GravitationalWave Tests of General Relativity with GroundBased Detectors and PulsarTiming Arrays
Abstract
This review is focused on tests of Einstein’s theory of general relativity with gravitational waves that are detectable by groundbased interferometers and pulsartiming experiments. Einstein’s theory has been greatly constrained in the quasilinear, quasistationary regime, where gravity is weak and velocities are small. Gravitational waves will allow us to probe a complimentary, yet previously unexplored regime: the nonlinear and dynamical strongfield regime. Such a regime is, for example, applicable to compact binaries coalescing, where characteristic velocities can reach fifty percent the speed of light and gravitational fields are large and dynamical. This review begins with the theoretical basis and the predicted gravitationalwave observables of modified gravity theories. The review continues with a brief description of the detectors, including both gravitationalwave interferometers and pulsartiming arrays, leading to a discussion of the data analysis formalism that is applicable for such tests. The review ends with a discussion of gravitationalwave tests for compact binary systems.
Keywords
General relativity Gravitational waves Pulsar timing Experimental tests Observational tests Alternative theories Compact binaries1 Introduction
1.1 The importance of testing
The era of precision gravitationalwave astrophysics is at our doorstep. With it, a plethora of previously unavailable information will flood in, allowing for unprecedented astrophysical measurements and tests of fundamental theories. Nobody would question the importance of more precise astrophysical measurements, but one may wonder whether fundamental tests are truly necessary, considering the many successes of Einstein’s theory of general relativity (GR). Indeed, GR has passed many tests, including solar system ones, binary pulsar ones and cosmological ones (for a recent review, see [438, 359]).
Let us provide some examples. For the EarthSun system, \(\mathcal{M}\) is essentially the mass of the sun, while \(\mathcal{R}\) is the orbital separation, which leads to \(\mathcal{C}\approx 9.8\times 10^{9}\) and \(\mathcal{V}\approx 9.9\times 10^{5}\). Even if an object were in a circular orbit at the surface of the sun, its gravitational compactness would be \(\mathcal{O}(10^{6})\) and its characteristic velocity \(\mathcal{O}(10^{3})\). Thus, we conclude that all solarsystem experiments are necessarily sampling the weak field regime of gravity. Similarly, for the binary pulsar J0737−3039 [298, 274], \(\mathcal{C}\approx 6\times 10^{6}\) and \(\mathcal{V}\approx 2\times 10^{3}\), where we have set the characteristic length \(\mathcal{R}\) to the orbital separation via \(\mathcal{R}\approx [MP^{2}/(4\pi^{2})]^{1/3}\approx 10^{6}\) km, where P = 0.1 days is the orbital period and M ≈ 3 M_{⊙} is the total mass. Although neutron stars are sources of strong gravity (the ratio of their mass to their radius is of order one tenth), binary pulsars are most sensitive to the quasistatic part of the postNewtonian effective potential or to the leadingorder (Newtonian piece) of the radiationreaction force. On the other hand, in compact binary coalescence the gravitational compactness and the characteristic speed can reach values much closer to unity. Therefore, although in much of the pulsartiming literature binary pulsar timing is said to allow for strongfield tests of gravity, gravitational information during compact binary coalescence would be a muchstronger—field test.
Even though current data does not give us access to the full nonlinear and dynamical regime of GR, solarsystem tests and binarypulsar observations have served (and will continue to serve) an invaluable role in testing Einstein’s theory. Solarsystem tests effectively cured an outbreak of modified gravity theories in the 1970s and 1980s, as summarized for example in [438]. Binary pulsars were crucial as the first indirect detectors of gravitational waves, and later to kill certain theories, like Rosen’s bimetric gravity [365], and heavily constrain others that predict dipolar energy loss, as we see in Sections 2 and 5. Binary pulsars are probes of GR in a certain sector of the strong field: in the dynamical but quasilinear sector, verifying that compact objects move as described by a perturbative, postNewtonian analysis to leading order. Binary pulsars can be used to test GR in the “strong field” only in the sense that they probe nonlinear stellarstructure effects, but they say very little to nothing about nonlinear radiative effects. Similarly, future electromagnetic observations of blackhole—accretion disks may probe GR in another strongfield sector: the nonlinear but fully stationary regime, verifying that black holes are described by the Kerr metric. As of this writing, only gravitational waves will allow for tests of GR in the full strongfield regime, where gravity is both heavily nonlinear and inherently dynamical.
No experiments exist to date that validate Einstein’s theory of GR in the highlydynamical, strongfield region. Due to previous successes of GR, one might consider such validation unnecessary. However, as most scientists would agree, the role of science is to predict and verify and not to assume without proof. Moreover, the incompleteness of GR in the quantum regime, together with the somewhat unsatisfactory requirement of the dark sector of cosmology (including dark energy and dark matter), have prompted more than one physicist to consider deviations from GR more seriously. Gravitational waves will soon allow us to verify Einstein’s theory in a regime previously inaccessible to us, and as such, these tests are invaluable.
However, in many areas of physics GR is so ingrained that questioning its validity (even in a regime where Einstein’s theory has not yet been validated) is synonymous with heresy. Dimensional arguments are usually employed to argue that any quantum gravitational correction will necessarily and unavoidably be unobservable with gravitational waves, as the former are expected at a (Planck) scale that is inaccessible to gravitationalwave detectors. This rationalization is dangerous, as it introduces a theoretical bias in the analysis of new observations of the universe, thus reducing the potential for new discoveries. For example, if astrophysicists had followed such a rationalization when studying supernova data, they would not have discovered that the universe is expanding. Dimensional arguments suggest that the cosmological constant is over 100 orders of magnitude larger than the value required to agree with observations. When observing the universe for the first time in a completely new way, it seems more conservative to remain agnostic about what is expected and what is not, thus allowing the data itself to guide our efforts toward theoretically understanding the gravitational interaction.
1.2 Testing general relativity versus testing alternative theories
When testing GR, one considers Einstein’s theory as a null hypothesis and searches for generic deviations. On the other hand, when testing alternative theories one starts from a particular modified gravity model, develops its equations and solutions and then predicts certain observables that then might or might not agree with experiment. Similarly, one may define a bottomup approach versus a topdown approach. In the former, one starts from some observables in an attempt to discover fundamental symmetries that may lead to a more complete theory, as was done when constructing the standard model of elementary particles. On the other hand, a topdown approach starts from some fundamental theory and then derives its consequence.
Both approaches possess strengths and weaknesses. In the topdown approach one has complete control over the theory under study, being able to write down the full equations of motion, answer questions about wellposedness and stability of solutions, and predict observables. But, as we see in Section 2, carrying out such an approach can be quixotic within any one model. What is worse, the lack of a complete and compelling alternative to GR makes choosing a particular modified theory difficult.
Given this, one might wish to attempt a bottomup approach, where one considers a set of principles one wishes to test without explicit mention of any particular theory. One usually starts by assuming GR as a nullhypothesis and then considers deformations away from GR. The hope is that experiments will be sensitive to such deformations, thus either constraining the size of the deformations or pointing toward a possible inconsistency. But if experiments do confirm a GR deviation, a bottomup approach fails at providing a given particular action from which to derive such a deformation. In fact, there can be several actions that lead to similar deformations, all of which can be consistent with the data within its experimental uncertainties.
Nonetheless, both approaches are complementary. The bottomup approach draws inspiration from particular examples carried out in the topdown approach. Given a plausible measured deviation from GR within a bottomup approach, one will still need to understand what plausible topdown theories can lead to such deviations. From this standpoint, then, both approaches are intrinsically intertwined and worth pursuing.
1.3 Gravitationalwave tests versus other tests of general relativity
Gravitationalwave tests differ from other tests of GR in many ways. Perhaps one of the most important differences is the spacetime regime gravitational waves sample. Indeed, as already mentioned, gravitational waves have access to the most extreme gravitational environments in nature. Moreover, gravitational waves travel essentially unimpeded from their source to Earth, and thus, they do not suffer from issues associated with obscuration. Gravitational waves also exist in the absence of luminous matter, thus allowing us to observe electromagnetically dark objects, such as blackhole inspirals.
This last point is particularly important as gravitational waves from inspiral—blackhole binaries are one of the cleanest astrophysical systems in nature. In the last stages of inspiral, when such gravitational waves would be detectable by groundbased interferometers, the evolution of a blackhole binary is essentially unaffected by any other matter or electromagnetic fields present in the system. As such, one does not need to deal with uncertainties associated with astrophysical matter. Unlike other tests of GR, such as those attempted with accretiondisk observations, blackhole—binary gravitationalwave tests may well be the cleanest probes of Einstein’s theory.
Of course, what is an advantage here, can also be a huge disadvantage in another context. Gravitational waves from compact binaries are intrinsically transient (they turn on for a certain amount of time and then shut off). This is unlike binary pulsar systems, for which astrophysicists have already collected tens of years of data. Moreover, gravitational wave tests rely on specific detections that cannot be anticipated beforehand. This is in contrast to Earthbased laboratory experiments, where one has complete control over the experimental setup. Finally, the intrinsic weakness of gravitational waves makes detection a very difficult task that requires complex dataanalysis algorithms to extract signals from the noise. As such, gravitationalwave tests are limited by the signaltonoise ratio and affected by systematics associated with the modeling of the waves, issues that are not as important in other loud astrophysical systems.
1.4 Groundbased vs spacebased detectors and interferometers vs pulsar timing
This review article focuses only on groundbased detectors, by which we mean both gravitationalwave interferometers, such as the Laser Interferometer Gravitational Observatory (LIGO) [3, 2, 217], Virgo [5, 6] and the Einstein Telescope (ET) [361, 377], as well as pulsartiming arrays (for a recent review of gravitationalwave tests of GR with spacebased detectors, see [183, 446]). Groundbased detectors have the limitation of being contaminated by manmade and naturemade noise, such as ground and air traffic, logging, earthquakes, ocean tides and waves, which are clearly absent in spacebased detectors. Groundbased detectors, however, have the clear benefit that they can be continuously upgraded and repaired in case of malfunction, which is obviously not possible with spacebased detectors.
As far as tests of GR are concerned, there is a drastic difference in spacebased and groundbased detectors: the gravitationalwave frequencies these detectors are sensitive to. For various reasons that we will not go into, spacebased interferometers are likely to have million kilometer long arms, and thus, be sensitive in the milliHz band. On the other hand, groundbased interferometers are bound to the surface and curvature of the Earth, and thus, they have kilometerlong arms and are sensitive in the deca and hectaHz band. Different types of interferometers are then sensitive to different types of gravitationalwave sources. For example, when considering binary coalescences, groundbased interferometers are sensitive to late inspirals and mergers of neutron stars and stellarmass black holes, while spacebased detectors will be sensitive to supermassive—blackhole binaries with masses around 10^{5} M_{⊙}.
The impact of a different population of sources in tests of GR depends on the particular modified gravity theory considered. When studying quadratic gravity theories, as we see in Section 2, the Einstein—Hilbert action is modified by introducing higherorder curvature operators, which are naturally suppressed by powers of the inverse of the radius of curvature. Thus, spacebased detectors will not be ideal at constraining these theories, as the radius of curvature of supermassive black holes is much larger than that of stellarmass black holes at merger. Moreover, spacebased detectors will not be sensitive to neutronstarbinary coalescences; they are sensitive to supermassive blackhole/neutronstar coalescences, where the radius of curvature of the system is controlled by the supermassive black hole.
On the other hand, spacebased detectors are unique in their potential to probe the spacetime geometry of supermassive black holes through gravitational waves emitted during extrememassratio inspirals. These inspirals consist of a stellarmass compact object in a generic decaying orbit around a supermassive black hole. Such inspirals produce millions of cycles of gravitational waves in the sensitivity band of spacebased detectors (in fact, they can easily outlive the observation time!). Therefore, even small changes to the radiationreaction force, or to the background geometry, can lead to noticeable effects in the waveform observable and thus strong tests of GR, albeit constrained to the radius of curvature of the supermassive black hole. For recent work on such systems and tests, see [23, 370, 371, 263, 196, 50, 289, 182, 390, 471, 31, 297, 184, 116, 93, 183].
Spacebased detectors also have the advantage of range, which is particularly important when considering theories where gravitons do not travel at light speed [316]. Spacebased detectors have a horizon distance much larger than groundbased detectors; the former can see blackhole mergers to redshifts of order 10 if there are any at such early times in the universe, while the latter are confined to events within redshift 1. Gravitational waves emitted from distant regions in spacetime need a longer time to propagate from the source to the detectors. Thus, theories that modify the propagation of gravitational waves will be best constrained by spacebased type systems. Of course, such theories are also likely to modify the generation of gravitational waves, which groundbased detectors should also be sensitive to.
Another important difference between detectors is in their response to an impinging gravitational wave. Groundbased detectors, as we see in Section 3, cannot separate between the two possible scalar modes (the longitudinal and the breathing modes) of metric theories of gravity, due to an intrinsic degeneracy in the response functions. Spacebased detectors in principle also possess this degeneracy, but they may be able to break it through Doppler modulation if the interferometer orbits the Sun. Pulsartiming arrays, on the other hand, lack this degeneracy altogether, and thus, they can in principle constrain the existence of both modes independently.
Pulsartiming arrays differ from interferometers in their potential to test GR mostly by the frequency space they are most sensitive to. The latter can observe the late inspiral and merger of compact binaries, while the former is restricted to the very early inspiral. This is why pulsar timing arrays do not need very accurate waveform templates that account for the highlydynamical and nonlinear nature of gravity to detect gravitational waves; leadingorder quadrupole waveforms are sufficient [120]. In turn, this implies that pulsar timing arrays cannot constrain theories that only deviate significantly from GR in the late inspiral, while they are exceptionally wellsuited for constraining lowfrequency deviations.

Groundbased detectors are best at constraining highercurvature type modified theories that deviate from GR the most in the late inspiral and merger phase.

Spacebased detectors are best at constraining modified graviton dispersion relations and the geometry of supermassive compact objects.

Pulsartiming arrays are best at independently constraining the existence of both scalar modes and any deviation from GR that dominates at low orbital frequencies.
1.5 Notation and conventions
We mainly follow the notation of [318], where Greek indices stand for spacetime coordinates and spatial indices in the middle of the alphabet (i, j, k, …) for spatial indices. Parenthesis and square brackets in index lists stand for symmetrization and antisymmetrization respectively, e.g., A_{(μν)} = (A_{ μν } + A_{ νμ })/2 and A_{[μν]} = (A_{ μν } − A_{ νμ })/2. Partial derivatives with respect to spacetime and spatial coordinates are denoted ∂_{ μ }A = A,_{ μ } and ∂_{ i }A = A,_{ i } respectively. Covariant differentiation is denoted ∇_{μA} = A_{;μ}, multiple covariant derivatives ∇^{ μν… } = ∇^{ μ }∇^{ ν } …, and the curved spacetime D’Alembertian □A = ∇_{ μ }∇^{ μ } A. The determinant of the metric g_{ μν } is g, R_{ μνδσ } is the Riemann tensor, R_{ μν } is the Ricci tensor, R is the Ricci scalar and g_{ μν } is the Einstein tensor. The LeviCivita tensor and symbol are ∈^{ μνδσ } and ∈^{ −μνδσ } respectively, with \({\bar \varepsilon ^{0123}} = + 1\) in an orthonormal, positivelyoriented frame. We use geometric units (G = c =1) and the Einstein summation convention is implied.
2 Alternative Theories of Gravity
In this section, we discuss the many possible alternative theories that have been studied so far in the context of gravitationalwave tests. We begin with a description of the theoretically desirable properties that such theories must have. We then proceed with a review of the theories so far explored as far as gravitational waves are concerned. We will leave out the description of many theories in this chapter, especially those which currently lack a gravitationalwave analysis. We will conclude with a brief description of unexplored theories as possible avenues for future research.
2.1 Desirable theoretical properties
 1.
Precision Tests. The theory must produce predictions that pass all solar system, binary pulsar, cosmological and experimental tests that have been carried out so far.
 1.a
General Relativity Limit. There must exist some limit, continuous or discontinuous, such as the weakfield one, in which the predictions of the theory are consistent with those of GR within experimental precision.
 1.b
Existence of Known Solutions [426]. The theory must admit solutions that correspond to observed phenomena, including but not limited to (nearly) flat spacetime, (nearly) Newtonian stars, and cosmological solutions.
 1.c
Stability of Solutions [426]. The special solutions described in property (1.b) must be stable to small perturbations on timescales smaller than the age of the universe. For example, perturbations to (nearly) Newtonian stars, such as impact by asteroids, should not render such solutions unstable.
 2.
Wellmotivated from Fundamental Physics. There must be some fundamental theory or principle from which the modified theory (effective or not) derives. This fundamental theory would solve some fundamental problem in physics, such as latetime acceleration or the incompatibility between quantum mechanics and GR.
 3.
Wellposed Initial Value Formulation [426]. A wide class of freely specifiable initial data must exist, such that there is a uniquely determined solution to the modified field equations that depends continuously on this data.
One might be concerned that Property (2) automatically implies that any predicted deviation to astrophysical observables will be too small to be detectable. This argument usually goes as follows. Any quantum gravitational correction to the action will “naturally” introduce at least one new scale, and this, by dimensional analysis, must be the Planck scale. Since this scale is usually assumed to be larger than 1 TeV in natural units (or 10^{−35} m in geometric units), gravitationalwave observations will never be able to observe quantumgravitational modifications (see, e.g., [155] for a similar argument). Although this might be true, in our view such arguments can be extremely dangerous, since they induce a certain theoretical bias in the search for new phenomena. For example, let us consider the supernova observations of the latetime expansion of the universe that led to the discovery of the cosmological constant. The above argument certainly fails for the cosmological constant, which on dimensional arguments is over 100 orders of magnitude too small. If the supernova teams had respected this argument, they would not have searched for a cosmological constant in their data. Today, we try to explain our way out of the failure of such dimensional arguments by claiming that there must be some exquisite cancellation that renders the cosmological constant small; but this, of course, came only after the constant had been measured. One is not trying to argue here that cancellations of this type are common and that quantum gravitational modifications are necessarily expected in gravitationalwave observations. Rather, we are arguing that one should remain agnostic about what is expected and what is not, and allow oneself to be surprised without suppressing the potential for new discoveries that will accompany the new era of gravitationalwave astrophysics.
 4.
Strong Field Inconsistency. The theory must lead to observable deviations from GR in the strongfield regime. Many modified gravity models have been proposed that pose infrared or cosmological modifications to GR, aimed at explaining certain astrophysical or cosmological observables, like the late expansion of the universe. Such modified models usually reduce to GR in the strongfield regime, for example via a Vainshteinlike mechanism [413, 140, 45] in a static sphericallysymmetric context. Extending this mechanism to highlydynamical strongfield scenarios has not been fully worked out yet [137, 138]. Gravitationalwave tests of GR, however, are concerned with modified theories that predict deviations in the strongfield, precisely where cosmological modified models do not. Clearly, Property (4) is not necessary for a theory to be a valid description of nature. This is because a theory might be identical to GR in the weak and strong fields, yet different at the Planck scale, where it would be unified with quantum mechanics. However, Property (4) is a desirable feature if one is to test this theory with gravitational wave observations.
2.2 Wellposedness and effective theories
However, such a perturbative analysis can say nothing about the wellposedness of the full theory from which the effective theory derives, or of the effective theory if treated as an exact one (i.e., not as a perturbative expansion). In fact, a wellposed full theory may have both stable and unstable solutions. The arguments presented above only discuss the stability of solutions in an effective theory, and thus, they are selfconsistent only within their perturbative scheme. A full theory may have nonperturbative instabilities, but these can only be studied once one has a full (nontruncated in g) theory, from which Eq. (6) derives as a truncated expansion. Lacking a full quantum theory of nature, quantum gravitational models are usually studied in a truncated lowenergy expansion, where the leadingorder piece is GR and higherorder pieces are multiplied by a small coupling constant. One can perturbatively explore the wellbehaved sector of the truncated theory about solutions to the leadingorder theory. However, such an analysis is incapable of answering questions about wellposedness or nonlinear stability of the full theory.
2.3 Explored theories
In this subsection we briefly describe the theories that have so far been studied in some depth as far as gravitational waves are concerned. In particular, we focus only on those theories that have been sufficiently studied so that predictions of the expected gravitational waveforms (the observables of gravitationalwave detectors) have been obtained for at least a typical source, such as the quasicircular inspiral of a compact binary.
2.3.1 Scalartensor theories
Let us now discuss whether scalartensor theories satisfy the properties discussed in Section 2.1. Massless JordanFierzBransDicke theory agrees with all known experimental tests provided ω_{BD} > 4 × 10^{4}, a bound imposed by the tracking of the Cassini spacecraft through observations of the Shapiro time delay [73]. Massive JordanFierzBransDicke theory has been recently constrained to ω_{BD} > 4 × 10^{4} and m_{s} < 2.5 × 10^{−20} eV, with m_{s} the mass of the scalar field [348, 20]. Of course, these bounds are not independent, as when m_{s} → 0 one recovers the standard massless constraint, while when m_{s} → ∞, ω_{BD} cannot be bounded as the scalar becomes nondynamical. Observations of the Nordtvedt effect with Lunar Laser Ranging observations, as well as observations of the orbital period derivative of whitedwarf/neutronstar binaries, yield similar constraints [131, 132, 20, 177]. Neglecting any homogeneous, cosmological solutions to the scalarfield evolution equation, it is clear that in the limit ω → ∞ one recovers GR, i.e., scalartensor theories have a continuous limit to Einstein’s theory, but see [164] for caveats for certain spacetimes. Moreover, [375, 278, 425] have verified that scalartensor theories with minimal or nonminimal coupling in the Jordan frame can be cast in a stronglyhyperbolic form, and thus, they possess a wellposed initialvalue formulation. Therefore, scalartensor theories possess both Properties (1) and (3).
Scalartensor theories also possess Property (2), since they can be derived from the lowenergy limit of certain string theories. The integration of string quantum fluctuations leads to a higherdimensional string theoretical action that reduces locally to a field theory similar to a scalartensor one [189, 176], the mapping being ϕ = e^{−2Ψ}, with Ψ one of the string moduli fields [133, 134]. Moreover, scalartensor theories can be mapped to f(R) theories, where one replaces the Ricci scalar by some functional of R. In particular, one can show that f(R) theories are equivalent to BransDicke theory with ω_{BD} = 0, via the mapping ϕ = df(R)/dR and V(ϕ) = R df (R)/dR − f(R) [104, 396]. For a recent review on this topic, see [135].
Black holes and stars continue to exist in scalartensor theories. Stellar configurations are modified from their GR profile [441, 131, 214, 215, 410, 132, 394, 139, 393, 235], while black holes are not, provided one neglects homogeneous, cosmological solutions to the scalar field evolution equation. Indeed, Hawking [224, 159, 222, 98, 244, 363] has proven that BransDicke black holes that are stationary and the endpoint of gravitational collapse are identical to those of GR. This proof has recently been extended to a general class of scalartensor models [398]. That is, stationary black holes radiate any excess “hair”, i.e., additional degrees of freedom, after gravitational collapse, a result sometimes referred to as the nohair theorem for black holes in scalartensor theories. This result has recently been extended even further to allow for quasistationary scenarios in generic scalartensor theories through the study of extrememassratio inspirals [465] (small black hole in orbit around a much larger one), postNewtonian comparablemass inspirals [315] and numerical simulations of comparablemass blackhole mergers [230, 67].
Damour and EspositoFarése [129, 130] proposed a different type of scalartensor theory, one that can be defined by the action in Eq. (15) but with the conformal factor \(A(\varphi) = {e^{\alpha \varphi + \beta {\varphi ^2}/2}}\) or the coupling function ω(ϕ) = −3/2 − 2πG/(β log ϕ), where α and β are constants. When β = 0 one recovers standard BransDicke theory. When β ≲ −4, nonperturbative effects that develop if the gravitational energy is large enough can force neutron stars to spontaneously acquire a nontrivial scalar field profile, to spontaneously scalarize. Through this process, a neutronstar binary that initially had no scalar hair in its early inspiral would acquire it before merger, when the binding energy exceeded some threshold [51]. Binary pulsar observations have constrained this theory in the (α,β) space; very roughly speaking β > −4 and α < 10^{−2} [131, 132, 177]
As for Property (4), scalar tensor theories are not built with the aim of introducing strongfield corrections to GR.^{3} Instead, they naturally lead to modifications of Einstein’s theory in the weakfield, modifications that dominate in scenarios with sufficiently weak gravitational interactions. Although this might seem strange, it is natural if one considers, for example, one of the key modifications introduced by scalartensor theories: the emission of dipolar gravitational radiation. Such dipolar emission dominates over the general relativistic quadrupolar emission for systems in the weak to intermediate field regime, such as in binary pulsars or in the very early inspiral of compact binaries. Therefore, one would expect scalartensor theories to be best constrained by experiments or observations of weaklygravitating systems, as it has recently been explicitly shown in [465].
2.3.2 Massive graviton theories and Lorentz violation
Massive graviton theories are those in which the gravitational interaction is propagated by a massive gauge boson, i.e., a graviton with mass m_{ g } ≠ 0 or Compton wavelength λ_{ g } = h/(m_{ g }c) < ∞. Einstein’s theory predicts massless gravitons and thus gravitational propagation at light speed, but if this were not the case, then a certain delay would develop between electromagnetic and gravitational signals emitted simultaneously at the source. Fierz and Pauli [169] were the first to write down an action for a free massive graviton, and ever since then, much work has gone into the construction of such models. For a detailed review, see, e.g., [232].
Gravitational theories with massive gravitons are somewhat wellmotivated from a fundamental physics perspective, and thus, one can say they possess Property (2). Indeed, in loop quantum cosmology [42, 77], the cosmological extension to loop quantum gravity, the graviton dispersion relation acquires holonomy corrections during loop quantization that endow the graviton with a mass [78] m_{ g } = Δ^{−1/2}γ^{−1}(ρ/ρ_{ c }), with γ the BarberoImmirzi parameter, Δ the area operator, and ρ and ρ_{ c } the total and critical energy density respectively. In stringtheoryinspired effective theories, such as Dvali’s compact, extradimensional theory [157], such massive modes also arise
Massive graviton modes also occur in many other modified gravity models. In Rosen’s bimetric theory [365], for example, photons and gravitons follow null geodesics of different metrics [438, 435]. In Visser’s massive graviton theory [424], the graviton is given a mass at the level of the action through an effective perturbative description of gravity, at the cost of introducing a nondynamical background metric, i.e., a prior geometry. A recent reincarnation of this model goes by the name of bigravity, where again two metric tensors are introduced [349, 346, 219, 220]. In Bekenstein’s TensorVectorScalar (TeVeS) theory [54], the existence of a scalar and a vector field lead to subluminal gravitationalwave propagation.
Massive graviton theories have a theoretical issue, the van DamVeltmanZakharov (vDVZ) discontinuity [418, 475], which is associated with Property 1.a, i.e., a GR limit. The problem is that certain predictions of massive graviton theories do not reduce to those of GR in the mg → 0 limit. This can be understood qualitatively by studying how the 5 spin states of the graviton behave in this limit. Two of them become the two GR helicity states of the massless graviton. Another two become helicity states of a massless vector that decouples from the tensor perturbations in the m_{ g } → 0 limit. However, the last state, the scalar mode, retains a finite coupling to the trace of the stressenergy tensor in this limit. Therefore, massive graviton theories in the m_{ g } → 0 limit do not reduce to GR, since the scalar mode does not decouple.
However, the vDVZ discontinuity can be evaded, for example, by carefully including nonlinearities. Vainshtein [413, 269, 140, 45] showed that around any sphericallysymmetric source of mass M, there exists a certain radius \(r < {r_V} \equiv {({r_S}\lambda _g^4)^{1/5}}\), with r_{ S } the Schwarzschild radius, where linear theory cannot be trusted. Since r_{ V } → ∞ as m_{ g } → 0, this implies that there is no radius at which the linear approximation (and thus vDVZ discontinuity) can be trusted. Of course, to determine then whether massive graviton theories have a continuous limit to GR, one must include nonlinear corrections to the action (see also an argument by [34]), which are more difficult to uniquely predict from fundamental theory. Recently, there has been much activity in the development of new, nonlinear massive gravity theories [60, 136, 211, 61, 137, 138].
 (i)
Modification to Newton’s laws;
 (ii)
Modification to gravitational wave propagation.
From the structure of the above phenomenological modifications, it is clear that GR can be recovered in the m_{ g } → 0 limit, avoiding the vDVZ issue altogether by construction. Such phenomenological modifications have been constrained by several types of experiments and observations. Using the modification to Newton’s third law and precise observations of the motion of the inner planets of the solar system together with Kepler’s third law, [437] found a bound of λ_{ g } > 2.8 × 10^{12} km. Such a constraint is purely static, as it does not sample the radiative sector of the theory. Dynamical constraints, however, do exist: through observations of the decay of the orbital period of binary pulsars, [174] found a bound of λ_{ g } > 1.6 × 10^{10} km;^{4} by investigating the stability of Schwarzschild and Kerr black holes, [88] placed the constraint λ_{ g } > 2.4 × 10^{13} km in FierzPauli theory [169]. New constraints that use gravitational waves have been proposed, including measuring a difference in time of arrival of electromagnetic and gravitational waves [126, 266], as well as direct observation of gravitational waves emitted by binary pulsars (see Section 5).
Although massive gravity theories unavoidably lead to a modification to the graviton dispersion relation, the converse is not necessarily true. A modification of the dispersion relation is usually accompanied by a modification to either the Lorentz group or its action in real or momentum space. Such Lorentzviolating effects are commonly found in quantum gravitational theories, including loop quantum gravity [78] and string theory [107, 403], as well as other effective models [58, 59]. In doublyspecial relativity [26, 300, 27, 28], the graviton dispersion relation is modified at high energies by modifying the law of transformation of inertial observers. Modified graviton dispersion relations have also been shown to arise in generic extradimensional models [381], in HořavaLifshitz theory [233, 234, 412, 76] and in theories with noncommutative geometries [186, 187, 188]. None of these theories necessarily requires a massive graviton, but rather the modification to the dispersion relation is introduced due to Lorentzviolating effects.
One might be concerned that the mass of the graviton and subsequent modifications to the graviton dispersion relation should be suppressed by the Planck scale. However, Collins, et al. [111, 110] have suggested that Lorentz violations in perturbative quantum field theories could be dramatically enhanced when one regularizes and renormalizes them. This is because terms that vanish upon renormalization due to Lorentz invariance do not vanish in Lorentzviolating theories, thus leading to an enhancement [185]. Whether such an enhancement is truly present cannot currently be ascertained.
2.3.3 Modified quadratic gravity
The action in Eq. (25) is wellmotivated from fundamental theories, as it contains all possible quadratic, algebraic curvature scalars with running (i.e., nonconstant) couplings. The only restriction here is that all quadratic terms are assumed to couple to the same field, which need not be the case. For example, in string theory some terms might couple to the dilaton (a scalar field), while others couple to the axion (a pseudo scalar field). Nevertheless, one can recover wellknown and motivated modified gravity theories in simple cases. For example, dynamical ChernSimons modified gravity [17] is recovered when α_{4} = −α_{ CS }/4 and all other α_{ i } = 0. EinsteinDilatonGaussBonnet gravity [343] is obtained when α_{4} = 0 and (α_{1}, α_{2}, α_{3}) = (1, −4, 1)α_{EDGB}.^{5} Both theories unavoidably arise as lowenergy expansions of heterotic string theory [203, 204, 12, 89]. As such, modified quadratic gravity theories should be treated as a class of effective field theories. Moreover, dynamical ChernSimons gravity also arises in loop quantum gravity [43, 366] when the BarberoImmirzi parameter is promoted to a field in the presence of fermions [41, 16, 406, 311, 192].
One should make a clean and clear distinction between the theory defined by the action of Eq. (25) and that of f(R) theories. The latter are defined as functionals of the Ricci scalar only, while Eq. (25) contains terms proportional to the Ricci tensor and Riemann tensor squared. One could think of the subclass of f(R) theories with f(R) = R^{2} as the limit of modified quadratic gravity with only α_{1} = 0 and f_{1}(ϑ) = 1. In that very special case, one can map quadratic gravity theories and f(R) gravity to a scalartensor theory. Another important distinction is that f(R) theories are usually treated as exact, while the action presented above is to be interpreted as an effective theory [89] truncated to quadratic order in the curvature in a lowenergy expansion of a more fundamental theory. This implies that there are cubic, quartic, etc. terms in the Riemann tensor that are not included in Eq. (25) and that presumably depend on higher powers of α_{ i }. Thus, when studying such an effective theory one should also orderreduce the field equations and treat all quantities that depend on perturbatively, the smallcoupling approximation. One can show that such an order reduction removes any additional polarization modes in propagating metric perturbations [390, 400] that naturally arise in f(R) theories. In analogy to the treatment of the Ostrogradski instability in Section 2.1, one would also expect that orderreduction would lead to a theory with a wellposed initialvalue formulation.
Here, S_{GR} is the EinsteinHilbert plus matter action, while S_{0} and S_{1} are corrections. The former is decoupled from υ, where the omitted term proportional to \(\alpha _4^{(0)}\) does not affect the classical field equations since it is topological, i.e., it can be rewritten as the total 4divergence of some 4current. Similarly, if the \(\alpha _i^{(0)}\) were chosen to reconstruct the GaussBonnet invariant, \((\alpha _1^{(0)},\alpha _2^{(0)},\alpha _3^{(0)}) = (1,  4,1){\alpha _{{\rm{GB}}}}\), then this combination would also be topological and not affect the classical field equations. On the other hand, S_{1} is a modification to GR with a direct (nonminimal) coupling to ϑ such that as the field goes to zero, the modified theory reduces to GR.
Another restriction one usually makes to simplify modified gravity theories is to neglect the \(\alpha _i^{(0)}\) terms and only consider the S_{1} modification, the restricted modified quadratic gravity. The \(\alpha _i^{(0)}\) terms represent corrections that are nondynamical. The term proportional to \(\alpha _1^{(0)}\) resembles a certain class of f(R) theories. As such, it can be mapped to a scalartensor theory with a complicated potential, which has been heavily constrained by torsionbalance EötWash experiments to \(\alpha _1^{(0)} < 2 \times {10^{ 8}}{{\rm{m}}^2}\) [237, 259, 62]. Moreover, these theories have a fixed coupling constant that does not run with energy or scale. In restricted modified gravity, the scalar field is effectively forcing the running of the coupling.
From the structure of the above equations, it should be clear that the dynamics of ϑ guarantee that the modified field equations are covariantly conserved exactly. That is, one can easily verify that the covariant divergence of Eq. (27) identically vanishes upon imposition of Eq. (30). Such a result had to be so, as the action is diffeomorphism invariant. If one neglected the kinetic and potential energies of ϑ in the action, as was originally done in [245], the theory would possess preferredframe effects and would not be covariantly conserved. Moreover, such a theory requires an additional constraint, i.e., the righthand side of (30) would have to vanish, which is an unphysical consequence of treating ϑ as prior structure [470, 207].
One last simplification that is usually made when studying modified quadratic gravity theories is to ignore the potential V(ϑ), i.e., set V(ϑ) = 0. This potential can in principle be nonzero, for example if one wishes to endow ϑ with a mass or if one wishes to introduce a cosine driving term, like that for axions in field and string theory. However, reasons exist to restrict the functional form of such a potential. First, a mass for ϑ will modify the evolution of any gravitational degree of freedom only if this mass is comparable to the inverse length scale of the problem under consideration (such as a binary system). This could be possible if there is an incredibly large number of fields with different masses in the theory, such as perhaps in the string axiverse picture [40, 268, 303]. However, in that picture the moduli fields are endowed with a mass due to shiftsymmetry breaking by nonperturbative effects; such masses are not expected to be comparable to the inverse length scale of binary systems. Second, no mass term may appear in a theory with a shift symmetry, i.e., invariance under the transformation ϑ → ϑ + const. Such symmetries are common in fourdimensional, lowenergy, effective string theories [79, 204, 203, 92, 89], such as dynamical Chern—Simons and EinsteinDilatonGaussBonnet theory. Similar considerations apply to other more complicated potentials, such as a cosine term.
Lastly, let us discuss what is known about whether modified quadratic gravities satisfy the requirements discussed in Section 2.1. As it should be clear from the action itself, this modified gravity theory satisfies the fundamental requirement, i.e., passing all precision tests, provided the couplings α_{ i } are sufficiently small. This is because such theories have a continuous limit to GR as α_{ i } → 0.^{6} Dynamical Chern—Simons gravity is constrained only weakly at the moment, \(\xi _4^{1/4} < {10^8}\) km, where \({\xi _4} \equiv \alpha _4^2/(\beta \kappa)\), only through observations of Lense—Thirring precession in the solar system [19]. The EinsteinDilatonGauss—Bonnet gravity coupling constant \({\xi _3} \equiv \alpha _3^2/(\beta \kappa)\), on the other hand, has been constrained by several experiments: solar system observations of the Shapiro time delay with the Cassini spacecraft placed the bound \(\xi _3^{1/4} < 1.3 \times {10^7}\) km [73, 29]; the requirement that neutron stars still exist in this theory placed the constraint \(\xi _3^{1/4} \underset{\sim}{<}26\) km [342], with the details depending somewhat on the central density of the neutron star; observations of the rate of change of the orbital period in the lowmass Xray binary A0620−00 [358, 255] has led to the constraint \(\xi _3^{1/4} < 1.9\) km [445].
However, not all subproperties of the fundamental requirement are known to be satisfied. One can show that certain members of modified quadratic gravity possess known solutions and these are stable, at least in the smallcoupling approximation. For example, in dynamical Chern—Simons gravity, sphericallysymmetric vacuum solutions are given by the Schwarzschild metric with constant ϑ to all orders in α_{ i } [245, 470]. Such a solution is stable to small perturbations [319, 190], as also are nonspinning black holes and branes in anti de Sitter space [144]. On the other hand, spinning solutions continue to be elusive, with approximate solutions in the slowrotation/smallcoupling limit known both for black holes [466, 272, 345, 455] and stars [19, 342]; nothing is currently known about the stability of these spinning solutions. In EinsteinDilatonGauss—Bonnet theory even sphericallysymmetric solutions are modified [473, 345] and these are stable to axial perturbations [343].
As for the other requirements discussed in Section 2.1, it is clear that modified quadratic gravity is wellmotivated from fundamental theory, but it is not clear at all whether it has a wellposed initialvalue formulation. From an effective point of view, a perturbative treatment in α_{ i } naturally leads to stable solutions and a wellposed initialvalue problem, but this is probably not the case when it is treated as an exact theory. In fact, if one were to treat such a theory as exact (to all orders in α_{ i }), then the evolution system would likely not be hyperbolic, as higherthansecond time derivatives now drive the evolution. Although no proof exists, it is likely that such an exact theory is not wellposed as an initialvalue problem. Notice, however, that this says nothing about the fundamental theories that modified quadratic gravity derives from. This is because even if the truncated theory were ill posed, higherorder corrections that are neglected in the truncated version could restore wellposedness.
As for the last requirement (that the theory modifies the strong field), modified quadratic theories are ideal in this respect. This is because they introduce corrections to the action that depend on higher powers of the curvature. In the strongfield, such higher powers could potentially become nonnegligible relative to the Einstein—Hilbert action. Moreover, since the curvature scales inversely with the mass of the objects under consideration, one expects the largest deviations in systems with small total mass, such as stellarmass blackhole mergers. On the other hand, deviations from GR should be small for small compact objects spiraling into a supermassive black hole, since here the spacetime curvature is dominated by the large object, and thus it is small, as discussed in [390].
2.3.4 Variable G theories and large extra dimensions
Variable G theories are defined as those where Newton’s gravitational constant is promoted to a spacetime function. Such a modification breaks the principle of equivalence (see [438]) because the laws of physics now become local position dependent. In turn, this implies that experimental results now depend on the spacetime position of the laboratory frame at the time of the experiment.
Many known alternative theories that violate the principle of equivalence, and in particular, the strong equivalence principle, predict a varying gravitational constant. A classic example is scalartensor theory [435], which, as explained in Section 2.3.1, modifies the gravitational sector of the action by multiplying the Ricci scalar by a scalar field (in the Jordan frame). In such theories, one can effectively think of the scalar as promoting the coupling between gravity and matter to a fielddependent quantity G → G(ϕ), thus violating local position invariance when ϕ varies. Another example are bimetric theories, such as that of Lightman—Lee [293], where the gravitational constant becomes timedependent even in the absence of matter, due to possibly timedependent cosmological evolution of the prior geometry. A final example are higherdimensional, braneworld scenarios, where enhanced Hawking radiation inexorably leads to a timevarying effective 4D gravitational constant [141], whose rate of change depends on the curvature radius of extra dimensions [255].
An important point to address is whether variable G theories can lead to modifications to a vacuum spacetime, such as a blackhole—binary inspiral. In Einstein’s theory, G appears as the coupling constant between geometry, encoded by the Einstein tensor g_{ μν }, and matter, encoded by the stress energy tensor \(T_{\mu \nu}^{{\rm{mat}}}\). When considering vacuum spacetimes, \(T_{\mu \nu}^{{\rm{mat}}} = 0\) and one might naively conclude that a variable G would not introduce any modification to such spacetimes. In fact, this is the case in scalartensor theories (without homogeneous, cosmological solutions to the scalar field equation), where the nohair theorem establishes that blackhole solutions are not modified. On the other hand, scalartensor theories with a cosmological, homogeneous scalar field solution can violate the nohair theorem, endowing black holes with timedependent hair, which in turn would introduce variability into G even in vacuum spacetimes [246, 236, 67].
In general, Newton’s constant plays a much more fundamental role than merely a coupling constant: it defines the relationship between energy and length. For example, for the vacuum Schwarzschild solution, G establishes the relationship between the radius R of the black hole and the restmass energy E of the spacetime via R = 2 GE/c^{4}. Similarly, in a blackhole—binary spacetime, each black hole introduces an energy scale into the problem that is quantified by a specification of Newton’s constant. Therefore, one can treat variable G modifications as induced by some effective theory that modifies the mapping between the curvature scale and the energy scale of the problem, as is done for example in theories with extra dimensions.
An explicit example of this idea is realized in braneworld models. Superstring theory suggests that physics should be described by 4 large dimensions, plus another 6 that are compactified and very small [354, 355]. The size of these extra dimensions is greatly constrained by particle theory experiments. However, braneworld models, where a certain higherdimensional membrane is embedded in a higherdimensional bulk spacetime, can evade this constraint as only gravitons can interact with the bulk. The ADD model [32, 33] is a particular example of such a braneworld, where the bulk is flat and compact and the brane is tensionless with ordinary fields localized on it. Since gravitationalwave experiments have not yet constrained deviations from Einstein’s theory in the strong field, the size of these extra dimensions is constrained to micrometer scales only by tabletop experiments [259, 7].
What is relevant to gravitationalwave experiments is that in many of these braneworld models black holes may not remain static [163, 405]. The argument goes roughly as follows: a fivedimensional black hole is dual to a fourdimensional one with conformal fields on it by the ADS/CFT conjecture [301, 9], but since the latter must evolve via Hawking radiation, the black hole must be losing mass. The Hawking mass loss rate is here enhanced by the large number of degrees of freedom in the conformal field theory, leading to an effective modification to Newton’s laws and to the emission of gravitational radiation. Effectively, one can think of the blackhole mass loss as due to the black hole being stretched away from the brane into the bulk due to a universal acceleration, that essentially reduces the size of the branelocalized black hole. For blackhole binaries, one can then draw an analogy between this induced time dependence in the blackhole mass and a variable G theory, where Newton’s constant decays due to the presence of black holes. Of course, this is only analogy, since large extra dimensions would not predict a timeevolving mass in neutronstar binaries.
Recently, however, Figueras et al. [170, 172, 171] numerically found stable solutions that do not require a radiation component. If such solutions were the ones realized in nature as a result of gravitational collapse on the brane, then the black hole mass would be time independent, up to quantum correction due to Hawking evaporation, a negligible effect for realistic astrophysical systems. Unfortunately, we currently lack numerical simulations of the dynamics of gravitational collapse in such scenarios.
Many experiments have been carried out to measure possible deviations from a constant G value, and they can broadly be classified into two groups: (a) those that search for the present or nearly present rate of variation (at redshifts close to zero); (b) those that search for secular variations over long time periods (at very large redshifts). Examples of experiments or observations of the first class include planetary radar ranging [350], surface temperature observations of lowredshift millisecond pulsars [249, 362], lunar ranging observations [442] and pulsar timing observations [260, 143], the latter two being the most stringent. Examples of experiments of the second class include the evolution of the sun [208] and BigBang Nucleosynthesis (BBN) calculations [119, 47], again with the latter being more stringent. For either class, the strongest constraints are about Ġ/G ≲ 10^{−13} yr^{−1}, varying somewhat from experiment to experiment.
Lacking a particularly compelling action to describe variable G theories, one is usually left with a phenomenological model of how such a modification to Einstein’s theory would impact gravitational waves. Given that the part of the waveform that detectors are most sensitive to is the gravitational wave phase, one can model the effect of variable G theories by studying how the rate of change of its frequency would be modified. Assuming a Taylor expansion for Newton’s constant one can derive the modification to the evolution equation for the gravitational wave frequency, given whichever physical scenario one is considering. Solving such an evolution equation then leads to a modification in the accumulated gravitationalwave phase observed at detectors on Earth. In Section 5 we will provide an explicit example of this for a compact binary system.
Let us discuss whether such theories satisfy the criteria defined in Section 2.1. The fundamental property can be satisfied if the rate of change of Newton’s constant is small enough, as variable G theories usually have a continuous limit to GR (as all derivatives of G go to zero). Whether variable theories are wellmotivated from fundamental physics (Property 2) depends somewhat on the particular effective model or action that one considers. But in general, Property 2 is usually satisfied, considering that such variability naturally arises in theories with extra dimensions, and the latter are also natural in all string theories. However, variable theories usually fail at introducing modifications in the strongfield region. Usually, such variability is parameterized as a Taylor expansion about some initial point with constant coefficients. That is, the variability of is not usually constructed so as to become stronger closer to merger.
The wellposed property and the subproperties of the fundamental property depend somewhat on the particular effective theory used to describe varying G modifications. In the f(R) case, one can impose restrictions on the functional form f(·) such that no ghosts (f′ > 0) or instabilities (f″ > 0) arise [180]. This, of course, does not guarantee that this (or any other such) theory is well posed. A much more detailed analysis would be required to prove wellposedness of the class of theories that lead to a variable Newton’s constant, but such is currently lacking.
2.3.5 Noncommutative geometry
Noncommutative geometry is a gravitational theory that generalizes the continuum Riemannian manifold of Einstein’s theory with the product of it with a tiny, discrete, finite noncommutative space, composed of only two points. Although the noncommutative space has zero spacetime dimension, as the product manifold remains four dimensional, its internal dimensions are 6 to account for Weyl and chiral fermions. This space is discrete to avoid the infinite tower of massive particles that would otherwise be generated, as in string theory. Through this construction, one can recover the standard model of elementary particles, while accounting for all (elementary particle) experimental data to date. Of course, the simple noncommutative space described above is expected to be replaced by a more complex model at Planckian energies. Thus, one is expected to treat such noncommutative geometry models as effective theories. Essentially nothing is currently known about the full noncommutative theory of which the theories described in this section are an effective lowenergy limit.
Before proceeding with an actionprinciple description of noncommutative geometry theories, we must distinguish between the spectral geometry approach championed by Connes [114], and Moyaltype noncommutative geometries [389, 206, 322]. In the former, the manifold is promoted to a noncommutative object through the product of a Riemann manifold with a noncommutative space. In the latter, instead, a nontrivial set of commutation relations is imposed between operators corresponding to position. These two theories are in principle unrelated. In this review, we will concentrate only on the former, as it is the only type of noncommutative GR extension that has been studied in the context of gravitationalwave theory.
Let us now discuss whether such a theory satisfies the properties discussed in Section 2.1. Noncommutative geometry theories clearly possess the fundamental property, as one can always take α_{0} → 0 (or equivalently β^{−2} → 0) to recover GR. Therefore, there must exist a sufficiently small α_{0} such that all precision tests carried out to date are satisfied. As for the existence and stability of known solutions, [326, 325] have shown that Minkowski spacetime is stable only for α_{0} < 0, as otherwise a tachyonic term appears in the evolution of the metric perturbation, as can be seen from Eq. (42). This then automatically implies that β must be real.
Current constraints on Weyl terms of this form come mostly from solar system experiments. Ni [328] recently studied an action of the form of Eq. (38) minimally coupled to matter in light of solar system experiments. He calculated the relativistic Shapiro timedelay and light deflection about a massive body in this theory and found that observations of the Cassini satellite place constraints on α_{0}^{1/2} < 5.7 km [328]. This is currently the strongest bound we are aware of on α_{0}.
2.3.6 Gravitational parity violation
Parity, the symmetry transformation that flips the sign of the spatial triad, has been found to be broken in the standard model of elementary interactions. Only the combination of charge conjugation, parity transformation and time inversion (CPT) still remains a true symmetry of the standard model. Experimentally, it is curious that the weak interaction exhibits maximal parity violation, while other fundamental forces seem to not exhibit any. Theoretically, parity violation unavoidably arises in the standard model [55, 8, 21], as there exist oneloop chiral anomalies that give rise to parityviolating terms coupled to lepton number [428]. In certain sectors of string theory, such as in heterotic and Type I superstring theories, parity violation terms are also generated through the Green—Schwarz gauge anomalycanceling mechanism [204, 355, 12]. Finally, in loop quantum gravity [41], the scalarization of the Barbero—Immirzi parameter coupled to fermions leads to an effective action that contains parityviolating terms [406, 90, 311, 192]. Even without a particular theoretical model, one can show that effective field theories of inflation generically contain nonvanishing, secondorder, parityviolating curvature corrections to the Einstein—Hilbert action [429]. Alternatively, phenomenological parityviolating extensions of GR have been proposed through a scalarization of the fundamental constants of nature [115].
Let us now discuss the properties of such an effective theory. Because of the structure of the modification to the field equations, one can always choose a sufficiently small value for α such that all solar system tests are satisfied. In fact, one can see from the equations in this section that in the limit α → 0, one recovers GR. Nondynamical Chern—Simons gravity leads to modifications to the nonradiative (nearzone) metric in the gravitomagnetic sector, leading to corrections to Lense—Thirring precession [14, 15]. This fact has been used to constrain the theory through observations of the orbital motion of the LAGEOS satellites [388] to \((\alpha/\kappa){\dot \vartheta ^{ 1}} < 2 \times {10^4}\) km, or equivalently \((\kappa/\alpha){\dot \vartheta ^{ 1}} \underset{\sim}{>}{10^{ 14}}{\rm{eV}}\). However, much better constraints can be placed through observations of the binary pulsar [472, 18]: \(({\alpha _4}/\kappa)\dot \vartheta < 0.8\) km.
Some of the subproperties of the fundamental requirement are satisfied in nondynamical Chern—Simons gravity. On the one hand, all sphericallysymmetric metrics that are solutions to the Einstein equations are also solutions in this theory for a “canonical” scalar field (θ ∞ t) [207]. On the other hand, axisymmetric solutions to the Einstein equations are generically not solutions in this theory. Moreover, although sphericallysymmetric solutions are preserved, perturbations of such spacetimes that are solutions to the Einstein equations are not generically solutions to the modified theory [470]. What is perhaps worse, the evolution of perturbations to nonspinning black holes have been found to be generically overconstrained [470]. This is a consequence of the lack of scalar field dynamics in the modified theory, which, via Eq. (47), tends to overconstrain it. Such a conclusion also suggests that this theory does not posses a wellposed initialvalue problem.
One can argue that nondynamical Chern—Simons gravity is wellmotivated from fundamental theories [17], except that in the latter, the scalar field is always dynamical, instead of having to be prescribed a priori. Thus, perhaps the strongest motivation for such a model is as a phenomenological proxy to test whether the gravitational interaction remains parity invariant in the strong field, a test that is uniquely suited to this modified model.
2.4 Currently unexplored theories in the gravitationalwave sector
3 Detectors
3.1 Gravitationalwave interferometers
3.2 Pulsar timing arrays
Neutron stars can emit powerful beams of radio waves from their magnetic poles. If the rotational and magnetic axes are not aligned, the beams sweep through space like the beacon on a lighthouse. If the line of sight is aligned with the magnetic axis at any point during the neutron star’s rotation the star is observed as a source of periodic radiowave bursts. Such a neutron star is referred to as a pulsar. Due to their large moment of inertia pulsars are very stable rotators, and their radio pulses arrive on Earth with extraordinary regularity. Pulsar timing experiments exploit this regularity: gravitational waves are expected to cause fluctuations in the time of arrival of radio pulses from pulsars.
4 Testing Techniques
4.1 Coalescence analysis
Gravitational waves emitted during the inspiral, merger and ringdown of compact binaries are the most studied in the context of data analysis and parameter estimation. In this section, we will review some of the main data analysis techniques employed in the context of parameter estimation and tests of GR. We begin with a discussion of matched filtering and Fisher theory (for a detailed review, see [173, 103, 125, 174, 248]). We then continue with a discussion of Bayesian parameter estimation and hypothesis testing (for a detailed review, see [387, 205, 123, 294]).
4.1.1 Matched filtering and Fisher’s analysis
The Fisher method to estimate projected constraints on modified gravity theory parameters is as follows. First, one constructs a waveform model in the particular modified gravity theory one wishes to constrain. Usually, this waveform will be similar to the GR one, but it will contain an additional parameter, κ, such that the template parameters are now λ^{ i } plus κ. Let us assume that as κ → 0, the modified gravity waveform reduces to the GR expectation. Then, the accuracy to which κ can be measured, or the accuracy to which we can say κ is zero, is approximately (Σ^{ κκ })^{1/2}, where the Fisher matrix associated with this variancecovariance matrix must be computed with the nonGR model evaluated at the GR limit (κ → 0). Such a method for estimating how well modified gravity theories can be constrained was pioneered by Will in [436, 353], and since then, it has been widely employed as a firstcut estimate of the accuracy to which different theories can be constrained.
The Fisher method described above can dangerously lead to incorrect results if abused [414, 415]. One must understand that this method is suitable only if the noise is stationary and Gaussian and if the SNR is sufficiently large. How large an SNR is required for Fisher methods to work depends somewhat on the signals considered, but usually for applications concerning tests of GR, one would be safe with ρ ≳ 30 or so. In real data analysis, the first two conditions are almost never satisfied. Moreover, the first detections that will be made will probably be of low SNR, i.e., ρ ∼ 8, for which again the Fisher method fails. In such cases, more sophisticated parameter estimation methods need to be employed.
4.1.2 Bayesian theory and model testing
When hypothesis A and B refer to fundamental theories of nature we can take different viewpoints regarding the priors. If we argue that we know nothing about whether hypothesis A or B better describes nature, then we would assign equal priors to both hypotheses. If, on the other hand, we believe GR is the correct theory of nature, based on all previous experiments performed in the solar system and with binary pulsars, then we would assign p(A) > p(B). This assigning of priors necessarily biases the inferences derived from the calculated posteriors, which is sometimes heavily debated when comparing Bayesian theory to a frequentist approach. However, this “biasing” is really unavoidable and merely a reflection of our state of knowledge of nature (for a more detailed discussion on such issues, please refer to [294]).
The main difficulty in Bayesian inference (both in parameter estimation and model selection) is sampling the PDF sufficiently accurately. Several methods have been developed for this purpose, but currently the two main workhorses in gravitationalwave data analysis are Markov Chain Monte Carlo and Nested Sampling. In the former, one samples the likelihood through the MetropolisHastings algorithm [314, 221, 122, 367]. This is computationally expensive in highdimensional cases, and thus, there are several techniques to improve the efficiency of the method, e.g., parallel tempering [402]. Once the PDF has been sampled, one can then calculate the evidence integral, for example via thermodynamic integration [420, 167, 419]. In Nested Sampling, the evidence is calculated directly by laying out a fixed number of points in the prior volume, which are then allowed to move and coalesce toward regions of high posterior probability. With the evidence in hand, one can then infer the PDF. As in the previous case, Nested Sampling can be computationally expensive in highdimensional cases.
Del Pozzo et al. [142] were the first to carry out a Bayesian implementation of model selection in the context of tests of GR. Their analysis focused on tests of a particular massive graviton theory, using the gravitational wave signal from quasicircular inspiral of nonspinning black holes. Cornish et al. [124, 376] extended this analysis by considering modelindependent deviations from GR, using the parameterized postEinsteinian (ppE) approach (Section 5.3.4) [467]. Recently, this was continued by Li et al. [290, 291], who carried out a similar analysis on a large statistical sample of Advanced LIGO (aLIGO) detections using simulated data and a restricted ppE model. All of these studies suggest that Bayesian tests of GR are possible, given sufficientlyhigh SNR events. Of course, whether deviations from GR are observable will depend on the strongfield character and strength of the deviation, as well as the availability of sufficientlyaccurate GR waveforms.
4.1.3 Systematics in model selection

Mismodeling Systematic, caused by inaccurate models of the gravitationalwave template.

Instrumental Systematic, caused by inaccurate models of the gravitationalwave response.

Astrophysical Systematic, caused by inaccurate models of the astrophysical environment.
Mismodeling systematics will prevent us from testing GR effectively with signals that we do not understand sufficiently well. For example, when considering signals from black hole coalescences, if the the total mass of the binary is sufficiently high, the binary will merge in band. The higher the total mass, the fewer the inspiral cycles that will be in band, until eventually only the merger is in band. Since the merger phase is the least understood phase, it stands to reason that our ability to test GR will deteriorate as the total mass increases. Of course, we do understand the ringdown phase very well, and tests of the nohair theorem would be allowed during this phase, provided a sufficientlyhigh SNR [65]. On the other hand, for neutron star binaries or verylowmass blackhole binaries, the merger phase is expected to be essentially out of band for aLIGO (above 1 kHz), and thus, the noise spectrum itself may shield us from our ignorance.
Instrumental systematics are introduced by our ignorance of the transfer function, which connects the detector output to the incoming gravitational waves. Through sophisticated calibration studies with real data, one can approximate the transfer function very well [4, 1]. However, this function is not timeindependent, because the noise in the instrument is not stationary or Gaussian. Thus, unmodeled drifts in the transfer function can introduce systematics in parameter estimation that are as large as 10% in the amplitude and the phase [4].
Instrumental systematics can affect tests of GR, if these are performed with a single instrument. However, one expects multiple detectors to be online in the future and for gravitationalwave detections to be made in several of them simultaneously. Instrumental systematics should be present in all such detections, but since the noise will be mostly uncorrelated between different instruments, one should be able to ameliorate its effects through crosscorrelating outputs from several instruments.
Astrophysical systematics are induced by our lack of a priori knowledge of the gravitational wave source. As explained above, matched filtering requires knowledge of a waveform template with which to filter the data. Usually, we assume the sources are in a perfect vacuum and isolated. For example, when considering inspiral signals, we ignore any third bodies, electric or magnetic fields, neutron star hydrodynamics, the expansion of the universe, etc. Fortunately, however, most of these effects are expected to be small: the probability of finding third bodies sufficiently close to a binary system is very small [463]; for low redshift events, the expansion of the universe induces an acceleration of the center of mass, which is also very small [468]; electromagnetic fields and neutronstar hydrodynamic effects may affect the inspiral of black holes and neutron stars, but not until the very last stages, when most signals will be out of band anyways. For example, tidal deformation effects enter a neutronstarbinary inspiral waveform at 5 postNewtonian order, which therefore affects the signal outside of the most sensitive part of the aLIGO sensitivity bucket.
Perhaps the most dangerous source of astrophysical systematics is due to the assumptions made about the astrophysical systems we expect to observe. For example, when considering neutronstarbinary inspirals, one usually assumes the orbit will have circularized by the time it enters the sensitivity band. Moreover, one assumes that any residual spin angular momentum that the neutron stars may possess is very small and aligned or counteraligned with the orbital angular momentum. These assumptions certainly simplify the construction of waveform templates, but if they happen to be wrong, they would introduce mismodeling systematics that could also affect parameter estimation and tests of GR.
4.2 Burst analyses
In alternative theories of gravity, gravitationalwave sources such as core collapse supernovae may result in the production of gravitational waves in more than just the plus and crosspolarizations [384, 380, 216, 334, 333, 369]. Indeed, the nearspherical geometry of the collapse can be a source of scalar breathingmode gravitational waves. However, the precise form of the waveform is unknown because it is sensitive to the initial conditions.
When searching for unmodeled bursts in alternative theories of gravity, a general approach involves optimized linear combinations of data streams from all available detectors to form maximum likelihood estimates of the waveforms in the various polarizations, and the use of null streams. In the context of groundbased detectors and GR, these ideas were first explored by Gürsel and Tinto [212] and later by Chatterji et al. [101] with the aim of separating falsealarm events from real detections. The main idea was to construct a linear combination of data streams received by a network of detectors, so that the combination contained only noise. Of course, in GR one need only include h_{+}. and h_{×} polarizations, and thus a network of three detectors suffices. This concept can be extended to develop null tests of GR, as originally proposed by Chatziioannou et al. [102] and recently implemented by Hayama et al. [228].
For pulsar timing experiments where one is dealing with data streams of about a few tens of pulsars, waveform reconstruction for all polarization states, as well as numerous null streams, can be constructed.
4.3 Stochastic background searches
Much work has been done on the response of groundbased interferometers to nontensorial polarization modes, stochastic background detection prospects, and data analysis techniques [299, 323, 191, 329, 121]. In the context of pulsar timing, the first work to deal with the detection of such backgrounds in the context of alternative theories of gravity is due to Lee et al. [284], who used a coherence statistic approach to the detection of nonEinsteinian polarizations. They studied the number of pulsars required to detect the various extra polarization modes, and found that pulsar timing arrays are especially sensitive to the longitudinal mode. Alves and Tinto [22] also found enhanced sensitivity to longitudinal and vector modes. Here we follow the work in [329, 99] that deals with the LIGO and pulsar timing cases using the optimal statistic, a crosscorrelation that maximizes the SNR.
In order to perform a search for a given polarization mode one first needs to compute the overlap reduction functions (using either Eq. (120) or (121)) for that mode. With that in hand and a form for the stochastic background spectrum Ω_{ A }(f), one can construct optimal filters for all pairs in the detector network using Eq. (119), and perform the crosscorrelations using either Eq. (109) (or equivalently Eq. (111)). Finally, we can calculate the overall network statistic Eq. (124), by first finding the variances using Eq. (114).
It is important to point out that the procedure outlined above is straightforward for groundbased interferometers. However, pulsar timing data are irregularly sampled, and have a pulsartiming model subtracted out. This needs to be accounted for, and generally, a timedomain approach is more appropriate for these data sets. The procedure is similar to what we have outlined above, but power spectra and gravitationalwave spectra in the frequency domain need to be replaced by autocovariance and crosscovariance matrices in the time domain that account for the model fitting (for an example of how to do this see [162]).
Interestingly, Nishizawa et al. [329] show that with three spatiallyseparated detectors the tensor, vector, and scalar contributions to the energy density in gravitational waves can be measured independently. Lee et al. [284] and Alves and Tinto [22] showed that pulsar timing experiments are especially sensitive to the longitudinal mode, and to a lesser extent the vector modes. Chamberlin and Siemens [99] showed that the sensitivity of the crosscorrelation to the longitudinal mode using nearby pulsar pairs can be enhanced significantly compared to that of the transverse modes. For example, for the NANOGrav pulsar timing array, two pulsar pairs separated by 3° result in an enhancement of 4 orders of magnitude in sensitivity to the longitudinal mode relative to the transverse modes. The main contribution to this effect is due to gravitational waves that are coming from roughly the same direction as the pulses from the pulsars. In this case, the induced redshift for any gravitationalwave polarization mode is proportional to fL, the product of the gravitationalwave frequency and the distance to the pulsar, which can be large. When the gravitational waves and the pulse direction are exactly parallel, the redshift for the transverse and vector modes vanishes, but it is proportional to f L for the scalarlongitudinal mode.
Lee et al. [285] studied the detectability of massive gravitons in pulsar timing arrays through stochastic background searches. They introduced a modification to Eq. (59) to account for graviton dispersion, and found the modified overlap reduction functions (i.e., modifications to the HellingsDowns curves Eq. (122)) for various values of the graviton mass. They conclude that a large number of stable pulsars (≥ 60) are required to distinguish between the massive and massless cases, and that future pulsar timing experiments could be sensitive to graviton masses of about 10−^{22} eV (∼ 10^{13} km). This method is competitive with some of the compact binary tests described later in Section 5.3.1 (see Table 2). In addition, since the method of Lee et al. [285] only depends on the form of the overlap reduction functions, it is advantageous in that it does not involve matched filtering (and therefore prior knowledge of the waveforms), and generally makes few assumptions about the gravitationalwave source properties.
5 Compact Binary Tests
In this section, we discuss gravitational wave tests of GR with signals emitted by compact binary systems. We begin by explaining the difference between direct and generic tests. We then proceed to describe the many direct or topdown tests and generic or bottomup tests that have been proposed once gravitational waves are detected, including tests of the nohair theorems. We concentrate here only on binaries composed of compact objects, such as neutron stars, black holes or other compact exotica. We will not discuss tests one could carry out with electromagnetic information from binary (or double) pulsars, as these are already described in [438]. We will also not review tests of GR with accretion disk observations, for which we refer the interested reader to [359].
5.1 Direct and generic tests
Gravitationalwave tests of Einstein’s theory can be classed into two distinct subgroups: direct tests and generic tests. Direct tests employ a topdown approach, where one starts from a particular modified gravity theory with a known action, derives the modified field equations and solves them for a particular gravitational waveemitting system. On the other hand, generic tests adopt a bottomup approach, where one takes a particular feature of GR and asks what type of signature its absence would leave on the gravitationalwave observable; one then asks whether the data presents a statisticallysignificant anomaly pointing to that particular signature.
Direct tests have by far been the traditional approach to testing GR with gravitational waves. The prototypical examples here are tests of JordanFierzBransDicke theory. As described in Section 2, one can solve the modified field equations for a binary system in the postNewtonian approximation to find a prediction for the gravitationalwave observable, as we will see in more detail later in this section. Other examples of direct tests include those concerning modified quadratic gravity models and noncommutative geometry theories.
The main advantage of such direct tests is also its main disadvantage: one has to pick a particular modified gravity theory. Because of this, one has a welldefined set of field equations that one can solve, but at the same time, one can only make predictions about that modified gravity model. Unfortunately, we currently lack a particular modified gravity theory that is particularly compelling; many modified gravity theories exist, but none possess all the criteria described in Section 2, except perhaps for the subclass of scalartensor theories with spontaneous scalarization. Lacking a clear alternative to GR, it is not obvious which theory one should pick. Given that the full development (from the action to the gravitational wave observable) of any particular theory can be incredibly difficult, time and computationally consuming, carrying out direct tests of all possible modified gravity models once gravitational waves are detected is clearly unfeasible.
Given this, one is led to generic tests of GR, where one asks how the absence of specific features contained in GR could impact the gravitational wave observable. For example, one can ask how such an observable would be modified if the graviton had a mass, if the gravitational interaction were Lorentz or parity violating, or if there existed large extra dimensions. From these general considerations, one can then construct a “meta”observable, i.e., one that does not belong to a particular theory, but that interpolates over all known possibilities in a welldefined way. This model has come to be known as the parameterized postEinsteinian framework, in analogy to the parameterized postNewtonian scheme used to test GR in the solar system [438]. Given such a construction, one can then ask whether the data points to a statisticallysignificant deviation from GR.
The main advantage of generic tests is precisely that one does not have to specify a particular model, but instead one lets the data select whether it contains any statisticallysignificant deviations from our canonical beliefs. Such an approach is, of course, not new to physics, having most recently been successfully employed by the WMAP team [57]. The intrinsic disadvantage of this method is that, if a deviation is found, there is no onetoone mapping between it and a particular action, but instead one has to point to a class of possible models. Of course, such a disadvantage is not that limiting, since it would provide strong hints as to what type of symmetries or properties of GR would have to be violated in an ultraviolet completion of Einstein’s theory.
5.2 Direct tests
5.2.1 Scalartensor theories
What is the sensitivity of black holes in generic scalartensor theories? Will and Zaglauer [474] have argued that the nohair theorems require s_{ a } = 1/2 for all black holes, no matter what their mass or spin is. As already explained in Section 2, stationary black holes that are the byproduct of gravitational collapse (i.e., with matter that satisfies the energy conditions) in a general class of scalartensor theories are identical to their GR counterparts [224, 408, 159, 398].^{9} This is because the scalar field satisfies a free wave equation in vacuum, which forces the scalar field to be constant in the exterior of a stationary, asymptoticallyflat spacetime, provided one neglects a homogeneous, cosmological solution. If the scalar field is to be constant, then by Eq. (127), s_{ a } = 1/2 for a single blackhole spacetime.
Such an argument formally applies only to stationary scenarios, so one might wonder whether a similar argument holds for binary systems that are in a quasistationary arrangement. Will and Zaglauer [474] and Mirshekari and Will [315] extended this discussion to quasistationary spacetimes describing blackhole binaries to higher postNewtonian order. They argued that the only possible deviations from ψ = 0 are due to tidal deformations of the horizon due to the companion, which are known to arise at very high order in postNewtonian theory, \(\psi = \mathcal{O}[{({m_a}/{r_a})^5}]\). Recently, Yunes et al. [465] extended this argument further by showing that to all orders in postNewtonian theory, but in the extreme massratio limit, black holes cannot have scalar hair in generic scalartensor theories. Finally, Healy et al. [230] have carried out a full numerical simulation of the nonlinear field equations, confirming this argument in the full nonlinear regime.
The activation of dynamics in the scalar field for a vacuum spacetime requires either a nonconstant distribution of initial scalar field (violating the constant cosmological scalar field condition at spatial infinity) or a pure geometrical source to the scalar field evolution equation. The latter would lead to the quadratic modified gravity theories discussed in Section 2.3.3. As for the former, Horbatsch and Burgess [235] have argued that if, for example, one lets ψ = μt, which clearly satisfies □ψ = 0 in a Minkowski background,^{10} then a Schwarzschild black hole will acquire modifications that are proportional to μ. Alternatively, scalar hair could also be induced by spatial gradients in the scalar field [67], possibly anchored in matter at galactic scales. Such cosmological hair, however, is likely to be suppressed by a long time scale; in the example above μ must have units of inverse time, and if it is to be associated with the expansion of the universe, then it would be natural to assume \(\mu = \mathcal{O}(H)\), where H is the Hubble parameter. Therefore, although such cosmological hair might have an effect on black holes in the early universe, it should not affect black hole observations at moderate to low redshifts.
Scalar field dynamics can be activated in nonvacuum spacetimes, even if initially the stars are not scalarized provided one considers a more general scalartensor theory, like the one introduced by Damour and EspositoFarèse [129, 130]. As discussed in Section 2.3.1, when the conformal factor takes on a particular functional form, nonlinear effects induced when the gravitational energy exceeds a certain threshold can spontaneously scalarize merging neutron stars, as demonstrated recently by Barausse, et al [51]. Therefore, neutron stars in binaries are likely to have hair in generic scalartensor theories, even if they start their inspiral unscalarized.
Many studies have been carried out to determine the level at which such corrections to the waveform could be measured or constrained once a gravitational wave has been detected. The first such study was carried out by Will [436], who determined that given a LIGO detection at SNR ρ = 10 of a (1.4, 3) M_{⊙} blackhole/neutronstar nonspinning, quasicircular inspiral, one could constrain ω_{BD} > 10^{3}. Scharre and Will [379] carried out a similar analysis but for a LISA detection with ρ = 10 of a (1.4, 10^{3}) M_{⊙} intermediatemass blackhole/neutronstar, nonspinning, quasicircular inspiral, and found that one could constrain ω_{BD} > 2.1 × 10^{4}. Such an analysis was then repeated by Will and Yunes [440] but as a function of the classic LISA instrument. They found that the bound is independent of the LISA arm length, but inversely proportional to the LISA position noise error, if the position error noise dominates over laser shot noise. All such studies considered an angleaveraged signal that neglected the spin of either body, assumptions that were relaxed by Berti et al. [63, 64]. They carried out MonteCarlo simulations over all signal sky positions that included spinorbit precession to find that the projected bound with LISA deteriorates to ω_{BD} > 0.7 × 10^{4} for the same system and SNR. This was confirmed and extended by Yagi et al. [450], who in addition to spinorbit precession allowed for noncircular (eccentric) inspirals. In fact, when eccentricity is included, the bound deteriorates even further to ω_{BD} > 0.5 × 10^{4}. The same authors also found that similar gravitationalwave observations with the nextgeneration detector DECIGO could constrain ω_{BD} > 1.6 × 10^{6}. Similarly, for a nonspinning neutronstar/blackhole binary, the future groundbased detector, the Einstein Telescope (ET) [361], could place constraints about 5 times stronger than the Cassini bound, as shown in [38].
Comparison of proposed tests of scalartensor theories.
Reference  Binary mass  ω_{BD}[10^{4}]  Properties 

[73]  X  4  Solar system 
[436]  (1.4, 3) M_{⊙}  0.1  LIGO, Fisher, Ang. Ave. circular, nonspinning 
[379]  (1.4,10^{3}) M_{⊙}  24  LISA, Fisher, Ang. Ave. circular, nonspinning 
[440]  (1.4,10^{3}) M_{⊙}  20  LISA, Fisher, Ang. Ave. circular, nonspinning 
[63]  (1.4,10^{3}) M_{⊙}  0.7  LISA, Fisher, MonteCarlo circular, w/spinorbit 
[450]  (1.4,10^{3}) M_{⊙}  0.5  LISA, Fisher, MonteCarlo eccentric, spinorbit 
[451]  (1.4,10) M_{⊙}  160  DECIGO, Fisher, MonteCarlo eccentric, spinorbit 
[38]  (1.4,10) M_{⊙}  10  ET, Fisher, Ang. Ave. circular, nonspinning 
The main reason that solarsystem constraints of JordanFierzBransDicke theory cannot be beaten with gravitationalwave observations is that the former are particularly wellsuited to constrain weakfield deviations of GR. One might have thought that scalartensor theories constitute strongfield tests of Einstein’s theory, but this is not quite true, as argued in Section 2.3.1. One can see this clearly by noting that scalartensor theory predicts dipolar radiation, which dominates at low velocities over the GR prediction (precisely the opposite behavior that one would expect from a strongfield modification to Einstein’s theory).
However, one should note that all the above analysis considered only the inspiral phase of coalescence, usually truncating their study at the innermost stablecircular orbit. The merger and ringdown phases, where most of the gravitational wave power resides, have so far been mostly neglected. One might expect that an increase in power will be accompanied by an increase in SNR, thus allowing us to constrain ω_{BD} further, as this scales with 1/SNR [262]. Moreover, during merger and ringdown, dynamical strongfield gravity effects in scalartensor theories could affect neutron star parameters and their oscillations [395], as well as possibly induce spontaneous scalarization [51]. All of these nonlinear effects could easily lead to a strengthening of projected bounds. However, to date no detailed analysis has attempted to determine how well one could constrain scalartensor theories using full information about the entire coalescence of a compact binary.
The subclass of scalartensor models described by JordanFierzBransDicke theory is not the only type of model that can be constrained with gravitationalwave observations. In the extrememassratio limit, for binaries consisting of a stellarmass compact object spiraling into a supermassive black hole, Yunes et al. [465] have recently shown that generic scalartensor theories reduce to either massless or massive JordanFierzBransDicke theory. Of course, in this case the sensitivities need to be calculated from the equations of structure within the full scalartensor theory. The inclusion of a scalar field mass leads to an interesting possibility: floating orbits [94]. Such orbits arise when the small compact object experiences superradiance, leading to resonances in the scalar flux that can momentarily counteract the gravitationalwave flux, leading to a temporarilystalled orbit that greatly modifies the orbitalphase evolution. These authors showed that if an extreme massratio inspiral is detected with a template consistent with GR, this alone allows us to rule out a large region of (m_{ s },ω_{ BD }) phase space, where m_{ s } is the mass of the scalar (see Figure 1 in [465]). This is because if such an inspiral had gone through a resonance, a GR template would be grossly different from the signal. Such bounds are dramatically stronger than the current most stringent bound ω_{BD} > 4 × 10^{4} and m_{ s } < 2.5 × 10^{−20} eV obtained from Cassini measurements of the Shapiro timedelay in the solar system [20]. Even if resonances are not hit, Berti et al. [71] have estimated that secondgeneration groundbased interferometers could constrain the combination m_{ s }/(ω_{BD})^{1}/^{2} ≲ 10^{−15} eV with the observation of gravitational waves from neutronstar/binary inspirals at an SNR of 10. These bounds can also be stronger than current constraints, especially for large scalar mass.
Another interesting scalartensor theory to consider is that studied by Damour and EspositoFarèse [129, 130]. As explained in Section 2.3.1, this theory is defined by the action of Eq. (14) with the conformal factor \(A(\psi) = {e^{\beta {\psi ^2}}}\). In standard BransDicke theory, only mixed binaries composed of a black hole and a neutron star lead to large deviations from GR due to dipolar emission. This is because dipole emission is proportional to the difference in sensitivities of the binary components. For neutronstar binaries with similar masses, this difference is close to zero, while for black holes it is identically zero (see Eqs. (134) and (144)). However, in the theory considered by Damour and EspositoFarèse, when the gravitational energy is large enough, as in the very late inspiral, nonlinear effects can lead to drastic modifications from the GR expectation, such as spontaneous scalarization [51]. Unfortunately, most of this happens at rather high frequency, and thus, it is not clear whether such effects are observable with current groundbased detectors.
5.2.2 Modified quadratic gravity
Black holes exist in the classes of modified quadratic gravity that have so far been considered. In nondynamical theories (when β = 0 and the scalarfields are constant, refer to Eq. (25)), Stein and Yunes [473] have shown that all metrics that are Ricci tensor flat are also solutions to the modified field equations (see also [360]). This is not so for dynamical theories, since then the ϑ field is sourced by curvature, leading to corrections to the field equations proportional to the Riemann tensor and its dual.
Neutron stars also exist in quadratic modified gravity. In dynamical ChernSimons gravity, the massradius relation remains unmodified to first order in the slowrotation expansion, but the moment of inertia changes to this order [469, 19], while the quadrupole moment and the mass measured at spatial infinity change to quadratic order in spin [448]. This is because the massradius relation, to first order in slowrotation, depends on the sphericallysymmetric part of the metric, which is unmodified in dynamical ChernSimons gravity. In EinsteinDilatonGaussBonnet gravity, the massradius relation is modified [342]. As in GR, these functions must be solved for numerically and they depend on the equation of state.
Gravitational waves are also modified in quadratic modified gravity. In dynamical ChernSimons gravity, Garfinkle et al. [190] have shown that the propagation of such waves on a Minkowski background remains unaltered, and thus, all modifications arise during the generation stage. In EinsteinDilatonGaussBonnet theory, no such analysis of the propagation of gravitational waves has yet been carried out. Yagi et al. [447] studied the generation mechanism in both theories during the quasicircular inspiral of comparablemass, spinning black holes in the postNewtonian and smallcoupling approximations. They found that a standard postNewtonian analysis fails for such theories because the assumption that black holes can be described by a distributional stressenergy tensor without any further structure fails. They also found that since black holes acquire scalar hair in these theories, and this scalar field is anchored to the curvature profiles, as black holes move, the scalar fields must follow the singularities, leading to dipole scalarfield emission.
From the above analysis, it should be clear that the corrections to the gravitationalwave observable in quadratic modified gravity are always proportional to the quantity \({\zeta _{3,4}} \equiv {\xi _{3,4}}/{m^4} = \alpha _{3,4}^2/(\beta \kappa {m^4})\). Thus, any measurement that is consistent with GR will allow a constraint of the form ζ_{3,4} < Nδ, where N is a number of order unity, and δ is the accuracy of the measurement. Solving for the coupling constants of the theory, such a measurement would lead to \(\xi _{3,4}^{1/4} < {(N\delta)^{1/4}}m\) [390]. Therefore, constraints on quadratic modified gravity will weaken for systems with larger characteristic mass. This can be understood by noticing that the corrections to the action scale with positive powers of the Riemann tensor, while this scales inversely with the mass of the object, i.e., the smaller a compact object is, the larger its curvature. Such an analysis then automatically predicts that LIGO will be able to place stronger constraints than LISAlike missions on such theories, because LIGO operates in the 100 Hz frequency band, allowing for the detection of stellarmass inspirals, while LISAlike missions operate in the mHz band, and are limited to supermassive blackholes inspirals.
How well can these modifications be measured with gravitationalwave observations? Yagi et al. [447] predicted, based on the results of Cornish et al. [124], that a skyaveraged LIGO gravitationalwave observation with SNR of 10 of the quasicircular inspiral of nonspinning black holes with masses (6,12)M_{⊙} would allow a constraint of \(\xi _3^{1/4} \underset{\sim}{<}20\) km, where we recall that \({\xi _3} = \alpha _3^2/(\beta \kappa)\). A similar skyaveraged, eLISA observation of a quasicircular, spinaligned blackhole inspiral with masses (10^{6},3 × 10^{6}M_{⊙}) would constrain \(\xi _3^{1/4} < {10^7}\) km [447]. The loss in constraining power comes from the fact that the constraint on ξ_{3} will scale with the total mass of the binary, which is six orders of magnitude larger for spaceborne sources. These constraints are not stronger than current bounds from the existence of compact objects [342] (ξ_{3} < 26 km) and from the change in the orbital period of the lowmass xray binary A0620–00 (ξ_{3} < 1.9 km) [444], but they are independent of the nature of the object and sample the theory in a different energy scale. In dynamical ChernSimons gravity, one expects similar projected gravitationalwave constraints on ξ_{4}, namely \(\xi _4^{1/4} < \mathcal{O}(M)\), where M is the total mass of the binary system in kilometers. Therefore, for binaries detectable with groundbased interferometers, one expects constraints of order \(\xi _4^{1/4} < 10\) km. In this case, such a constraint would be roughly six orders of magnitude stronger than current LAGEOS bounds [19]. Dynamical ChernSimons gravity cannot be constrained with binary pulsar observations, since the theory’s corrections to the postKeplerian observables are too high postNewtonian order, given the current observational uncertainties [448]. However, the gravitational wave constraint is more difficult to achieve in the dynamical ChernSimons case, because the correction to the gravitational wave phase is degenerate with spin. However, Yagi et al. [454] argued that precession should break this degeneracy, and if a signal with sufficiently high SNR is observed, such bounds would be possible. One must be careful, of course, to check that the smallcoupling approximation is still satisfied when saturating such a constraint [454].
5.2.3 Noncommutative geometry
Black holes exist in noncommutative geometry theories, as discussed in Section 2.3.5. What is more, the usual Schwarzschild and Kerr solutions of GR persist in these theories. This is not because such solutions have vanishing Weyl tensor, but because the quantity ∇^{ αβ }C_{ μανβ } happens to vanish for such metrics. Similarly, one would expect that the twobody, postNewtonian metric that describes a blackholebinary system should also satisfy the noncommutative geometry field equations, although this has not been proven explicitly. Similarly, although neutronstar spacetimes have not yet been considered in noncommutative geometries, it is likely that if such spacetimes are stationary and satisfy the Einstein equations, they will also satisfy the modified field equations. Much more work on this is still needed to establish all of these concepts on a firmer basis.
The above analysis was used by Nelson et al. [326, 325] to compute the rate of change of the orbital period of binary pulsars, in the hopes of using this to constrain β. Using data from the binary pulsar, they stipulated an orderofmagnitude constraint of β ≥ 10^{−13} m^{−1}. However, such an analysis could be revisited to relax a few assumptions used in [326, 325]. First, binary pulsar constraints on modified gravity theories require the use of at least three observables. These observables can be, for example, the rate of change of the period Ṗ, the line of nodes \(\dot \Omega\) and the perihelion shift ẇ. Any one observable depends on the parameters (m_{1}, m_{2}) in GR or (m_{1}, m_{2}, β) in noncommutative geometries, where m_{1,2} are the component masses. Therefore, each observable corresponds to a surface of codimension one, i.e., a twodimensional surface or sheet in the threedimensional space (m_{1}, m_{2}, β). If the binary pulsar observations are consistent with Einstein’s theory, then all sheets will intersect at some point, within a certain uncertainty volume given by the observational error. The simultaneous fitting of all these observables is what allows one to place a bound on β. The analysis of [326, 325] assumed that all binary pulsar observables were known, except for β, but degeneracies between (m_{1}, m_{2}, β) could potentially dilute constraints on these quantities. Moreover, this analysis should be generalized to eccentric and inclined binaries, since binary pulsars are known to not be on exactly circular orbits.
But perhaps the most important modification that ought to be made has to do with the calculation of the energy flux itself. The expression for Ė in Eq. (164) in terms of derivatives of the metric perturbation derives from the effective gravitationalwave stressenergy tensor, obtained by perturbatively expanding the action or the field equations and averaging over several wavelengths (the Isaacson procedure [241, 242]). In modified gravity theories, the definition of the effective stressenergy tensor in terms of the metric perturbation is usually modified, as found for example in [400]. In the case of noncommutative geometries, Stein and Yunes [400] showed that Eq. (164) still holds, provided one considers fluxes at spatial infinity. However, the analysis of [326, 325] evaluated this energy flux at a fixed distance, instead of taking the r → ∞ limit.
The balance law relates the rate of change of a binary’s binding energy with the gravitational wave flux emitted by the binary, but for it to hold, one must require the following: (i) that the binary be isolated and possess a welldefined binding energy; (ii) the total stressenergy of the spacetime satisfies a local covariant conservation law. If (ii) holds, one can use this conservation law to relate the rate of change of the volume integral of the energy density, i.e., the energy flux, to the volume integral of the current density, which can be rewritten as an integral over the boundary of the volume through Stokes’ theorem. Since in principle one can choose any integration volume, any physicallymeaningful result should be independent of the surface of that volume. This is indeed the case in GR, provided one takes the integration 2sphere to spatial infinity. Presumably, if one included all the relevant terms in Ė, without taking the limit to i^{0}, one would still find a result that is independent of the surface of this twosphere. However, this has not yet been verified. Therefore, the analysis of [326, 325] should be taken as an interesting first step toward understanding possible changes in the gravitationalwave metric perturbation in noncommutative geometries.
Not much beyond this has been done regarding noncommutative geometries and gravitational waves. In particular, one lacks a study of what the final response function would be if the gravitationalwave propagation were modified, which of course depends on the timeevolution of all propagating gravitationalwave degrees of freedom, and whether there are only the two usual dynamical degrees of freedom in the metric perturbation.
5.3 Generic tests
5.3.1 Massive graviton theories and Lorentz violation
Several massive graviton theories have been proposed to later be discarded due to ghosts, nonlinear or radiative instabilities. Thus, little work has gone into studying whether black holes and neutron stars in these theories persist and are stable, and how the generation of gravitational waves is modified. Such questions will depend on the specific massive gravity model considered, and of course, if a Vainshtein mechanism is employed, then there will not be any modifications.
Another generic consequence of a graviton mass is the appearance of additional propagating degrees of freedom in the gravitational wave metric perturbation. In particular, one expects scalar, longitudinal modes to be excited (see, e.g., [148]). This is, for example, the case if the action is of PauliFierz type [169, 148]. Such longitudinal modes arise due to the nonvanishing of the Ψ_{2} and Ψ_{3} NewmanPenrose scalars, and can be associated with the presence of spin0 particles, if the theory is of Type N in the E(2) classification [438]. The specific form of the scalar mode will depend on the structure of the modified field equations, and thus, it is not possible to generically predict its associated contribution to the response function.
Comparison of proposed tests of massive graviton theories. Ang. Ave. stands for an angular average over all sky locations.
Reference  Binary mass  λ_{ g }[10^{15} km]  Properties 

[404]  x  0.0028  Solarsystem dynamics 
[174]  x  1.6 × 10^{−5}  Binary pulsar orbital period in Visser’s theory [424] 
[88]  x  0.024  Stability of black holes in PauliFierz theory [169] 
[437]  (10, 10) M_{⊙}  0.006  LIGO, Fisher, Ang. Ave. circular, nonspinning 
[437]  (10^{7}, 10^{7}) M_{⊙}  69  LISA, Fisher, Ang. Ave. circular, nonspinning 
[281, 126]  (0.5, 0.5) M_{⊙}  0.03  LISA, WDWD, coincident with electromagnetic signal 
[440]  (10^{7}, 10^{7}) M_{⊙}  50  LISA, Fisher, Ang. Ave. circular, nonspinning 
[63]  (10^{6}, 10^{6})M_{⊙}  10  LISA, Fisher, MonteCarlo circular, w/spinorbit 
[39]  (10^{5},10^{5})M_{⊙}  10  LISA, Fisher, Ang. Ave. higherharmonics, circular, nonspinning 
[450]  (10^{6}, 10^{7})M_{⊙}  22  LISA, Fisher, MonteCarlo eccentric, spinorbit 
[451]  (10^{6}, 10^{7})M_{⊙}  2.4  DECIGO, Fisher, MonteCarlo eccentric, spinorbit 
[399]  (10^{6}, 10^{6})M_{⊙}  50  LISA, Fisher, MonteCarlo circular, w/spin modulations 
[262]  (10^{7}, 10^{7}) M_{⊙}  400  LISA, Fisher, Ang. Ave. circular, nonspinning, w/merger 
[142]  (13, 3)M_{⊙}  0.006–0.014  LIGO, Bayesian, Ang. Ave. circular, nonspinning 
[70]  (13, 3)M_{⊙}  30  eLISA, Fisher, MonteCarlo multiple detections, circular, nonspinning 
Before proceeding, we should note that the correction to the propagation of gravitational waves due to a nonzero graviton mass are not exclusive to binary systems. In fact, any gravitational wave that propagates a significant distance from the source will suffer from the time delays described in this section. Binary inspirals are particularly useful as probes of this effect because one knows the functional form of the waveform, and thus, one can employ matched filtering to obtain a strong constraint. But, in principle, one could use gravitationalwave bursts from supernovae or other sources.
Different α_{LV} limits deserve further discussion here. Of course, when α_{LV} = 0, one recovers the standard massive graviton result with the mapping \(\lambda _g^{ 2} \rightarrow \lambda _g^{ 2} + \lambda _A^{ 2}\). When α_{LV} = 2, the dispersion relation is identical to that in Eq. (23), but with a redefinition of the speed of light, and should thus be unobservable. Indeed, in this limit the correction to the Fourier phase in Eq. (176) becomes linear in frequency, and this is 100% degenerate with the time of coalescence parameter in the standard GR Fourier phase. Finally, relative to the standard GR terms that arise in the postNewtonian expansion of the Fourier phase, the new corrections are of (1 + 3α_{LV}/2) postNewtonian order. Then, if LIGO gravitationalwave observations were incapable of discerning between a 4 postNewtonian and a 5 postNewtonian waveform, then such observations would not be able to see the modified dispersion effect if α_{LV} > 2. Mirshekari et al. [316] confirmed this expectation with a Fisher analysis of nonspinning, comparablemass quasicircular inspirals. They found that for α_{LV} = 3, one can place very weak bounds on λ_{ A }, namely A < 10^{−7} eV^{−1} with a LIGO observation of a (1.4, 1.4) M_{⊙} neutron star inspiral, A < 0.2 eV^{−1} with an enhancedLISA or NGO observation of a (10^{5}, 10^{5}) M_{⊙} blackhole inspiral, assuming a SNR of 10 and 100 respectively. A word of caution is due here, though, as these analyses neglect any Lorentzviolating correction to the generation of gravitational waves, including the excitation of additional polarization modes. One would expect that the inclusion of such effects would only strengthen the bounds one could place on Lorentzviolating theories, but this must be done on a theory by theory basis.
5.3.2 Variable G theories and large extra dimensions
The lack of a particular Lagrangian associated with variable G theories, excluding scalartensor theories, or extra dimensions, makes it difficult to ascertain whether blackhole or neutronstar binaries exist in such theories. Whether this is so will depend on the particular variable G model considered. In spite of this, if such binaries do exist, the gravitational waves emitted by such systems will carry some generic modifications relative to the GR expectation.
Most current tests of the variability of Newton’s gravitational constant rely on electromagnetic observations of massive bodies, such as neutron stars. As discussed in Section 2.3.4, scalartensor theories can be interpreted as variableG theories, where the variability of G is really a variation in the coupling between gravity and matter. However, Newton’s constant serves the more fundamental role of defining the relationship between geometry or length and energy, and such a relationship is not altered in most scalartensor theories, unless the scalar fields are allowed to vary on a cosmological scale (background, homogeneous scalar solution).
For this reason, one might wish to consider a possible temporal variation of Newton’s constant in pure vacuum spacetimes, such as in blackholebinary inspirals. Such temporal variation would encode Ġ/G at the time and location of the merger event. Thus, once a sufficiently large number of gravitational wave events has been observed and found consistent with GR, one could reconstruct a constraint map that bounds Ġ/G along our past light cone (as a function of redshift and sky position). Since our pastlight cone with gravitational waves would have extended to roughly redshift 10 with classic LISA (limited by the existence of merger events at such high redshifts), such a constraint map would have been much more complete than what one can achieve with current tests at redshift almost zero. Big Bang nucleosynthesis constraints also allow us to bound a linear drift in Ġ/G from z ≫ 10^{3} to zero, but these become degenerate with limits on the number of relativistic species. Moreover, these bounds exploit the huge leverarm provided by integrating over cosmic time, but they are insensitive to local, oscillatory variations of G with periods much less than the cosmic observation time. Thus, gravitationalwave constraint maps would test one of the pillars of GR: local position invariance. This principle (encoded in the equivalence principle) states that the laws of physics (and thus the fundamental constants of nature) are the same everywhere in the universe.
Given such corrections to the gravitationalwave response function, one can investigate the level to which a gravitationalwave observation consistent with GR would allow us to constrain Ġ_{ c }. Yunes et al. [468] carried out such a study and found that for comparablemass blackhole inspirals of total redshifted mass m_{ z } = 10^{6} M_{⊙} with LISA, one could constrain Ġ_{ c }/G_{ c } ≲ 10^{−9} yr^{−1} or better to redshift 10 (assuming an SNR of 10^{3}). Similar constraints are possible with observations of extreme massratio inspirals. The constraint is strengthened when one considers intermediatemass blackhole inspirals, where one would be able to achieve a bound of Ġ_{ c }/G_{ c } ≲ 10^{−11} yr^{−1}. Although this is not as stringent as the strongest constraints from other observations (see Section 2.3.4), we recall that gravitationalwave constraints would measure local variations at the source, as opposed to local variations at zero redshift or integrated variations from the very early universe.
Given a gravitationalwave detection consistent with GR, one could then, in principle, place an upper bound on ℓ. Yagi et al. [449] carried out a Fisher analysis and found that a 1year LISA detection would constrain ℓ ≤ 10^{3} μm with a (10, 10^{5}) M_{Q} binary inspiral at an SNR of 100. This constraint is roughly two orders of magnitude weaker than current tabletop experiment constraints [7]. Moreover, the constraint weakens somewhat for more generic inspirals, due to degeneracies between ℓ and eccentricity and spin. However, a similar observation with the third generation detector DECIGO/BBO should be able to beat current constraints by roughly one order of magnitude. Such a constraint could be strengthened by roughly one order of magnitude further, if one included the statistical enhancement in parameter estimation due to detection of order 10^{5} sources by DECIGO/BBO.
Another way to place a constraint on ℓ is to consider the effect of mass loss in the orbital dynamics [308]. When a system loses mass, the evolution of its semimajor axis a will acquire a correction of the form \(\dot a =  (\dot M/M)a\), due to conservation of specific orbital angular momentum. There is then a critical semimajor axis a_{ c } at which this correction balances the semimajor decay rate due to gravitational wave emission. McWilliams [308] argues that systems with a < a_{ c } are then gravitationalwave dominated and will thus inspiral, while systems with a > a_{ c } will be massloss dominated and will thus outspiral. If a gravitational wave arising from an inspiraling binary is detected at a given semimajor axis, then ℓ is automatically constrained to about \(\mathcal{O}(20\mu {\rm{m}})\). Yagi et al. [449] extended this analysis to find that such a constraint is weaker than what one could achieve via matched filtering with a waveform in the form of Eq. (181), using the DECIGO detector.
A final possible degeneracy arises with the effect of a third body [463], accretion disk migration [267, 462] and the interaction of a binary with a circumbinary accretion disk [229]. All of these effects introduce corrections to the gravitationalwave phase of negative PN order, just like the effect of a variable gravitational constant. However, degeneracies of this type are only expected to affect a small subset of blackholebinary observations, namely those with a third body sufficiently close to the binary, or a sufficiently massive accretion disk.
5.3.3 Parity violation
As discussed in Section 2.3.6 the simplest action to model parity violation in the gravitational interaction is given in Eq. (45). Black holes and neutron stars exist in this theory, albeit nonrotating. A generic feature of this theory is that parity violation imprints onto the propagation of gravitational waves, an effect that has been dubbed amplitude birefringence. Such birefringence is not to be confused with optical or electromagnetic birefringence, in which the gauge boson interacts with a medium and is doublyrefracted into two separate rays. In amplitude birefringence, right(left)circularly polarized gravitational waves are enhanced or suppressed (suppressed or enhanced) relative to the GR expectation as they propagate [245, 295, 11, 460, 17, 464].
Another test of parity violation was proposed by Yunes et al. [464], who considered the coincident detection of a gravitational wave and a gammaray burst with the SWIFT [193] and GLAST/Fermi [97] gammaray satellites, and the groundbased LIGO [2] and Virgo [6] gravitational wave detectors. If the progenitor of the gammaray burst is a neutronstar/neutronstar merger, the gammaray jet is expected to be highly collimated. Therefore, an electromagnetic observation of such an event implies that the binary’s orbital angular momentum at merger must be pointing along the line of sight to Earth, leading to a stronglycircularlypolarized gravitationalwave signal and to maximal parity violation. If the gammaray burst observation were to provide an accurate sky location, one would be able to obtain an accurate distance measurement from the gravitational wave signal alone. Moreover, since GLAST/Fermi observations of gammaray bursts occur at low redshift, one would also possess a purely electromagnetic measurement of the distance to the source. Amplitude birefringence would manifest itself as a discrepancy between these two distance measurements. Therefore, if no discrepancy is found, the error ellipse on the distance measurement would allow us to place an upper limit on any possible gravitational parity violation. Because of the nature of such a test, one is constraining generic parity violation over distances of hundreds of Mpc, along the light cone on which the gravitational waves propagate.
The coincident gammaray burst/gravitationalwave test compares favorably to the pure LISA test, with the sensitivity to parity violation being about 2–3 orders of magnitude better in the former case. This is because, although the fractional error in the gravitationalwave distance measurement is much smaller for LISA than for LIGO, since it is inversely proportional to the SNR, the parity violating effect also depends on the gravitationalwave frequency, which is much larger for neutronstar inspirals than massive blackhole coalescences. Mathematically, the simplest models of gravitational parity violation will lead to a signature in the response function that is proportional to the gravitationalwave wavelength^{12} λ_{GW} ∝ Df. Although the coincident test requires small distances and low SNRs (by roughly 1–2 orders of magnitude), the frequency is also larger by a factor of 5–6 orders of magnitude for the LIGOVirgo network.
The coincident gammaray burst/gravitationalwave test also compares favorably to current solar system constraints. Using the motion of the LAGEOS satellites, Smith et al. [388] have placed the 1σ bound \({\dot \vartheta _0} < 2000\) km assuming \({\ddot \vartheta _0} = 0\). A similar assumption leads to a 2σ bound of \({\dot \vartheta _0} < 200\) km with a coincident gammaray burst/gravitationalwave observation. Moreover, the latter test also allows us to constrain the second timederivative of the scalar field. Finally, a LISA observation would constrain the integrated history of ϑ along the past light cone on which the gravitational wave propagated. However, these tests are not as stringent as the recently proposed test by Dyda et al. [158], \({\dot \vartheta _0} < {10^{ 7}}\) km, assuming the effective theory cutoff scale is less than 10 eV and obtained by demanding that the energy density in photons created by vacuum decay over the lifetime of the universe not violate observational bounds.
The coincident test is somewhat idealistic in that there are certain astrophysical uncertainties that could hamper the degree to which we could constrain parity violation. One of the most important uncertainties relates to our knowledge of the inclination angle, as gammaray burst jets are not necessarily perfectly aligned with the line of sight. If the inclination angle is not known a priori, it will become degenerate with the distance in the waveform template, decreasing the accuracy to which the luminosity could be extracted from a pure gravitational wave observation by at least a factor of two. Even after taking such uncertainties into account, Yunes et al. [464] found that \({\dot \vartheta _0}\) could be constrained much better with gravitational waves than with current solar system observations.
5.3.4 Parameterized postEinsteinian framework
One of the biggest disadvantages of a topdown or direct approach toward testing GR is that one must pick a particular theory from the beginning of the analysis. However, given the large number of possible modifications to Einstein’s theory and the lack of a particularly compelling alternative, it is entirely possible that none of these will represent the correct gravitational theory in the strong field. Thus, if one carries out a topdown approach, one will be forced to make the assumption that we, as theorists, know which modifications of gravity are possible and which are not [467]. The parameterized postEinsteinian (ppE) approach is a framework developed specifically to alleviate such a bias by allowing the data to select the correct theory of nature through the systematic study of statistically significant anomalies.
For detection purposes, one usually expects to use match filters that are consistent with GR. But if GR happened to be wrong in the strong field, it is possible that a GR template would still extract the signal, but with the wrong parameters. That is, the best fit parameters obtained from a matched filtering analysis with GR templates will be biased by the assumption that GR is sufficiently accurate to model the entire coalescence. This fundamental bias could lead to a highly distorted image of the gravitationalwave universe. In fact, recent work by Vallisneri and Yunes [417] indicates that such fundamental bias could indeed be present in observations of neutron star inspirals, if GR is not quite the right theory in the strongfield.
One of the primary motivations for the development of the ppE scheme was to alleviate fundamental bias, and one of its most dangerous incarnations: stealthbias [124]. If GR is not the right theory of nature, yet all our future detections are of low SNR, we may estimate the wrong parameters from a matchedfiltering analysis, yet without being able to identify that there is a nonGR anomaly in the data. Thus, stealth bias is nothing but fundamental bias hidden by our limited SNR observations. Vallisneri and Yunes [417] have found that such stealthbias is indeed possible in a certain sector of parameter space, inducing errors in parameter estimation that could be larger than statistical ones, without us being able to identify the presence of a nonGR anomaly.
5.3.4.1 Historical development
The template family in Eq. (204) allows for postNewtonian tests of GR, i.e., consistency checks of the signal with the postNewtonian expansion. For example, let us imagine that a gravitational wave has been detected with sufficient SNR that the chirp mass and mass ratio have been measured from the Newtonian and 1 postNewtonian terms in the waveform phase. One can then ask whether the 1.5 postNewtonian term in the phase is consistent with these values of chirp mass and mass ratio. Put another way, each term in the phase can be thought of as a curve in \(({M_c},\eta)\) space. If GR is correct, all these curves should intersect inside some uncertainty box, just like when one tests GR with binary pulsar data. From that standpoint, these tests can be thought of as nulltests of GR and one can ask: given an event, is the data consistent with the hypothesis β_{rppE} = 0 for the restricted set of frequency exponents b_{PN}?
A Fisher and a Bayesian data analysis study of how well β_{PNT} could be constrained given a certain b_{PN} was carried out in [317, 240, 290]. Mishra et al. [317] considered the quasicircular inspiral of nonspinning compact objects and showed that aLIGO observations would allow one to constrain β_{PNT} to 6% up to the 1.5 postNewtonian order correction (b_{PN} = −2). Thirdgeneration detectors, such as ET, should allow for better constraints on all postNewtonian coefficients to roughly 2%. Clearly, the higher the value of b_{PN}, the worse the bound on β_{PN} because the power contained in higher frequency exponent terms decreases, i.e., the number of useful additional cycles induced by the \({\beta _{PNT}}{u^{b{\rm{PN}}}}\) term decreases as b_{PN} increases. Huwyler et al. [240] repeated this analysis but for LISA observations of the quasicircular inspiral of black hole binaries with spin precession. They found that the inclusion of precessing spins forces one to introduce more parameters into the waveform, which dilutes information and weakens constraints on β_{PNT} by as much as a factor of 5. Li et al. [290] carried out a Bayesian analysis of the oddsratio between GR and restricted ppE templates given a nonspinning, quasicircular compact binary inspiral observation with aLIGO and adVirgo. They calculated the odds ratio for each value of b_{PN} listed above and then combined all of this into a single probability measure that allows one to quantify how likely the data is to be consistent with GR.
5.3.4.2 The simplest ppE model
Parameters that define the deformation of the response function in a variety of modified gravity theories. The notation · means that a value for this parameter is irrelevant, as its amplitude is zero.
Theory  α _{ppE}  a _{ppE}  β _{ppE}  b _{ppE} 

JordanFierzBransDicke  \( {5 \over {96}}{{{S^2}} \over {{\omega _{{\rm{BD}}}}}}{\eta ^{2/5}}\)  −2  \( {5 \over {3584}}{{{S^2}} \over {{\omega _{{\rm{BD}}}}}}{\eta ^{2/5}}\)  −7 
Dissipative EinsteinDilatonGaussBonnet Gravity  0  ·  \( {5 \over {7168}}{\varsigma _3}{\eta ^{ 18/5}}\delta _m^2\)  −7 
Massive Graviton  0  ·  \( {{{\pi ^2}D{{\mathcal M}_c}} \over {\lambda _g^2(1 + z)}}\)  −3 
Lorentz Violation  0  ·  \( {{{\pi ^{2  {\gamma _{{\rm{LV}}}}}}} \over {(1  {\gamma _{{\rm{LV}}}})}}{{{D_{{\gamma _{{\rm{LV}}}}}}} \over {\lambda _{{\rm{LV}}}^{2  {\gamma _{{\rm{LV}}}}}}}{{{\mathcal M}_c^{1  {\gamma _{{\rm{LV}}}}}} \over {{{(1 + z)}^{1  {\gamma _{{\rm{LV}}}}}}}}\)  −3γ_{LV} − 3 
G(t) Theory  \( {5 \over {512}}\dot G{{\mathcal M}_c}\)  −8  \( {{25} \over {65536}}{{\dot G}_c}{{\mathcal M}_c}\)  −13 
Extra Dimensions  ·  ·  \( {{75} \over {2554344}}{{dM} \over {dt}}{\eta ^{ 4}}(3  26\eta + 24{\eta ^2})\)  −13 
NonDynamical ChernSimons Gravity  α _{PV}  3  β _{PV}  6 
Dynamical ChernSimons Gravity  0  ·  β _{dCS}  −1 
In Table 3, recall that S is the difference in the square of the sensitivities and ω_{BD} is the BransDicke coupling parameter (see Section 5.2.1; we have here neglected the scalar mode), ζ_{3} is the coupling parameter in EinsteinDilatonGaussBonnet theory (see Section 5.2.2), where we have here included both the dissipative and the conservative corrections, D is a certain distance measure and λ_{ g } is the Compton wavelength of the graviton (see Section 5.3.1), λ_{LV} is a distance scale at which Lorentzviolation becomes important and γ_{LV} is the graviton momentum exponent in the deformation of the dispersion relation (see Section 5.3.1), Ġ_{ c } is the value of the time derivative of Newton’s constant at coalescence and dM/dt is the mass loss due to enhanced Hawking radiation in extradimensional scenarios (see Section 5.3.2), β_{dCS} is given in Eq. (157) and (α_{PV},β_{PV}) are given in Eqs. (198) and (199) of Section 5.3.3.
Although there are only a few modified gravity theories where the leadingorder postNewtonian correction to the Fourier transform of the response function can be parameterized by postNewtonian waveforms of Eq. (204), all such predictions can be modeled with the ppE templates of Eq. (205). In fact, only massive graviton theories, certain classes of Lorentzviolating theories and dynamical ChernSimons gravity lead to waveform corrections that can be parameterized via Eq. (204). For example, the lack of amplitude corrections in Eq. (204) does not allow for tests of gravitational parity violation or nondynamical ChernSimons gravity.
However, this does not imply that Eq. (205) can parameterize all possible deformations of GR. First, Eq. (205) can be understood as a singleparameter deformation away from Einstein’s theory. If the correct theory of nature happens to be a deformation of GR with several parameters (e.g., several coupling constants, mass terms, potentials, etc.), then Eq. (205) will only be able to parameterize the one that leads to the most useful cycles. This was recently verified by Sampson et al. [376]. Second, Eq. (205) assumes that the modification can be represented as a power series in velocity, with possibly noninteger values. Such an assumption does not allow for possible logarithmic terms, which are known to arise due to nonlinear memory interactions at sufficientlyhigh postNewtonian order. It also does not allow for interactions that are screened, e.g., in theories with massive degrees of freedom. Nonetheless, the parameterization in Eq. (205) will still be able to signal that the detection is not a pure Einstein event, at the cost of biasing their true value.
5.3.4.3 More complex ppE models
In fact, this is precisely one of the most important differences between the ppE and ppN frameworks. In ppN, it does not matter how many ppN parameters are introduced, because the observations are of very high SNR, and thus, templates are not needed to extract the signal from the noise. On the other hand, in gravitational wave astrophysics, templates are essential to make detections and do parameter estimation. Spurious parameters in these templates that are not needed to match the signal will deteriorate the accuracy to which all parameters can be measured because of an Occam penalty. Thus, in gravitational wave astrophysics and data analysis one wishes to minimize the number of theory parameters when testing GR [124, 376]. One must then find a balance between the number of additional theory parameters to introduce and the amount of bias contained in the templates.
At this junction, one must emphasize that frequency exponents in the amplitude and phase correction were above assumed to be integers, i.e., (a_{ppE}, b_{ppE},n) ∈ ℤ. This must be the case if these corrections arise due to modifications that can be represented as integer powers of the momenta or velocity. We are not aware of any theory that predicts corrections proportional to fractional powers of the velocity for circular inspirals. Moreover, one can show that theories that introduce noninteger powers of the velocity into the equations of motion will lead to issues with analyticity at zero velocity and a breakdown of uniqueness of solutions [102]. In spite of this, modified theories can introduce logarithmic terms, that for example enter at high postNewtonian order in GR due to nonlinear propagation effects (see, e.g., [75] and references therein). Moreover, certain modified gravity theories introduce screened modifications that become “active” only above a certain frequency. Such effects would be modeled through a Heaviside function, for example needed when dealing with massive BransDicke gravity [147, 94, 20, 465]. However, even these nonpolynomial injections would be detectable with the simplest ppE model. In essence, one finds similar results as if one were trying to fit a 3parameter injection with the simplest 1parameter ppE model [376].
Of course, one can also generalize the inspiral ppE waveform families to more general orbits, for example through the inclusion of spins aligned or counteraligned with the orbital angular momentum. More general inspirals would still lead to waveform families of the form of Eq. (205) or (209), but where the parameters (α_{ppE}, β_{ppE}) would now depend on the mass ratio, mass difference, and the spin parameters of the black holes. With a single detection, one cannot break the degeneracy in the ppE parameters and separately fit for its system parameter dependencies. However, given multiple detections one should be able to break such a degeneracy, at least to a certain degree [124]. Such breaking of degeneracies begins to become possible when the number of detections exceeds the number of additional parameters required to capture the physical parameter dependencies of (α_{ppE}, β_{ppE}).
The ppE theory parameters are now \(\vec \theta = ({b_{{\rm{ppE}}}},{\beta _{{\rm{ppE}}}},{k_{{\rm{ppE}}}},{\kappa _{{\rm{ppE}}}},{\gamma _{{\rm{ppE}}}},\Phi _c^{(1)})\). Of course, one may ignore (k_{ppE}, κ_{ppE}) altogether, if one wishes to ignore propagation effects. Such a parameterization recovers the predictions of JordanFierzBransDicke theory for a singledetector response function [102], as well as Arun’s analysis for generic dipole radiation [36].
One might worry that the corrections introduced by the ℓ = 1 harmonic, i.e., terms proportional to γ_{ppE} in Eq. (217), will be degenerate with postNewtonian corrections to the amplitude of the ℓ = 2 mode (not displayed in Eq. (217)). However, this is clearly not the case, as the latter scale as \({(\pi {\mathcal{M}_c}f)^{ 7/6 + n/3}}\) with n an integer greater than 0, while the ℓ = 1 mode is proportional to \({(\pi {M_c}f)^{ 3/2}}\), which would correspond to a (−0.5) postNewtonian order correction, i.e., n = − 1. On the other hand, the ppE amplitude corrections to the ℓ = 2 mode, i.e., terms proportional to β_{ppE} in the amplitude of Eq. (217), can be degenerate with such postNewtonian corrections when b_{ppE} is an integer greater than −4.
5.3.4.4 Applications of the ppE formalism
The two models in Eq. (205) and (209) answer different questions. The latter contains a stronger prior (that ppE frequency exponents be integers), and thus, it is ideal for fitting a particular set of theoretical models. On the other hand, Eq. (205) with continuous ppE frequency exponents allows one to search for generic deviations that are statistically significant, without imposing such theoretical priors. That is, if a deviation from GR is present, then Eq. (205) is more likely to be able to fit it, than Eq. (209). If one prioritizes the introduction of the least number of new parameters, Eq. (205) with (a_{ppE}, b_{ppE}) ∈ ℝ can still recover deviations from GR, even if the latter cannot be represented as a correction proportional to an integer power of velocity.
Given these ppE waveforms, how should they be used in a data analysis pipeline? The main idea behind the ppE framework is to match filter or perform Bayesian statistics with ppE enhanced template banks to allow the data to select the bestfit values of θ^{ a }. As discussed in [467, 124] and then later in [290], one might wish to first run detection searches with GR template banks, and then, once a signal has been found, do a Bayesian model selection analysis with ppE templates. The first such Bayesian analysis was carried out by Cornish et al. [124], who concluded that an aLIGO detection at SNR of 20 for a quasicircular, nonspinning blackhole inspiral would allow us to constrain α_{ppE} and β_{ppE} much better than existent constraints for sufficiently strongfield corrections, e.g., b_{ppE} > −5. This is because for lower values of the frequency exponents, the corrections to the waveform are weakfield and better constrained with binary pulsar observations [461]. The large statistical study of Li et al. [290] uses a reduced set of ppE waveforms and investigates our ability to detect deviations of GR when considering a catalogue of aLIGO/adVirgo detections. Of course, the disadvantage of such a pipeline is that it requires a first detection, and if the gravitational interaction is too different from GR’s prediction, it is possible that a search with GR templates might miss the signal all together; we deem this possibility to be less likely.
Interpretation of nonzero ppE parameters.
a _{ppE}  b _{ppE}  Interpretation 

1  ·  Parity violation 
−8  −13  Anomalous acceleration, Extra dimensions, Violation of position invariance 
·  7  Dipole gravitational radiation, Electric dipole scalar radiation 
·  3  Massive graviton propagation 
∝ spin  −1  Magnetic dipole scalar radiation, Quadrupole moment correction, Scalar dipole force 
Moreover, if a followup search is done with the ppE model in Eq. (209), one could infer whether the correction is one due to modifications to the generation or the propagation of gravitational waves. In this way, a nonzero ppE detection could inform theories of what type of GR modification is preferred by nature.
5.3.4.5 Degeneracies
However, much care must be taken to avoid confusing a ppE theory modification with some other systematic, such as an astrophysical, a mismodeling or an instrumental effect. Instrumental effects can be easily remedied by requiring that several instruments, with presumably unrelated instrumental systematics, independently derive a posterior probability for (α_{ppE}, β_{ppE}) that peaks away from zero. Astrophysical uncertainties can also be alleviated by requiring that different events lead to the same posteriors for ppE parameters (after breaking degeneracies with system parameters). However, astrophysically there are a limited number of scenarios that could lead to corrections in the waveforms that are large enough to interfere with these tests. For comparablemassratio inspirals, this is usually not a problem as the inertia of each binary component is too large for any astrophysical environment to affect the orbital trajectory [229]. Magnetohydrodynamic effects could affect the merger of neutronstar binaries, but this usually occurs outside of the sensitivity band of groundbased interferometers. However, in extrememassratio inspirals the small compact object can be easily nudged away by astrophysical effects, such as the presence of an accretion disk [462, 267] or a third supermassive black hole [463]. However, these astrophysical effects present the interesting feature that they correct the waveform in a form similar to Eq. (205) but with b_{ppE} < −5. This is because the larger the orbital separation, the stronger the perturbations of the astrophysical environment, either because the compact object gets closer to the third body or because it leaves the inner edge of the accretion disk and the disk density increases with separation. Such effects, however, are not likely to be present in all sources observed, as few extrememassratio inspirals are expected to be embedded in an accretion disk or sufficiently close to a third body (≲ 0.1 pc) for the latter to have an effect on the waveform.
Perhaps the most dangerous systematic is mismodeling, which is due to the use of approximation schemes when constructing waveform templates. For example, in the inspiral one uses the postNewtonian approximation series, expanding and truncating the waveform at a given power of orbital velocity. Moreover, neutron stars are usually modeled as testparticles (with a Dirac distributional density profile), when in reality they have a finite radius, which will depend on its equation of state. Such finitesize effects enter at 5 postNewtonian order (the effacement principle [227, 128]), but with a postNewtonian coefficient that can be rather large [320, 72, 175]. Ignorance of the postNewtonian series beyond 3 postNewtonian order can lead to systematics in the determination of physical parameters and possibly also to confusion when carrying out ppElike tests. Much more work is needed to determine the systems and SNRs for which such systematics are truly a problem.
5.3.5 Searching for nontensorial gravitationalwave polarizations
Another way to search for generic deviations from GR is to ask whether any gravitationalwave signal detected contains more than the two traditional polarizations expected in GR. A general approach to answer this question is through null streams, as discussed in Section 4.3. This concept was first studied by Gürsel and Tinto [212] and later by Chatterji et al. [101] with the aim to separate falsealarm events from real detections. Chatziioannou et al. [102] proposed the extension of the idea of null streams to develop null tests of GR, which was proposed using stochastic gravitational wave backgrounds in [329, 330] and recently implemented in [228] to reconstruct the independent polarization modes in timeseries data of a groundbased detector network.
Given a gravitationalwave detection, one can ask whether the data is consistent with two polarizations by constructing a null stream through the combination of data streams from 3 or more detectors. As explained in Section 4.3, such a null stream should be consistent with noise in GR, while it would present a systematic deviation from noise if the gravitational wave metric perturbation possessed more than two polarizations. Notice that such a test would not require a template; if one were parametrically constructed, such as in [102], more powerful null tests could be applied to such a null steam. In the future, we expect several gravitational wave detectors to be online: the two aLIGO ones in the United States, adVIRGO in Italy, LIGOIndia in India, and KAGRA in Japan. Given a gravitationalwave observation that is detected by all five detectors, one can then construct three enhanced GR null streams, each with power in a signal null direction.
5.3.6 ILoveQ tests
Neutron stars in the slowrotation limit can be characterized by their mass and radius (to zerothorder in spin), by their moment of inertia (to firstorder in spin), and by their quadrupole moment and Love numbers (to secondorder in spin). One may expect these quantities to be quite sensitive to the neutron star’s internal structure, which can be parameterized by its equation of state, i.e., the relation between its internal pressure and its internal energy density. Since the equation of state cannot be wellconstrained at supernuclear densities in the laboratory, one is left with a variety of possibilities that predict different neutronstar massradius relations.
Given the independent measurement of any two members of the ILoveQ trio, one could carry out a (null) modelindependent and equationofstateindependent test of GR [453, 452]. For example, assume that electromagnetic observations of the binary pulsar J07373039 have measured the moment of inertia to 10% accuracy [282, 273, 274]. The slowrotation approximation is perfectly valid for this binary pulsar, due to its relatively long spin period. Assume further that a gravitationalwave observation of a neutronstarbinary inspiral, with individual masses similar to that of the primary in J07373039, manages to measure the neutron star tidal Love number to 60% accuracy [453, 452]. These observations then lead to an error box in the ILove plane, which must contain the curve in the leftpanel of Figure 5.
A similar test could be carried out by using data from only binary pulsar observations or only gravitational wave detections. In the case of the latter, one would have to simultaneously measure or constrain the value of the quadrupole moment and the Love number, since the moment of inertia is not measurable with gravitational wave observations. In the case of the former, one would have to extract the moment of inertia and the quadrupole moment, the latter of which will be difficult to measure. Therefore, the combination of electromagnetic and gravitational wave observations would be the ideal way to carry out such tests.
Such a test of GR, of course, is powerful only as long as modified gravity theories predict ILoveQ relations that are not degenerated with the general relativistic ones. Yagi and Yunes [453, 452] investigated such a relation in dynamical ChernSimons gravity to find that such degeneracy is only present in the limit ζ_{CS} → 0. That is, for any finite value of ζ_{CS}, the dynamical ChernSimons ILoveQ relation differs from that of GR, with the distance to the GR expectation increasing for larger ζ_{CS}. Yagi and Yunes [453, 452] predicted that a test similar to the one described above could constrain dynamical ChernSimons gravity to roughly \(\xi _{{\rm{CS}}}^{1/4} < 10{M_{NS}}\sim 15\) km, where recall that \({\xi _{{\rm{CS}}}} = \alpha _{{\rm{CS}}}^2/(\beta \kappa)\).
The test described above, of course, only holds provided the ILoveQ relations are valid, which in turn depends on the assumptions made in deriving them. In particular, Yagi and Yunes [453, 452] assumed that the neutron stars are uniformly and slowly rotating, as well as only slightly tidally deformed by their rotational velocity or companion. These assumptions would not be valid for newlyborn neutron stars, which are probably differentially rotating and doing so quickly. However, the gravitational waves emitted by neutronstar inspirals are expected to have binary components that are old and not rapidly spinning by the time they enter the detector sensitivity band [74]. Some shortperiod, millisecond pulsars may spin at a nonnegligible rate, for which the normalized moment of inertia, quadrupole moment and Love number would not be independent of the rotational angular velocity. However, if then the above tests should still be possible, since binary pulsar observations would also automatically determine the rotational angular velocity, for which a unique ILoveQ relation should exist in GR.
5.4 Tests of the nohair theorems
Another important class of generic tests of GR are those that concern the nohair theorems. Since much work has been done on this area, we have decided to separate this topic from the main generic tests section (5.3). In what follows, we describe what these theorems are and the possible tests one could carry out with gravitationalwave observations emitted by blackholebinary systems.
5.4.1 The nohair theorems
The nohair theorems state that the only stationary, vacuum solution to the Einstein equations that is nonsingular outside the event horizon is completely characterized by three quantities: its mass M, its spin S and its charge Q. This conclusion is arrived at by combining several different theorems. First, Hawking [223, 222] proved that a stationary black hole must have an event horizon with a spherical topology and that it must be either static or axially symmetric. Israel [243, 244] then proved that the exterior gravitational field of such static black holes is uniquely determined by M and Q and it must be given by the Schwarzschild or the ReissnerNordström metrics. Carter [98] constructed a similar proof for uncharged, stationary, axiallysymmetric black holes, where this time black holes fall into disjoint families, not deformable into each other and with an exterior gravitational field uniquely determined by M and S. Robinson [363] and Mazur [306] later proved that such black holes must be described by either the Kerr or the KerrNewman metric. See also [318, 352] for more details.
The nohair theorems apply under a restrictive set of conditions. First, the theorems only apply in stationary situations. Blackhole horizons can be tidally deformed in dynamical situations, and if so, Hawking’s theorems [223, 222] about spherical horizon topologies do not apply. This then implies that all other theorems described above also do not apply, and thus, dynamical black holes will generically have hair. Second, the theorems only apply in vacuum. Consider, for example, an axiallysymmetric black hole in the presence of a nonsymmetrical matter distribution outside the event horizon. One might naively think that this would tidally distort the event horizon, leading to a rotating, stationary black hole that is not axisymmetric. However, Hawking and Hartle [226] showed that in such a case the matter distribution torques the black hole forcing it to spin down, thus leading to a nonstationary scenario. If the black hole is nonstationary, then again the nohair theorems do not apply by the arguments described at the beginning of this paragraph, and thus nonisolated black holes can have hair. Third, the theorems only apply within GR, i.e., through the use of the Einstein equations. Therefore, it is plausible that black holes in modified gravity theories or in GR with singularities outside any event horizons (naked singularities) will have hair.
An astrophysical observation of a hairy black hole would not imply that the nohair theorems are wrong, but rather that one of the assumptions made in deriving these theorems is not appropriate to describe nature. As described above, the three main assumptions are stationarity, vacuum and that GR and the regularity condition hold. Astrophysical black holes will generically be hairy due to a violation of the first two assumptions, since they will neither be perfectly stationary, nor exist in a perfect vacuum. Astrophysical black holes will always suffer small perturbations by other stars, electromagnetic fields, other forms of matter, like dust, plasma or dark matter, etc, which will induce nonzero deviations from Eq. (219) and thus evade the nohair theorems. However, in all cases of interest such perturbations are expected to be too small to be observable, which is why one argues that even astrophysical black holes should obey the nohair theorems if GR holds. Put another way, an observation of the violation of the nohair theorems would be more likely to indicate a failure of GR in the strongfield, than an unreasonably large amount of astrophysical hair.
Tests of the nohair theorems come in two flavors: through electromagnetic observations [250, 251, 253, 254] and through gravitational wave observations [370, 371, 112, 196, 44, 50, 289, 390, 471, 422, 421, 184, 423, 364]. The former rely on radiation emitted by accelerating particles in an accretion disk around black holes. However, such tests are not clean as they require the modeling of complicated astrophysics, with matter and electromagnetic fields. Gravitational wave tests are clean in that respect, but unlike electromagnetic tests, they cannot be carried out yet due to lack of data. Other electromagnetic tests of the nohair theorems exist, for example through the observation of close stellar orbits around Sgr A* [312, 313, 373] and pulsarblackhole binaries [431], but these cannot yet probe the nearhorizon, strongfield regime, since electromagnetic observations cannot yet resolve horizon scales. See [359] for reviews on this topic.
5.4.2 Extreme massratio tests of the nohair theorem
Gravitational wave tests of the nohair theorems require the detection of either extreme massratio inspirals or the ringdown of comparablemass blackhole mergers with future spaceborne gravitationalwave detectors [25, 24]. Extreme massratio inspirals consist of a stellarmass compact object spiraling into a supermassive black hole in a generic orbit within astronomical units from the event horizon of the supermassive object [23]. These events outlive the observation time of future detectors, emitting millions of gravitational wave cycles, with the stellarmass compact object essentially acting as a tracer of the supermassive black hole spacetime [397]. Ringdown gravitational waves are always emitted after black holes merge and the remnant settles down into its final configuration. During the ringdown, the highlydistorted remnant radiates all excess degrees of freedom and this radiation carries a signature of whether the nohair theorems hold in its quasinormal mode spectrum (see, e.g., [68] for a recent review).
Both electromagnetic and gravitational wave tests need a metric with which to model accretion disks, quasiperiodic oscillations, or extreme massratio inspirals. One can classify these metrics as direct or generic, paralleling the discussion in Section 5.2. Direct metrics are exact solutions to a specific set of field equations, with which one can derive observables. Examples of such metrics are the MankoNovikov metric [302] and the slowlyspinning blackhole metric in dynamical ChernSimons gravity [466]. When computing observables with these metrics, one usually assumes that all radiative and dynamical process (e.g., the radiationreaction force) are as predicted in GR. Generic metrics are those that parametrically modify the Kerr spacetime, such that for certain parameter choices one recovers identically the Kerr metric, while for others, one has a deformation of Kerr. Generic metrics can be further classified into two subclasses, Ricciflat versus nonRicciflat, depending on whether they satisfy R_{ μν } = 0.
Let us first consider direct metric tests of the nohair theorem. The most studied direct metric is the MankoNovikov one, which although an exact, stationary and axisymmetric solution to the vacuum Einstein equations, does not represent a black hole, as the event horizon is broken along the equator by a ring singularity [302]. Just like the Kerr metric, the MankoNovikov metric possesses an ergoregion, but unlike the former, it also possesses regions of closed timelike curves that overlap the ergoregion. Nonetheless, an appealing property of this metric is that it deviates continuously from the Kerr metric through certain parameters that characterize the higher multiple moments of the solution.
The first geodesic study of MankoNovikov spacetimes was carried out by Gair et al. [182]. They found that there are two ringlike regions of bound orbits: an outer one where orbits look regular and integrable, as there exist four isolating integrals of the motion; and an inner one where orbits are chaotic and thus ergodic. Gair et al. [182] suggested that orbits that transition from the integrable to the chaotic region would leave a clear observable signature in the frequency spectrum of the emitted gravitational waves. However, they also noted that chaotic regions exist only very close to the central body and are probably not astrophysically accessible. The study of Gair et al. [182] was recently confirmed and followed up by Contopoulos et al. [116]. They studied a wide range of geodesics and found that, in addition to an inner chaotic region and an outer regular region, there are also certain Birkhoff islands of stability. When an extreme massratio inspiral traverses such a region, the ratio of resonant fundamental frequencies would remain constant in time, instead of increasing monotonically. Such a feature would impact the gravitational waves emitted by such a system, and it would signal that the orbit equations are nonintegrable and the central object is not a Kerr black hole.
The study of chaotic motion in geodesics of nonKerr spacetimes is by no means new. Chaos has also been found in geodesics of ZipoyVoorheesWeyl and Curzon spacetimes with multiple singularities [391, 392] and in general for ZipoyVoorhees spacetimes in [296], of perturbed Schwarzschild spacetimes [287], of Schwarzschild spacetimes with a dipolar halo [286, 288, 209] of ErezRosen spacetimes [210], and of deformed generalizations of the TomimatsySato spacetime [154]. One might worry that such chaotic orbits will depend on the particular spacetime considered, but recently Apostolatos et al. [31] and LukesGerakopoulos et al. [297] have argued that the Birkhoff islands of stability are a general feature. Although the Kolmogorov, Arnold, and Moser theorem [270, 35, 321] states that phase orbit tori of an integrable system are only deformed if the Hamiltonian is perturbed, the PoincareBirkhoff theorem [292] states that resonant tori of integrable systems actually disintegrate, leaving behind a chain of Birkhoff islands. These islands are only characterized by the ratio of winding frequencies that equals a rational number, and thus, they constitute a distinct and generic feature of nonintegrable systems [31, 297]. Given an extreme massratio gravitationalwave detection, one can monitor the ratio of fundamental frequencies and search for plateaus in their evolution, which would signal nonintegrability. Of course, whether detectors can resolve such plateaus depends on the initial conditions of the orbits and the physical system under consideration (these determine the thickness of the islands), as well as the mass ratio (this determines the radiationreaction timescale) and the distance and mass of the central black hole (this determines the SNR).
Another example of a direct metric test of the nohair theorem is through the use of the slowlyrotating dynamical ChernSimons black hole metric [466]. Unlike the MankoNovikov metric, the dynamical ChernSimons one does represent a black hole, i.e., it possesses an event horizon, but it evades the nohair theorems because it is not a solution to the Einstein equations. Sopuerta and Yunes [390] carried out the first extreme massratio inspiral analysis when the background supermassive black hole object is taken to be such a ChernSimons black hole. They used a semirelativistic model [368] to evolve extreme massratio inspirals and found that the leadingorder modification comes from a modification to the geodesic trajectories, induced by the nonKerr modifications of the background. Because the latter correspond to a strongfield modification to GR, modifications in the trajectories are most prominent for zoomwhirl orbits, as the small compact object zooms around the supermassive black hole in a region of unstable orbits, close to the event horizon. These modifications were then found to propagate into the gravitational waves emitted, leading to a dephasing that could be observed or ruled out with future gravitationalwave observations to roughly the horizon scale of the supermassive black hole, as has been recently confirmed by Canizares et al. [93]. However, these studies may be underestimates, given that they treat the black hole background in dynamical ChernSimons gravity only to firstorder in spin.
A final example of a direct metric test of the nohair theorems is to consider black holes that are not in vacuum. Barausse et al. [52] studied extrememassratio inspirals in a Kerrblackhole background that is perturbed by a selfgravitating, homogeneous torus that is compact, massive and close to the Kerr black hole. They found that the presence of this torus impacts the gravitational waves emitted during such inspirals, but only weakly, making it difficult to distinguish the presence of matter. Yunes et al. [462] and Kocsis et al. [267] carried out a similar study, where this time they considered a small compact object inspiraling completely within a geometrically thin, radiationpressure dominated accretion disk. They found that diskinduced migration can modify the radiationreaction force sufficiently so as to leave observable signatures in the waveform, provided the accretion disk is sufficiently dense in the radiationdominated regime and a gap opens up. However, these tests of the nohair theorem will be rather difficult as most extrememassratio inspirals are not expected to be in an accretion disk.
That the bumps represent unphysical matter should not be a surprise, since by the nohair theorems, if the bumps are to satisfy the vacuum Einstein equations they must either break stationarity or violate the regularity condition. Naked singularities are an example of the latter. A Lorentzviolating massive field coupled to the Einstein tensor is another example [155]. Gravitational wave tests with bumpy black holes must then be understood as null tests: one assumes the default hypothesis that GR is correct and then sets out to test whether the data rejects or fails to reject this hypothesis (a null hypothesis can never be proven). Unfortunately, however, bumpy black hole metrics cannot parameterize spacetimes in modified gravity theories that lead to corrections in the field equations that are not proportional to the Ricci tensor, such as for example in dynamical ChernSimons or in EinsteinDilatonGaussBonnet modified gravity.
Other bumpy black hole metrics have also been recently proposed. Glampedakis and Babak [196] proposed a different type of stationary and axisymmetric bumpy black hole through the HartleThorne metric [218], with modifications to the quadrupole moment. They then constructed a “kludge” extreme massratio inspiral waveform and estimated how well the quadrupole deformation could be measured [44]. However, this metric is valid only when the supermassive black hole is slowlyrotating, as it derives from the HartleThorne ansatz. Recently, Johansen and Psaltis [252] proposed yet another metric to represent bumpy stationary and sphericallysymmetric spacetimes. This metric introduces one new degree of freedom, which is a function of radius only and assumed to be a series in M/r. Johansen and Psaltis then rotated this metric via the NewmanJanis method [327, 151] to obtain a new bumpy metric for axiallysymmetric spacetimes. However, such a metric possesses a naked ring singularity on the equator, and naked singularities on the poles. As before, none of these bumpy metrics can be mapped to known modified gravity black hole solutions, in the Glampedakis and Babak case [196] because the Einstein equations are assumed to hold to leading order in the spin, while in the Johansen and Psaltis case [252] because a single degree of freedom is not sufficient to model the three degrees of freedom contained in stationary and axisymmetric spacetimes [401, 423].
The only generic nonRicciflat bumpy blackhole metric so far is that of Vigeland, Yunes and Stein [423]. They allowed generic deformations in the metric tensor, only requiring that the new metric perturbatively retained the Killing symmetries of the Kerr spacetime: the existence of two Killing vectors associated with stationarity and axisymmetry, as well as the perturbative existence of a Killing tensor (and thus a Carterlike constant), at least to leading order in the metric deformation. Such requirements imply that the geodesic equations in this new background are fully integrable, at least perturbatively in the metric deformation, which then allows one to solve for the orbital motion of extrememassratio inspirals by adapting previously existing tools. Brink [83, 84, 85, 86, 87] studied the existence of such a secondorder Killing tensor in generic, vacuum, stationary and axisymmetric spacetimes in Einstein’s theory and found that these are difficult to construct exactly. By relaxing this exact requirement, Vigeland, Yunes and Stein [423] found that the existence of a perturbative Killing tensor poses simple differential conditions on the metric perturbation that can be analytically solved. Moreover, they also showed how this new bumpy metric can reproduce all known modified gravity black hole solutions in the appropriate limits, provided these have an at least approximate Killing tensor; thus, these metrics are still vacuum solutions even though R ≠ 1, since they satisfy a set of modified field equations. Although unclear at this junction, it seems that the imposition that the spacetime retains the Kerr Killing symmetries leads to a bumpy metric that is wellbehaved everywhere outside the event horizon (no singularities, no closedtimelike curves, no loss of Lorentz signature). Recently, Gair and Yunes [184] studied how the geodesic equations are modified for a testparticle in a generic orbit in such a spacetime and showed that the bumps are indeed encoded in the orbital motion, and thus, in the gravitational waves emitted during an extrememassratio inspiral.
One might be concerned that such nohair tests of GR cannot constrain modified gravity theories, because Kerr black holes can also be solutions in the latter [360]. This is indeed true provided the modified field equations depend only on the Ricci tensor or scalar. In EinsteinDilatonGaussBonnet or dynamical ChernSimons gravity, the modified field equations depend on the Riemann tensor, and thus, Ricciflat metric need not solve these modified set [473]. Moreover, just because the metric background is identically Kerr does not imply that inspiral gravitational waves will be identical to those predicted in GR. All studies carried out to date, be it direct metric tests or generic metric tests, assume that the only quantity that is modified is the metric tensor, or equivalently, the Hamiltonian or binding energy. Inspiral motion, of course, does not depend just on this quantity, but also on the radiationreaction force that pushes the small object from geodesic to geodesic. Moreover, the gravitational waves generated during such an inspiral depend on the field equations of the theory considered. Therefore, all metric tests discussed above should be considered as partial tests. In general, strongfield modified gravity theories will modify the Hamiltonian, the radiationreaction force and the wave generation.
5.4.3 Ringdown tests of the nohair theorem
Tests of the nohair theorems through the observation of blackhole ringdown date back to Detweiler [146], and it was recently worked out in detail by Dreyer et al. [152]. Let us first imagine that a single complex mode is detected \({\omega _{{\ell _1}{m_1}{n_1}}}\) and one measures separately its real and imaginary parts. Of course, from such a measurement, one cannot extract the measured harmonic triplet (ℓ_{1},m_{1},n_{1}), but instead one only measures the complex frequency \({\omega _{{\ell _1}{m_1}{n_1}}}\). This information is not sufficient to extract the mass and spin angular momentum of the black hole because different quintuplets (M, a, ℓ, m, n) can lead to the same complex frequency \({\omega _{{\ell _1}{m_1}{n_1}}}\). The best way to think of this is graphically: a given observation of \(\omega _{{\ell _1}{m_1}{n_1}}^{(1)}\) traces a line in the complex \({\Omega _{{\ell _1}{m_1}{n_1}}} = M\omega _{{\ell _1}{m_1}{n_1}}^{(1)}\) plane; a given (ℓ, m, n) triplet defines a complex frequency ω_{ ℓmn } that also traces a curve in the complex Ω_{ ℓmn } plane; each intersection of the measured line \({\Omega _{{\ell _1}{m_1}{n_1}}}\) with Ω_{ ℓmn } defines a possible doublet (M, a); since different (ℓ, m, n) triplets lead to different ω_{ ℓmn } curves and thus different intersections, one ends up with a set of doublets S_{1}, out of which only one represents the correct blackhole parameters. We thus conclude that a single mode observation of ringdown gravitational waves is not sufficient to test the nohair theorem [152, 69]
Let us then imagine that one has detected two complex modes, \({\omega _{{\ell _1}{m_1}{n_1}}}\) and \({\omega _{{\ell _2}{m_2}{n_2}}}\). Each detection leads to a separate line \({\Omega _{{\ell _1}{m_1}{n_1}}}\) and \({\Omega _{{\ell _2}{m_2}{n_2}}}\) the complex plane. As before, each (n, ℓ, m) triplet leads to separate curves Ω_{ ℓmn } which will intersect with both \({\Omega _{{\ell _1}{m_1}{n_1}}}\) and \({\Omega _{{\ell _2}{m_2}{n_2}}}\) in the complex plane. Each intersection between Ω_{ ℓmn } and \({\Omega _{{\ell m n}}}\) leads to a set of doublets S_{1}, while each intersection between Ωℓmn and \({\Omega _{{\ell _1}{m_1}{n_1}}}\) leads to another set of doublets S_{2}. However, if the nohair theorems hold sets S_{1} and S_{2} must have at least one element in common. Therefore, a twomode detection allows for tests of the nohair theorem [152, 69]. However, when dealing with a quasicircular blackholebinary inspiral within GR one knows that the dominant mode is ℓ = m = 2. In such a case, the observation of this complex mode by itself allows one to extract the mass and spin angular momentum of the black hole. Then, the detection of the real frequency in an additional mode can be used to test the nohair theorem [69, 65].
Although the logic behind these tests is clear, one must study them carefully to determine whether all systematic and statistical errors are sufficiently under control so that they are feasible. Berti et al. [69, 65] investigated such tests carefully through a frequentist approach. First, they found that a matchedfiltering type analysis with twomode ringdown templates would increase the volume of the template manifold by roughly three orders of magnitude. A better strategy then is perhaps to carry out a Bayesian analysis, like that of Gossan et al. [256, 201]; through such a study one can determine whether a given detection is consistent with a twomode or a onemode hypothesis. Berti et al. [69, 65] also calculated that a SNR of \(\mathcal{O}({10^2})\) would be sufficient to detect the presence of two modes in the ringdown signal and to resolve their frequencies, so that nohair tests would be possible. Strong signals are necessary because one must be able to distinguish at least two modes in the signal. Unfortunately, however, whether the ringdown leads to such strong SNRs and whether the subdominant ringdown modes are of a sufficiently large amplitude depends on a plethora of conditions: the location of the source in the sky, the mass of the final black hole, which depends on the rest mass fraction that is converted into ringdown gravitational waves (the ringdown efficiency), the mass ratio of the progenitor, the magnitude and direction of the spin angular momentum of the final remnant and probably also of the progenitor and the initial conditions that lead to ringdown. Thus, although such tests are possible, one would have to be quite fortunate to detect a signal with the right properties so that a twomode extraction and a test of the nohair theorems is feasible.
5.4.4 The hairy search for exotica
Another way to test GR is to modify the matter sector of the theory through the introduction of matter corrections to the EinsteinHilbert action that violate the assumptions made in the nohair theorems. More precisely, one can study whether gravitational waves emitted by binaries composed of strange stars, like quark stars, or horizonless objects, such as boson stars or gravastars, are different from waves emitted by more traditional neutronstar or blackhole binaries. In what follows, we will describe such hairy tests of the existence of compact exotica.
Boson stars are a classic example of a compact object that is essentially indistinguishable from a black hole in the weak field, but which differs drastically from one in the strong field due to its lack of an event horizon. A boson star is a coherent scalarfield configuration supported against gravitational collapse by its selfinteraction. One can construct several Lagrangian densities that would allow for the existence of such an object, including miniboson stars [178, 179], axiallysymmetric solitons [372], and nonsolitonic stars supported by a noncanonical scalar potential energy [113]. Boson stars are wellmotivated from fundamental theory, since they are the gravitationallycoupled limit of qballs [108, 276], a coherent scalar condensate that can be described classically as a nontopological soliton and that arises unavoidably in viable supersymmetric extensions of the standard model [275]. In all studies carried out to date, boson stars have been studied within GR, but they are also allowed in scalartensor theories [46].
At this junction, one should point out that the choice of a boson star is by no means special; the key point here is to select a strawman to determine whether gravitational waves emitted during the coalescence of compact binaries are sensitive to the presence of an event horizon or the evasion of the nohair theorems induced by a nonvacuum spacetime. Of course, depending on the specific model chosen, it is possible that the exotic object will be unstable to evolution or even to its own rotation. For example, in the case of an extreme massratio inspiral, one could imagine that as the small compact object enters the boson star’s surface, it will accrete the scalar field, forcing the boson star to collapse into a black hole. Alternatively, one can imagine that as two supermassive boson stars merge, the remnant might collapse into a black hole, emitting the usual GR quasinormal modes. What is worse, even when such objects are in isolation, they are unstable under small perturbations if their angular momentum is large, possibly leading to gravitational collapse into a black hole or possibly a scalar explosion [95, 96]. Since most astrophysical black hole candidates are believed to have high spins, such instabilities somewhat limit the interest of horizonless objects. Even then, however, the existence of slowly spinning or non spinning horizonless compact objects cannot be currently ruled out by observation.
Boson stars evade the nohair theorems within GR because they are not vacuum spacetimes, and thus, their metric and quasinormal mode spectrum cannot be described by just their mass and spin angular momentum; one also requires other quantities intrinsic to the scalarfield energy momentum tensor, scalar hair. Therefore, as before, two types of gravitational wave tests for scalar hair have been proposed: extrememassratio inspiral tests and ringdown tests. As for the former, several studies have been carried out that considered a supermassive boson star background. Kesden et al. [263] showed that stable circular orbits exist both outside and inside of the surface of the boson star, provided the small compact object interacts with the background only gravitationally. This is because the effective potential for geodesic motion in such a bosonstar background lacks the Schwarzschildlike singular behavior at small radius, instead turning over and allowing for a new minimum. Gravitational waves emitted in such a system would then stably continue beyond what one would expect if the background had been a supermassive black hole; in the latter case the small compact object would simply disappear into the horizon. Kesden et al. [263] found that orbits inside the boson star exhibit strong precession, exciting high frequency harmonics in the waveform, and thus allowing one to easily distinguish between such boson stars from blackhole backgrounds.
Just as the inspiral phase is modified by the presence of a boson star, the merger phase is also greatly altered, but this must be treated fully numerically. A few studies have found that the merger of boson stars leads to a spinning bar configuration that either fragments or collapses into a Kerr black hole [339, 338]. Of course, the gravitational waves emitted during such a merger will be drastically different from those produced when black holes merge. Unfortunately, the complexity of such simulations makes predictions difficult for any one given example, and the generalization to other more complicated scenarios, such as theories with modified field equations, is currently not feasible.
Recently, Pani et al. [340, 341] revisited this problem, but instead of considering a supermassive boson star, they considered a gravastar. This object consists of a Schwarzschild exterior and a de Sitter interior, separated by an infinitely thin shell with finite tension [307, 100]. Pani et al. [341] calculated the gravitational waves emitted by a stellarmass compact object in a quasicircular orbit around such a gravastar background. In addition to considering a different background, Pani et al. used a radiativeadiabatic waveform generation model to describe the gravitational waves [351, 238, 239, 458, 456, 459], instead of the kludge scheme used by Kesden et al. [49, 44, 456]. Pani el al. [341] concluded that the waves emitted during such inspirals are sufficiently different that they could be used to discern between a Kerr black hole and a gravastar.
On the ringdown side of nohair tests, several studies have been carried out. Berti and Cardoso [66] calculated the quasinormal mode spectrum of boson stars. Chirenti and Rezzolla [105] studied the nonradial, axial perturbations of gravastars, and Pani et al. [340] the nonradial, axial and polar oscillations of gravastars. Medved et al. [309, 310] considered the quasinormal ringdown spectrum of skyrmion black holes [386]. In all cases, it was found that the quasinormal mode spectrum of such objects could be used to discern between them and Kerr black holes. Of course, such tests still require the detection of ringdown gravitational waves with the right properties, such that more than one mode can be discerned and extracted from the signal (see Section 5.4.3).
6 Musings About the Future
Gravitational waves hold the key to testing Einstein’s theory of general relativity (GR) to new exciting levels in the previously unexplored strongfield regime. Depending on the type of wave that is detected, e.g., compact binary inspirals, mergers, ringdowns, continuous sources, supernovae, etc, different tests will be possible. Irrespective of the type of wave detected, two research trends seem currently to be arising: direct tests and generic tests. These trends aim at answering different questions. With direct tests, one wishes to determine whether a certain modified theory is consistent with the data. Generic tests, on the other hand, ask whether the data is statistically consistent with our canonical beliefs. Or put another way: are there any statisticallysignificant deviations present in the data from what we expected to observe? This approach is currently used in cosmological observations, for example by the WMAP team, and it is particularly wellsuited when one tries to remain agnostic as to which is the correct theory of nature. Given that we currently have no data in the strongfield, it might be too restrictive to assume GR is correct prior to verifying that this is the case.
What one would like to believe is that gravitational waves will be detected by the end of this decade, either through groundbased detectors or through pulsar timing arrays. Given this, there is a concrete effort to develop the proper formalism and implementation pipelines to test Einstein’s theory once data becomes available. Currently, the research groups separate into two distinct classes: theory and implementation. The theory part of the research load is being carried out at a variety of institutions without a given focal point. The implementation part is being done mostly within the LIGO Scientific Collaboration and the pulsar timing consortia. Crosscommunication between the theory and implementation groups has recently flourished and one expects more interdisciplinary work in the future.
So many accomplishments have been made in the past 50 years that it is almost impossible to list them here. From the implementation side, perhaps one of the most important is the actual construction and operation of the initial LIGO instruments at design sensitivity in all of their frequency domain. This is a tremendously important engineering and physics challenge. Similarly, the construction of impressive pulsar timing arrays, and the timing of these pulses to nanosecond precision is an instrumental and data analysis feat to be admired. Without these observatories no waves would be detectable in the future, and of course, no tests of Einstein’s theory would be feasible. On the theory side, perhaps the most important accomplishment has been the understanding of the inspiral phase to reallyhigh postNewtonian order and the merger phase with numerical simulations. The latter, in particular, had been an unsolved problem for over 50 years, until very recently. It is these accomplishments that then allow us to postulate modified inspiral template families, since we understand what the GR expectation is. This is particularly true if one is considering small deformations away from Einstein’s theory, as it would be impossible to perturb about an unknown solution.
The main questions that are currently at the forefront are the following. On the theory side of things, one would wish to understand the inspiral to high postNewtonian order in certain strongfield modifications to GR, like dynamical ChernSimons gravity or EinsteinDilatonGaussBonnet theory. One would also like to investigate theories with preferred frames, such as EinsteinAether theory or HořavaLifshitz gravity, which will lead to Lorentz violating observables. Understanding these theories to high postNewtonian order is particularly important for those that predict dipolar gravitational emission, such as EinsteinDilatonGaussBonnet theory. Such corrections dominate over Einstein’s quadrupole emission at sufficiently low velocities.
Of course, a full inspiralmergerringdown template is not complete unless we also understand the merger. This would require full numerical simulations, which are very taxing even within GR. Once one modifies the Einstein field equations, the characteristic structure of the evolution equations will also likely change, and it is unclear whether the standard evolution methods will continue to work. Moreover, when dealing with the merger phase, one is usually forced to treat the modified theory as exact, instead of as an effective theory. Without the latter, it is likely that certain modified theories will not have a wellposed initial value problem, which would force any numerical evolution to fail. Of course, one could orderreduce these equations and then use these to evolve blackhole spacetimes. Much work still remains to be done to understand whether this is feasible.
On the implementation side of things, there is also much work that remains to be done. Currently, efforts are only beginning on the implementation of Bayesian frameworks for hypothesis testing. This seems today like one of the most promising approaches to testing Einstein’s theory with gravitational waves. Current studies concentrate mostly on singledetectors, but by the beginning of the next decade we expect four or five detectors to be online, and thus, one would like to see these implementations extended. The use of multiple detectors also opens the door to the extraction of new information, such as multiple polarization modes, a precise location of the source in the sky, etc. Moreover, the evidence for a given model increases dramatically if the event is observed in several detectors. One therefore expects that the strongest tests of GR will come from leveraging the data from all detectors in a multiplycoincident event.
Ultimately, research is moving toward the construction of robust techniques to test Einstein’s theory. A general push is currently observed toward the testing of general principles that serve as foundations of GR. This allows one to answer general questions, such as: Does the graviton have a mass? Are compact objects represented by the Kerr metric and the nohair theorems satisfied? Does the propagating metric perturbation possess only two transversetraceless polarization modes? What is the rate of change of a binary’s binding energy? Do naked singularities exist in nature and are orbits chaotic? Is Lorentzviolation present in the propagation of gravitons? These are examples of questions that can be answered once gravitational waves are detected. The more questions of this type that are generated and the more robust the methods to answer them are, the more stringent the test of Einstein’s theories and the more information we will obtain about the gravitational interaction in a previously unexplored regime.
Footnotes
 1.
Notice that “strong field” is not synonymous with Planckscale physics in this context. In fact, a stationary black hole would not serve as a probe of the strong field, even if one were to somehow acquire information about the gravitational potential close to the singularity. This is because any such observation would necessarily be lacking information about the dynamical sector of the gravitational interaction. Planckscale physics is perhaps more closely related to strongcurvature physics.
 2.
Stability and wellposedness are not the same concepts and they do not necessarily imply each other. For example, a wellposed theory might have stable and unstable solutions. For illposed theories, it does not make sense to talk about stability of solutions.
 3.
The process of spontaneous scalarization in a particular type of scalartensor theory [129, 130] does introduce strongfield modifications because it induces nonperturbative corrections that can affect the structure of neutron stars. This subclass of scalartensor theories would satisfy Property (4).
 4.
The model considered by [174] is not phenomenological, but it contains a ghost mode.
 5.
Technically, EinsteinDilatonGauss—Bonnet gravity has a very particular set of coupling functions f_{1}(ϑ) = f_{2} (ϑ) = f_{3} (ϑ) ∞ e^{ γϑ }, where γ is a constant. However, in most cases one can expand about γϑ ≪ 1, so that the functions become linear in the scalar field.
 6.
Formally, as α_{ i } → 0, one recovers GR with a dynamical scalar field. However, the latter does not couple to the metric or the matter sector, so it does not lead to any observable effects that distinguish it from GR.
 7.
The tensor \(\mathcal{K}_{\mu \nu}^{(1)}\) is sometimes written as C_{ μν } and referred to as the Ctensor.
 8.
 9.
 10.
The scalar field of Horbatsch and Burgess satisfies \(\square \psi = \mu {g^{\mu \nu}}\Gamma _{\mu \nu}^t\), and thus □ψ = 0 for stationary and axisymmetric spacetimes, since the metric is independent of time an azimuthal coordinate. However, notice that is not necessarily needed for Jacobson’s construction [246] to be possible.
 11.
All LISA bounds refer to the classic LISA configuration.
 12.
Even if it is not linear, the effect should scale with positive powers of λ_{GW}. It is difficult to think of any parityviolating theory that would lead to an inversely proportional relation.
 13.
Notice that these relations are independent of the polytropic constant K, where p = Kρ^{(1+1/n)}, as shown in [452].
Notes
Acknowledgements
We would like to thank Emanuele Berti, Vitor Cardoso, William Nelson, Bangalore Sathyaprakash, and Leo Stein for many discussions. We would also like to thank Laura Sampson and Tyson Littenberg for helping us write parts of the data analysis sections. Finally, we would like to thank Matt Adams, Katerina Chatziioannou, Tyson Littenberg, Laura Sampson, and Kent Yagi for proofreading earlier versions of this manuscript. Nicolás Yunes would like to acknowledge support from NSF grant PHY1114374 and NASA grant NNX11AI49G, under subaward 00001944. Xavier Siemens would like to acknowledge support from the NSF CAREER award number 0955929, the PIRE award number 0968126, and award number 0970074.
References
 [1]Abadie, J. et al. (LIGO Scientific Collaboration), “Calibration of the LIGO Gravitational Wave Detectors in the Fifth Science Run”, Nucl. Instrum. Methods A, 624, 223–240, (2010). [DOI], [arXiv: 1007.3973 [grqc]]. (Cited on page 42.)ADSCrossRefGoogle Scholar
 [2]Abbott, B. et al. (LIGO Scientific Collaboration), “LIGO: The Laser Interferometer GravitationalWave Observatory”, Rep. Prog. Phys., 72, 076901, (2009). [DOI], [arXiv:0711.3041 [grqc]]. (Cited on pages 8 and 72.)ADSCrossRefGoogle Scholar
 [3]Abramovici, A. et al., “LIGO: The Laser Interferometer GravitationalWave Observatory”, Science, 256, 325–333, (1992). [DOI], [ADS]. (Cited on page 8.)ADSCrossRefGoogle Scholar
 [4]Accadia, T. et al. (Virgo Collaboration), “Calibration and sensitivity of the Virgo detector during its second science run”, Class. Quantum Grav., 28, 025005, (2011). [DOI], [arXiv:1009.5190 [grqc]]. Erratum: 10.1088/02649381/28/7/079501. (Cited on page 42.)ADSCrossRefGoogle Scholar
 [5]Acernese, F. et al. (VIRGO Collaboration), “The Virgo Detector”, in Tricomi, A., Albergo, S. and Chiorboli, M., eds., IFAE 2005: XVII Incontri de Fisica delle Alte Energie; 17th Italian Meeting on High Energy, Catania, Italy, 30 March–2 April 2005, AIP Conference Proceedings, 794, pp. 307–310, (American Institute of Physics, Melville, NY, 2005). [DOI]. (Cited on page 8.)Google Scholar
 [6]Acernese, F. et al. (Virgo Collaboration), “Status of Virgo detector”, Class. Quantum Grav., 24, S381–S388 (2007). [DOI], [ADS]. (Cited on pages 8 and 72.)ADSzbMATHCrossRefGoogle Scholar
 [7]Adelberger, E.G., Heckel, B.R., Hoedl, S.A., Hoyle, C.D., Kapner, D.J. and Upadhye, A., “ParticlePhysics Implications of a Recent Test of the Gravitational InverseSquare Law”, Phys. Rev. Lett., 98, 131104, (2007). [DOI], [arXiv:hepph/0611223]. (Cited on pages 25 and 69.)ADSCrossRefGoogle Scholar
 [8]Adler, S.L., “AxialVector Vertex in Spinor Electrodynamics”, Phys. Rev., 177, 2426–2438, (1969). [DOI], [ADS]. (Cited on page 28.)ADSCrossRefGoogle Scholar
 [9]Aharony, O., Gubser, S.S., Maldacena, J.M., Ooguri, H. and Oz, Y., “Large N field theories, string theory and gravity”, Phys. Rep., 323, 183–386, (2000). [DOI], [arXiv:hepth/9905111]. (Cited on page 25.)ADSMathSciNetCrossRefGoogle Scholar
 [10]Akmal, A., Pandharipande, V.R. and Ravenhall, D.G., “The equation of state of nucleon matter and neutron star structure”, Phys. Rev. C, 58, 1804–1828, (1998). [DOI], [ADS], [arXiv:nuclth/9804027]. (Cited on page 81.)ADSCrossRefGoogle Scholar
 [11]Alexander, S., Finn, L.S. and Yunes, N., “Gravitationalwave probe of effective quantum gravity”, Phys. Rev. D, 78, 066005, (2008). [DOI], [ADS], [arXiv:0712.2542 [grqc]]. (Cited on pages 29, 69, 71, and 72.)ADSMathSciNetCrossRefGoogle Scholar
 [12]Alexander, S. and Gates Jr, S. J., “Can the string scale be related to the cosmic baryon asymmetry?”, J. Cosmol. Astropart. Phys., 2006(06), 018, (2006). [DOI], [ADS], [arXiv:hepth/0409014]. (Cited on pages 20 and 28.)ADSCrossRefGoogle Scholar
 [13]Alexander, S. and Martin, J., “Birefringent gravitational waves and the consistency check of inflation”, Phys. Rev. D, 71, 063526, (2005). [DOI], [arXiv:hepth/0410230]. (Cited on page 29.)ADSCrossRefGoogle Scholar
 [14]Alexander, S. and Yunes, N., “New PostNewtonian Parameter to Test ChernSimons Gravity”, Phys. Rev. Lett., 99, 241101, (2007). [DOI], [ADS], [arXiv:hepth/0703265]. (Cited on page 30.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [15]Alexander, S. and Yunes, N., “Parametrized postNewtonian expansion of ChernSimons gravity”, Phys. Rev. D, 75, 124022, (2007). [DOI], [ADS], [arXiv:0704.0299 [hepth]]. (Cited on page 30.)ADSMathSciNetCrossRefGoogle Scholar
 [16]Alexander, S. and Yunes, N., “ChernSimons modified gravity as a torsion theory and its interaction with fermions”, Phys. Rev. D, 77, 124040, (2008). [DOI], [ADS], [arXiv:0804.1797 [grqc]]. (Cited on page 20.)ADSMathSciNetCrossRefGoogle Scholar
 [17]Alexander, S. and Yunes, N., “ChernSimons modified general relativity”, Phys. Rep., 480, 1–55, (2009). [DOI], [ADS], [arXiv:0907.2562 [hepth]]. (Cited on pages 20, 29, 30, and 69.)ADSMathSciNetCrossRefGoogle Scholar
 [18]AliHaïmoud, Y., “Revisiting the doublebinarypulsar probe of nondynamical ChernSimons gravity”, Phys. Rev. D, 83, 124050, (2011). [DOI], [ADS], [arXiv:1105.0009 [astroph.HE]]. (Cited on page 30.)ADSCrossRefGoogle Scholar
 [19]AliHaïmoud, Y. and Chen, Y., “Slowlyrotating stars and black holes in dynamical ChernSimons gravity”, Phys. Rev. D, 84, 124033, (2011). [DOI], [ADS], [arXiv: 1110.5329 [astroph.HE]]. (Cited on pages 22, 23, 58, and 60.)ADSCrossRefGoogle Scholar
 [20]Alsing, J., Berti, E., Will, C.M. and Zaglauer, H., “Gravitational radiation from compact binary systems in the massive BransDicke theory of gravity”, Phys. Rev. D, 85, 064041, (2012). [DOI], [arXiv:1112.4903 [grqc]]. (Cited on pages 16, 57, and 78.)ADSCrossRefGoogle Scholar
 [21]AlvarezGaumé, L. and Witten, E., “Gravitational anomalies”, Nucl. Phys. B, 234, 269–330, (1984). [DOI], [ADS]. (Cited on page 28.)ADSMathSciNetCrossRefGoogle Scholar
 [22]Alves, M.E.S. and Tinto, M., “Pulsar Timing Sensitivities to Gravitational Waves from Relativistic Metric Theories of Gravity”, Phys. Rev. D, 83, 123529, (2011). [DOI], [ADS], [arXiv:1102.4824 [grqc]]. (Cited on pages 37, 45, and 49.)ADSCrossRefGoogle Scholar
 [23]AmaroSeoane, P., Gair, J.R., Freitag, M., Miller, M.C., Mandel, I., Cutler, C.J. and Babak, S., “Intermediate and extreme massratio inspirals — astrophysics, science applications and detection using LISA”, Class. Quantum Grav., 24, R113–R169 (2007). [DOI], [ADS], [arXiv:astroph/0703495]. (Cited on pages 8 and 85.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [24]AmaroSeoane, P. et al., “Lowfrequency gravitationalwave science with eLISA/NGO”, Class. Quantum Grav., 29, 124016, (2012). [DOI], [ADS], [arXiv:1202.0839 [grqc]]. (Cited on page 84.)ADSCrossRefGoogle Scholar
 [25]AmaroSeoane, P. et al., “eLISA: Astrophysics and cosmology in the millihertz regime”, GW Notes, 6, 4–110, (2013). [ADS], [arXiv:1201.3621 [astroph.CO]]. URL (accessed 10 October 2013): http://brownbag.lisascience.org/lisagwnotes/. (Cited on page 84.)Google Scholar
 [26]AmelinoCamelia, G., “Testable scenario for relativity with minimum length”, Phys. Lett. B, 510, 255–263, (2001). [DOI], [ADS], [arXiv:hepth/0012238]. (Cited on page 19.)ADSzbMATHCrossRefGoogle Scholar
 [27]AmelinoCamelia, G., “Doubly special relativity”, Nature, 418, 34–35, (2002). [DOI], [arXiv:grqc/0207049]. (Cited on page 19.)ADSzbMATHCrossRefGoogle Scholar
 [28]AmelinoCamelia, G., “DoublySpecial Relativity: Facts, Myths and Some Key Open Issues”, Symmetry, 2, 230–271, (2010). [DOI], [arXiv: 1003.3942 [grqc]]. (Cited on page 19.)MathSciNetzbMATHCrossRefGoogle Scholar
 [29]Amendola, L., Charmousis, C. and Davis, S.C., “Solar System Constraints on GaussBonnet Mediated Dark Energy”, J. Cosmol. Astropart. Phys., 2007(10), 004, (2007). [DOI], [arXiv:0704.0175 [astroph]]. (Cited on page 22.)CrossRefGoogle Scholar
 [30]Anholm, M., Ballmer, S., Creighton, J.D.E., Price, L.R. and Siemens, X., “Optimal strategies for gravitational wave stochastic background searches in pulsar timing data”, Phys. Rev. D, 79, 084030, (2009). [DOI], [arXiv:0809.0701 [grqc]]. (Cited on page 48.)ADSCrossRefGoogle Scholar
 [31]Apostolatos, T.A., LukesGerakopoulos, G. and Contopoulos, G., “How to Observe a NonKerr Spacetime Using Gravitational Waves”, Phys. Rev. Lett., 103, 111101, (2009). [DOI], [ADS], [arXiv:0906.0093 [grqc]]. (Cited on pages 8, 85, and 86.)ADSMathSciNetCrossRefGoogle Scholar
 [32]ArkaniHamed, N., Dimopoulos, S. and Dvali, G., “The hierarchy problem and new dimensions at a millimeter”, Phys. Lett. B, 429, 263–272, (1998). [DOI], [arXiv:hepph/9803315]. (Cited on page 25.)ADSCrossRefGoogle Scholar
 [33]ArkaniHamed, N., Dimopoulos, S. and Dvali, G., “Phenomenology, astrophysics, and cosmology of theories with submillimeter dimensions and TTeV scale quantum gravity”, Phys. Rev. D, 59, 086004, (1999). [DOI], [ADS], [arXiv:hepph/9807344]. (Cited on page 25.)ADSCrossRefGoogle Scholar
 [34]ArkaniHamed, N., Georgi, H. and Schwartz, M.D., “Effective field theory for massive gravitons and gravity in theory space”, Ann. Phys. (N.Y.), 305, 96–118, (2003). [DOI], [ADS], [arXiv:hepth/0210184]. (Cited on page 18.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [35]Arnold, V.I., “Proof of a theorem of A. N. Kolmogorov on the invariance of quasiperiodic motions under small perturbations of the Hamiltonian”, Russ. Math. Surv., 18(5), 9–36 (1963). [DOI]. (Cited on page 85.)CrossRefGoogle Scholar
 [36]Arun, K.G., “Generic bounds on dipolar gravitational radiation from inspiralling compact binaries”, Class. Quantum Grav., 29, 075011, (2012). [DOI], [ADS], [arXiv:1202.5911 [grqc]]. (Cited on pages 78 and 79.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [37]Arun, K.G., Iyer, B.R., Qusailah, M.S.S. and Sathyaprakash, B.S., “Testing postNewtonian theory with gravitational wave observations”, Class. Quantum Grav., 23, L37–L43 (2006). [DOI], [ADS], [arXiv:grqc/0604018]. (Cited on page 74.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [38]Arun, K.G. and Pai, A., “Tests of General Relativity and Alternative theories of gravity using Gravitational Wave observations”, Int. J. Mod. Phys. D, 22, 1341012, (2013). [DOI], [ADS], [arXiv:1302.2198 [grqc]]. (Cited on pages 55 and 56.)ADSCrossRefGoogle Scholar
 [39]Arun, K.G. and Will, C.M., “Bounding the mass of the graviton with gravitational waves: effect of higher harmonics in gravitational waveform templates”, Class. Quantum Grav., 26, 155002, (2009). [DOI], [ADS], [arXiv:0904.1190 [grqc]]. (Cited on pages 64 and 65.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [40]Arvanitaki, A. and Dubovsky, S., “Exploring the string axiverse with precision black hole physics”, Phys. Rev. D, 83, 044026, (2011). [DOI], [ADS], [arXiv: 1004.3558 [hepth]]. (Cited on page 22.)ADSCrossRefGoogle Scholar
 [41]Ashtekar, A., Balachandran, A.P. and Jo, S., “The CP Problem in Quantum Gravity”, Int. J. Mod. Phys. A, 4, 1493–1514, (1989). [DOI], [ADS]. (Cited on pages 20 and 28.)ADSMathSciNetCrossRefGoogle Scholar
 [42]Ashtekar, A., Bojowald, M. and Lewandowski, J., “Mathematical structure of loop quantum cosmology”, Adv. Theor. Math. Phys., 7, 233–268, (2003). [arXiv:grqc/0304074]. (Cited on page 17.)MathSciNetCrossRefGoogle Scholar
 [43]Ashtekar, A. and Lewandowski, J., “Background independent quantum gravity: a status report”, Class. Quantum Grav., 21, R53–R152 (2004). [DOI], [arXiv:grqc/0404018]. (Cited on page 20.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [44]Babak, S., Fang, H., Gair, J.R., Glampedakis, K. and Hughes, S.A., “‘Kludge’ gravitational waveforms for a testbody orbiting a Kerr black hole”, Phys. Rev. D, 75, 024005 (2007). [DOI], [arXiv:grqc/0607007]. Erratum: 10.1103/PhysRevD.77.049902. (Cited on pages 84, 87, and 91.)ADSMathSciNetCrossRefGoogle Scholar
 [45]Babichev, E. and Deffayet, C., “An introduction to the Vainshtein mechanism”, Class. Quantum Grav., 30, 184001, (2013). [DOI], [ADS], [arXiv:1304.7240 [grqc]]. (Cited on pages 12 and 18.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [46]Balakrishna, J. and Shinkai, H., “Dynamical evolution of boson stars in BransDicke theory”, Phys. Rev. D, 58, 044016, (1998). [DOI], [arXiv:grqc/9712065]. (Cited on page 90.)ADSCrossRefGoogle Scholar
 [47]Bambi, C., Giannotti, M. and Villante, F.L., “Response of primordial abundances to a general modification of GN and/or of the early universe expansion rate”, Phys. Rev. D, 71, 123524, (2005). [DOI], [arXiv:astroph/0503502]. (Cited on page 25.)ADSCrossRefGoogle Scholar
 [48]Bañados, M. and Ferreira, P.G., “Eddington’s Theory of Gravity and Its Progeny”, Phys. Rev. Lett., 105, 011101, (2010). [DOI], [ADS], [arXiv:1006.1769 [astroph.CO]]. (Cited on page 30.)ADSMathSciNetCrossRefGoogle Scholar
 [49]Barack, L. and Cutler, C., “LISA capture sources: Approximate waveforms, signaltonoise ratios, and parameter estimation accuracy”, Phys. Rev. D, 69, 082005, (2004). [DOI], [ADS], [arXiv:grqc/0310125]. (Cited on page 91.)ADSCrossRefGoogle Scholar
 [50]Barack, L. and Cutler, C., “Using LISA extrememassratio inspiral sources to test offKerr deviations in the geometry of massive black holes”, Phys. Rev. D, 75, 042003, (2007). [DOI], [ADS], [arXiv:grqc/0612029]. (Cited on pages 8 and 84.)ADSCrossRefGoogle Scholar
 [51]Barausse, E., Palenzuela, C., Ponce, M. and Lehner, L., “Neutronstar mergers in scalartensor theories of gravity”, Phys. Rev. D, 87, 081506, (2013). [DOI], [ADS], [arXiv:1212.5053 [grqc]]. (Cited on pages 17, 53, 56, and 57.)ADSCrossRefGoogle Scholar
 [52]Barausse, E., Rezzolla, L., Petroff, D. and Ansorg, M., “Gravitational waves from extreme mass ratio inspirals in nonpure Kerr spacetimes”, Phys. Rev. D, 75, 064026, (2007). [DOI], [ADS], [arXiv:grqc/0612123]. (Cited on page 86.)ADSMathSciNetCrossRefGoogle Scholar
 [53]Baskaran, D., Polnarev, A.G., Pshirkov, M.S. and Postnov, K.A., “Limits on the speed of gravitational waves from pulsar timing”, Phys. Rev. D, 78, 044018, (2008). [DOI], [arXiv:0805.3103 [astroph]]. (Cited on page 64.)ADSCrossRefGoogle Scholar
 [54]Bekenstein, J.D., “Relativistic gravitation theory for the MOND paradigm”, Phys. Rev. D, 70, 083509, (2004). [DOI], [arXiv:astroph/0403694]. (Cited on page 17.)ADSCrossRefGoogle Scholar
 [55]Bell, J.S. and Jackiw, R., “A PCAC Puzzle: π^{0} → γγ in the σModel”, Nuovo Cimento A, 60, 47–61, (1969). [DOI]. (Cited on page 28.)ADSCrossRefGoogle Scholar
 [56]Bender, C.M. and Orszag, S.A., Advanced Mathematical Methods for Scientists and Engineers I: Asymptotic Methods and Perturbation Theory, International Series in Pure and Applied Mathematics, (McGrawHill, New York, 1978). (Cited on pages 54 and 68.)zbMATHGoogle Scholar
 [57]Bennett, C.L. et al. (WMAP Collaboration), “Sevenyear Wilkinson Microwave Anisotropy Probe (WMAP) Observations: Are There Cosmic Microwave Background Anomalies?”, Astrophys. J. Suppl. Ser., 192, 17, (2011). [DOI], [ADS], [arXiv:1001.4758 [astroph.CO]]. (Cited on page 51.)ADSCrossRefGoogle Scholar
 [58]Berezhiani, Z., Comelli, D., Nesti, F. and Pilo, L., “Spontaneous Lorentz Breaking and Massive Gravity”, Phys. Rev. Lett., 99, 131101, (2007). [DOI], [arXiv:hepth/0703264]. (Cited on page 19.)ADSCrossRefGoogle Scholar
 [59]Berezhiani, Z., Comelli, D., Nesti, F. and Pilo, L., “Exact Spherically Symmetric Solutions in Massive Gravity”, J. High Energy Phys., 0807, 130, (2008). [DOI], [arXiv:0803.1687 [hepth]]. (Cited on page 19.)ADSMathSciNetCrossRefGoogle Scholar
 [60]Bergshoeff, E.A., Hohm, O. and Townsend, P.K., “New massive gravity”, in Damour, T., Jantzen, R. and Ruffini, R., eds., On Recent Developments in Theoretical and Experimental General Relativity, Astrophysics and Relativistic Field Theories, Proceedings of the MG12 Meeting on General Relativity, Paris, France, 12–18 July 2009, pp. 2329–2331, (World Scientific, Singapore; Hackensack, NJ, 2009). [DOI]. (Cited on pages 18 and 30.)Google Scholar
 [61]Bergshoeff, E.A., Kovacevic, M., Rosseel, J. and Yin, Y., “Massive Gravity: A Primer”, in Calcagni, G., Papantonopoulos, L., Siopsis, G. and Tsamis, N., eds., Quantum Gravity and Quantum Cosmology, Lecture Notes in Physics, 863, pp. 119–145, (Springer, Berlin; New York, 2013). [DOI], [ADS]. (Cited on page 18.)CrossRefGoogle Scholar
 [62]Berry, C.P.L. and Gair, J.R., “Linearized f(R) gravity: Gravitational radiation and solar system tests”, Phys. Rev. D, 83, 104022, (2011). [DOI], [ADS], [arXiv:1104.0819 [grqc]]. (Cited on page 21.)ADSCrossRefGoogle Scholar
 [63]Berti, E., Buonanno, A. and Will, C.M., “Estimating spinning binary parameters and testing alternative theories of gravity with LISA”, Phys. Rev. D, 71, 084025, (2005). [DOI], [ADS], [arXiv:grqc/0411129]. (Cited on pages 55, 56, 64, and 65.)ADSCrossRefGoogle Scholar
 [64]Berti, E., Buonanno, A. and Will, C.M., “Testing general relativity and probing the merger history of massive black holes with LISA”, Class. Quantum Grav., 22, S943–S954 (2005). [DOI], [ADS], [arXiv:grqc/0504017]. (Cited on page 55.)ADSzbMATHCrossRefGoogle Scholar
 [65]Berti, E., Cardoso, J., Cardoso, V. and Cavaglià, M., “Matched filtering and parameter estimation of ringdown waveforms”, Phys. Rev. D, 76, 104044, (2007). [DOI], [ADS], [arXiv:0707.1202 [grqc]]. (Cited on pages 42 and 89.)ADSCrossRefGoogle Scholar
 [66]Berti, E. and Cardoso, V., “Supermassive black holes or boson stars? Hair counting with gravitational wave detectors”, Int. J. Mod. Phys. D, 15, 2209–2216, (2006). [DOI], [ADS], [arXiv:grqc/0605101]. (Cited on page 91.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [67]Berti, E., Cardoso, V., Gualtieri, L., Horbatsch, M.W. and Sperhake, U., “Numerical simulations of single and binary black holes in scalartensor theories: Circumventing the nohair theorem”, Phys. Rev. D, 87, 124020, (2013). [DOI], [ADS], [arXiv:1304.2836 [grqc]]. (Cited on pages 16, 25, and 53.)ADSCrossRefGoogle Scholar
 [68]Berti, E., Cardoso, V. and Starinets, A.O., “Quasinormal modes of black holes and black branes”, Class. Quantum Grav., 26, 163001, (2009). [DOI], [ADS], [arXiv:0905.2975 [grqc]]. (Cited on pages 85 and 88.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [69]Berti, E., Cardoso, V. and Will, C.M., “Gravitationalwave spectroscopy of massive black holes with the space interferometer LISA”, Phys. Rev. D, 73, 064030, (2006). [DOI], [ADS], [arXiv:grqc/0512160]. (Cited on pages 88 and 89.)ADSMathSciNetCrossRefGoogle Scholar
 [70]Berti, E., Gair, J.R. and Sesana, A., “Graviton mass bounds from spacebased gravitationalwave observations of massive black hole populations”, Phys. Rev. D, 84, 101501, (2011). [DOI], [ADS], [arXiv:1107.3528 [grqc]]. (Cited on pages 64 and 65.)ADSCrossRefGoogle Scholar
 [71]Berti, E., Gualtieri, L., Horbatsch, M.W. and Alsing, J., “Light scalar field constraints from gravitationalwave observations of compact binaries”, Phys. Rev. D, 85, 122005, (2012). [DOI], [ADS], [arXiv:1204.4340 [grqc]]. (Cited on page 57.)ADSCrossRefGoogle Scholar
 [72]Berti, E., Iyer, S. and Will, C.M., “PostNewtonian diagnosis of quasiequilibrium configurations of neutron starneutron star and neutron starblack hole binaries”, Phys. Rev. D, 77, 024019, (2008). [DOI], [arXiv:0709.2589 [grqc]]. (Cited on page 81.)ADSCrossRefGoogle Scholar
 [73]Bertotti, B., Iess, L. and Tortora, P., “A test of general relativity using radio links with the Cassini spacecraft”, Nature, 425, 374–376, (2003). [DOI], [ADS]. (Cited on pages 16, 22, 55, and 56.)ADSCrossRefGoogle Scholar
 [74]Bildsten, L. and Cutler, C., “Tidal interactions of inspiraling compact binaries”, Astrophys. J., 400, 175–180, (1992). [DOI], [ADS]. (Cited on page 83.)ADSCrossRefGoogle Scholar
 [75]Blanchet, L., “Gravitational Radiation from PostNewtonian Sources and Inspiralling Compact Binaries”, Living Rev. Relativity, 9, lrr20064 (2006). [DOI], [ADS]. URL (accessed 15 April 2013): http://www.livingreviews.org/lrr20064. (Cited on pages 52 and 78.)
 [76]Blas, D. and Sanctuary, H., “Gravitational radiation in Hořava gravity”, Phys. Rev. D, 84, 064004, (2011). [DOI], [arXiv:1105.5149 [grqc]]. (Cited on page 19.)ADSCrossRefGoogle Scholar
 [77]Bojowald, M., “Loop Quantum Cosmology”, Living Rev. Relativity, 8, lrr200511 (2005). [DOI], [arXiv:grqc/0601085]. URL (accessed 15 April 2013): http://www.livingreviews.org/lrr200511. (Cited on page 17.)
 [78]Bojowald, M. and Hossain, G.M., “Loop quantum gravity corrections to gravitational wave dispersion”, Phys. Rev. D, 77, 023508, (2008). [DOI], [ADS], [arXiv:0709.2365 [grqc]]. (Cited on pages 17 and 19.)ADSMathSciNetCrossRefGoogle Scholar
 [79]Boulware, D.G. and Deser, S., “StringGenerated Gravity Models”, Phys. Rev. Lett., 55, 2656, (1985). [DOI]. (Cited on page 22.)ADSCrossRefGoogle Scholar
 [80]Boyle, L., “The general theory of porcupines, perfect and imperfect”, arXiv, eprint, (2010). [ADS], [arXiv:1008.4997 [grqc]]. (Cited on page 44.)Google Scholar
 [81]Boyle, L., “Perfect porcupines: ideal networks for low frequency gravitational wave astronomy”, arXiv, eprint, (2010). [ADS], [arXiv:1003.4946 [grqc]]. (Cited on page 44.)Google Scholar
 [82]Brans, C. and Dicke, R.H., “Mach’s Principle and a Relativistic Theory of Gravitation”, Phys. Rev., 124, 925–935, (1961). [DOI], [ADS]. (Cited on pages 14 and 15.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [83]Brink, J., “Spacetime encodings. I. A spacetime reconstruction problem”, Phys. Rev. D, 78, 102001, (2008). [DOI], [arXiv:0807.1178 [grqc]]. (Cited on page 88.)ADSMathSciNetCrossRefGoogle Scholar
 [84]Brink, J., “Spacetime encodings. II. Pictures of integrability”, Phys. Rev. D, 78, 102002, (2008). [DOI], [ADS], [arXiv:0807.1179 [grqc]]. (Cited on page 88.)ADSMathSciNetCrossRefGoogle Scholar
 [85]Brink, J., “Spacetime encodings. III. Second order Killing tensors”, Phys. Rev. D, 81, 022001, (2010). [DOI], [arXiv:0911.1589 [grqc]]. (Cited on page 88.)ADSMathSciNetCrossRefGoogle Scholar
 [86]Brink, J., “Spacetime encodings. IV. The relationship between Weyl curvature and killing tensors in stationary axisymmetric vacuum spacetimes”, Phys. Rev. D, 81, 022002, (2010). [DOI], [arXiv:0911.1595 [grqc]]. (Cited on page 88.)ADSMathSciNetCrossRefGoogle Scholar
 [87]Brink, J., “Formal solution of the fourth order Killing equations for stationary axisymmetric vacuum spacetimes”, Phys. Rev. D, 84, 104015, (2011). [DOI], [arXiv:0911.4161 [grqc]]. (Cited on page 88.)ADSCrossRefGoogle Scholar
 [88]Brito, R., Cardoso, V. and Pani, P., “Massive spin2 fields on black hole spacetimes: Instability of the Schwarzschild and Kerr solutions and bounds on graviton mass”, Phys. Rev. D, 88, 023514, (2013). [DOI], [ADS], [arXiv: 1304.6725 [grqc]]. (Cited on pages 19, 64, and 65.)ADSCrossRefGoogle Scholar
 [89]Burgess, C.P., “Quantum Gravity in Everyday Life: General Relativity as an Effective Field Theory”, Living Rev. Relativity, 7, lrr20045 (2004). [DOI], [arXiv:grqc/0311082]. URL (accessed 15 April 2013): http://www.livingreviews.org/lrr20045. (Cited on pages 20 and 22.)
 [90]Calcagni, G. and Mercuri, S., “The BarberoImmirzi field in canonical formalism of pure gravity”, Phys. Rev. D, 79, 084004, (2009). [DOI], [arXiv:0902.0957 [grqc]]. (Cited on page 28.)ADSMathSciNetCrossRefGoogle Scholar
 [91]Campanelli, M. and Lousto, C.O., “Are black holes in BransDicke theory precisely the same as a general relativity?”, Int. J. Mod. Phys. D, 2, 451–462, (1993). [DOI], [arXiv:grqc/9301013]. (Cited on page 52.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [92]Campbell, B.A., Kaloper, N. and Olive, K.A., “Classical hair for KerrNewman black holes in string gravity”, Phys. Lett. B, 285, 199–205, (1992). [DOI], [ADS]. (Cited on page 22.)ADSMathSciNetCrossRefGoogle Scholar
 [93]Canizares, P., Gair, J.R. and Sopuerta, C.F., “Testing ChernSimons modified gravity with observations of extrememassratio binaries”, J. Phys.: Conf. Ser., 363, 012019, (2012). [DOI], [ADS], [arXiv:1206.0322 [grqc]]. (Cited on pages 8 and 86.)ADSGoogle Scholar
 [94]Cardoso, V., Chakrabarti, S., Pani, P., Berti, E. and Gualtieri, L., “Floating and sinking: The Imprint of massive scalars around rotating black holes”, Phys. Rev. Lett., 107, 241101, (2011). [DOI], [ADS], [arXiv:1109.6021 [grqc]]. (Cited on pages 56 and 78.)ADSCrossRefGoogle Scholar
 [95]Cardoso, V., Pani, P., Cadoni, M. and Cavaglià, M., “Ergoregion instability of ultracompact astrophysical objects”, Phys. Rev. D, 77, 124044, (2008). [DOI], [arXiv:0709.0532 [grqc]]. (Cited on page 90.)ADSCrossRefGoogle Scholar
 [96]Cardoso, V., Pani, P., Cadoni, M. and Cavaglià, M., “Instability of hypercompact Kerrlike objects”, Class. Quantum Grav., 25, 195010, (2008). [DOI], [arXiv:0808.1615 [grqc]]. (Cited on page 90.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [97]Carson, J.E., “GLAST: Physics goals and instrument status”, J. Phys.: Conf. Ser., 60, 115–118, (2007). [DOI], [arXiv:astroph/0610960]. (Cited on page 72.)ADSGoogle Scholar
 [98]Carter, B., “Axisymmetric Black Hole Has Only Two Degrees of Freedom”, Phys. Rev. Lett., 26, 331–333, (1971). [DOI], [ADS]. (Cited on pages 16 and 83.)ADSCrossRefGoogle Scholar
 [99]Chamberlin, S.J. and Siemens, X., “Stochastic backgrounds in alternative theories of gravity: overlap reduction functions for pulsar timing arrays”, Phys. Rev. D, 85, 082001, (2012). [DOI], [ADS], [arXiv:1111.5661 [astroph.HE]]. (Cited on pages 37, 46, 48, and 49.)ADSCrossRefGoogle Scholar
 [100]Chapline, G., “Quantum Phase Transitions and the Failure of Classical General Relativity”, Int. J. Mod. Phys. A, 18, 3587–3590, (2003). [DOI], [ADS], [arXiv:grqc/0012094]. (Cited on page 91.)ADSCrossRefGoogle Scholar
 [101]Chatterji, S., Lazzarini, A., Stein, L., Sutton, P.J., Searle, A. and Tinto, M., “Coherent network analysis technique for discriminating gravitationalwave bursts from instrumental noise”, Phys. Rev. D, 74, 082005, (2006). [DOI], [arXiv:grqc/0605002]. (Cited on pages 43, 45, and 81.)ADSCrossRefGoogle Scholar
 [102]Chatziioannou, K., Yunes, N. and Cornish, N.J., “Modelindependent test of general relativity: An extended postEinsteinian framework with complete polarization content”, Phys. Rev. D, 86, 022004, (2012). [DOI], [ADS], [arXiv: 1204.2585 [grqc]]. (Cited on pages 43, 44, 45, 53, 54, 76, 78, 79, and 81.)ADSCrossRefGoogle Scholar
 [103]Chernoff, D.F. and Finn, L.S., “Gravitational radiation, inspiraling binaries, and cosmology”, Astrophys. J., 411, L5–L8 (1993). [DOI], [arXiv:grqc/9304020]. (Cited on page 39.)ADSCrossRefGoogle Scholar
 [104]Chiba, T., “1/R gravity and scalartensor gravity”, Phys. Lett. B, 575, 1–3, (2003). [DOI], [ADS], [arXiv:astroph/0307338]. (Cited on page 16.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [105]Chirenti, C.B.M.H. and Rezzolla, L., “How to tell a gravastar from a black hole”, Class. Quantum Grav., 24, 4191–4206, (2007). [DOI], [arXiv:0706.1513 [grqc]]. (Cited on page 91.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [106]Choudhury, S.R., Joshi, G.C., Mahajan, S. and McKellar, B.H.J., “Probing large distance higher dimensional gravity from lensing data”, Astropart. Phys., 21, 559–563, (2004). [DOI], [arXiv:hepph/0204161]. (Cited on page 18.)ADSCrossRefGoogle Scholar
 [107]Chouha, P.R. and Brandenberger, R.H., “TDuality and the Spectrum of Gravitational Waves”, arXiv, eprint, (2005). [ADS], [arXiv:hepth/0508119]. (Cited on page 19.)Google Scholar
 [108]Coleman, S.R., “Qballs”, Nucl. Phys. B, 262, 263–283, (1985). [DOI], [ADS]. (Cited on page 90.)ADSCrossRefGoogle Scholar
 [109]Colladay, D. and Kostelecký, V.A., “Lorentzviolating extension of the standard model”, Phys. Rev. D, 58, 116002, (1998). [DOI], [arXiv:hepph/9809521]. (Cited on page 30.)ADSCrossRefGoogle Scholar
 [110]Collins, J., Perez, A. and Sudarsky, D., “Lorentz invariance violation and its role in Quantum Gravity phenomenology”, in Oriti, D., ed., Approaches to Quantum Gravity: Toward a New Understanding of Space, Time and Matter, pp. 528–547, (Cambridge University Press, Cambridge; New York, 2009). [arXiv:hepth/0603002]. (Cited on page 19.)CrossRefGoogle Scholar
 [111]Collins, J., Perez, A., Sudarsky, D., Urrutia, L. and Vucetich, H., “Lorentz Invariance and Quantum Gravity: An Additional FineTuning Problem?”, Phys. Rev. Lett., 93, 191301, (2004). [DOI], [ADS], [arXiv:grqc/0403053]. (Cited on page 19.)ADSMathSciNetCrossRefGoogle Scholar
 [112]Collins, N.A. and Hughes, S.A., “Towards a formalism for mapping the spacetimes of massive compact objects: Bumpy black holes and their orbits”, Phys. Rev. D, 69, 124022, (2004). [DOI], [ADS], [arXiv:grqc/0402063]. (Cited on pages 84, 86, and 87.)ADSMathSciNetCrossRefGoogle Scholar
 [113]Colpi, M., Shapiro, S.L. and Wasserman, I., “Boson Stars: Gravitational Equilibria of SelfInteracting Scalar Fields”, Phys. Rev. Lett., 57, 2485–2488, (1986). [DOI], [ADS]. (Cited on page 90.)ADSMathSciNetCrossRefGoogle Scholar
 [114]Connes, A., “Gravity coupled with matter and foundation of noncommutative geometry”, Commun. Math. Phys., 182, 155–176, (1996). [DOI], [arXiv:hepth/9603053]. (Cited on page 26.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [115]Contaldi, C.R., Magueijo, J. and Smolin, L., “Anomalous CosmicMicrowaveBackground Polarization and Gravitational Chirality”, Phys. Rev. Lett., 101, 141101, (2008). [DOI], [arXiv:0806.3082 [astroph]]. (Cited on page 28.)ADSCrossRefGoogle Scholar
 [116]Contopoulos, G., LukesGerakopoulos, G. and Apostolatos, T.A., “Orbits in a nonKerr Dynamical System”, Int. J. Bifurcat. Chaos, 21, 2261–2277, (2011). [ADS], [arXiv:1108.5057 [grqc]]. (Cited on pages 8 and 85.)CrossRefGoogle Scholar
 [117]Cooney, A., DeDeo, S. and Psaltis, D., “Gravity with Perturbative Constraints: Dark Energy Without New Degrees of Freedom”, Phys. Rev. D, 79, 044033, (2009). [DOI], [arXiv:0811.3635 [astroph]]. (Cited on page 14.)ADSCrossRefGoogle Scholar
 [118]Cooney, A., DeDeo, S. and Psaltis, D., “Neutron stars in f(R) gravity with perturbative constraints”, Phys. Rev. D, 82, 064033, (2010). [DOI], [arXiv:0910.5480 [astroph.HE]]. (Cited on page 14.)ADSCrossRefGoogle Scholar
 [119]Copi, C.J., Davis, A.N. and Krauss, L.M., “New Nucleosynthesis Constraint on the Variation of G”, Phys. Rev. Lett., 92, 171301, (2004). [DOI], [arXiv:astroph/0311334]. (Cited on page 25.)ADSCrossRefGoogle Scholar
 [120]Corbin, V. and Cornish, N.J., “Pulsar Timing Array Observations of Massive Black Hole Binaries”, arXiv, eprint, (2010). [ADS], [arXiv:1008.1782 [astroph.HE]]. (Cited on page 9.)Google Scholar
 [121]Corda, C., “Massive relic gravitational waves from f(R) theories of gravity: Production and potential detection”, Eur. Phys. J. C, 65, 257–267, (2010). [DOI], [arXiv: 1007.4077 [grqc]]. (Cited on page 45.)ADSCrossRefGoogle Scholar
 [122]Cornish, N.J. and Crowder, J., “LISA data analysis using MCMC methods”, Phys. Rev. D, 72, 043005, (2005). [DOI], [arXiv:grqc/0506059]. (Cited on page 41.)ADSCrossRefGoogle Scholar
 [123]Cornish, N.J. and Littenberg, T.B., “Tests of Bayesian model selection techniques for gravitational wave astronomy”, Phys. Rev. D, 76, 083006, (2007). [DOI], [arXiv:0704.1808 [grqc]]. (Cited on page 39.)ADSCrossRefGoogle Scholar
 [124]Cornish, N.J., Sampson, L., Yunes, N. and Pretorius, F., “Gravitational wave tests of general relativity with the parameterized postEinsteinian framework”, Phys. Rev. D, 84, 062003, (2011). [DOI], [ADS], [arXiv:1105.2088 [grqc]]. (Cited on pages 42, 60, 73, 77, 78, and 79.)ADSCrossRefGoogle Scholar
 [125]Cutler, C. and Flanagan, É.É., “Gravitational waves from merging compact binaries: How accurately can one extract the binary’s parameters from the inspiral wave form?”, Phys. Rev. D, 49, 2658–2697, (1994). [DOI], [arXiv:grqc/9402014]. (Cited on pages 39, 54, and 68.)ADSCrossRefGoogle Scholar
 [126]Cutler, C., Hiscock, W.A. and Larson, S.L., “LISA, binary stars, and the mass of the graviton”, Phys. Rev. D, 67, 024015, (2003). [DOI], [ADS], [arXiv:grqc/0209101]. (Cited on pages 19, 64, and 65.)ADSCrossRefGoogle Scholar
 [127]Cutler, C. and Vallisneri, M., “LISA detections of massive black hole inspirals: Parameter extraction errors due to inaccurate template waveforms”, Phys. Rev. D, 76, 104018, (2007). [DOI], [arXiv:0707.2982 [grqc]]. (Cited on page 42.)ADSCrossRefGoogle Scholar
 [128]Damour, T., “The general relativistic problem of motion and binary pulsars”, in Iyer, B.R., Kembhavi, A., Narlikar, J.V. and Vishveshwara, C.V., eds., Highlights in Gravitation and Cosmology, Proceedings of the Conference on Gravitation and Cosmology held in Goa, India, December 14–19, 1987, pp. 393–401, (Cambridge University Press, Cambridge; New York, 1988). (Cited on page 81.)Google Scholar
 [129]Damour, T. and EspositoFarèse, G., “Tensormultiscalar theories of gravitation”, Class. Quantum Grav., 9, 2093–2176, (1992). [DOI], [ADS]. (Cited on pages 14, 17, 53, and 57.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [130]Damour, T. and EspositoFarèse, G., “Nonperturbative strongfield effects in tensorscalar theories of gravitation”, Phys. Rev. Lett., 70, 2220–2223, (1993). [DOI], [ADS]. (Cited on pages 17, 53, and 57.)ADSCrossRefGoogle Scholar
 [131]Damour, T. and EspositoFarèse, G., “Tensorscalar gravity and binary pulsar experiments”, Phys. Rev. D, 54, 1474–1491, (1996). [DOI], [arXiv:grqc/9602056]. (Cited on pages 16 and 17.)ADSCrossRefGoogle Scholar
 [132]Damour, T. and EspositoFarèse, G., “Gravitationalwave versus binarypulsar tests of strongfield gravity”, Phys. Rev. D, 58, 042001, (1998). [DOI], [ADS], [arXiv:grqc/9803031]. (Cited on pages 16 and 17.)ADSCrossRefGoogle Scholar
 [133]Damour, T. and Polyakov, A.M., “The string dilaton and a least coupling principle”, Nucl. Phys. B, 423, 532–558, (1994). [DOI], [arXiv:hepth/9401069]. (Cited on page 16.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [134]Damour, T. and Polyakov, A.M., “String theory and gravity”, Gen. Relativ. Gravit., 26, 1171–1176, (1994). [DOI], [arXiv:grqc/9411069]. (Cited on page 16.)ADSMathSciNetCrossRefGoogle Scholar
 [135]De Felice, A. and Tsujikawa, S., “f(R) Theories”, Living Rev. Relativity, 13, lrr20103 (2010). [DOI], [ADS], [arXiv:1002.4928 [grqc]]. URL (accessed 15 April 2013): http://www.livingreviews.org/lrr20103. (Cited on page 16.)
 [136]de Rham, C., Gabadadze, G. and Tolley, A.J., “Resummation of Massive Gravity”, Phys. Rev. Lett., 106, 231101, (2011). [DOI], [ADS], [arXiv:1011.1232 [hepth]]. (Cited on pages 18 and 30.)ADSCrossRefGoogle Scholar
 [137]de Rham, C., Matas, A. and Tolley, A.J., “Galileon Radiation from Binary Systems”, Phys. Rev. D, 87, 064024, (2013). [DOI], [arXiv:1212.5212 [hepth]]. (Cited on pages 12 and 18.)ADSCrossRefGoogle Scholar
 [138]de Rham, C., Tolley, A.J. and Wesley, D.H., “Vainshtein mechanism in binary pulsars”, Phys. Rev. D, 87, 044025, (2013). [DOI], [arXiv: 1208.0580 [grqc]]. (Cited on pages 12 and 18.)ADSCrossRefGoogle Scholar
 [139]DeDeo, S. and Psaltis, D., “Towards New Tests of Strongfield Gravity with Measurements of Surface Atomic Line Redshifts from Neutron Stars”, Phys. Rev. Lett., 90, 141101, (2003). [DOI], [ADS], [arXiv:astroph/0302095]. (Cited on page 16.)ADSCrossRefGoogle Scholar
 [140]Deffayet, C., Dvali, G., Gabadadze, G. and Vainshtein, A.I., “Nonperturbative continuity in graviton mass versus perturbative discontinuity”, Phys. Rev. D, 65, 044026, (2002). [DOI], [arXiv:hepth/0106001]. (Cited on pages 12 and 18.)ADSCrossRefGoogle Scholar
 [141]Deffayet, C. and Menou, K., “Probing Gravity with Spacetime Sirens”, Astrophys. J., 668, L143–L146 (2007). [DOI], [arXiv:0709.0003 [astroph]]. (Cited on page 24.)ADSCrossRefGoogle Scholar
 [142]Del Pozzo, W., Veitch, J. and Vecchio, A., “Testing general relativity using Bayesian model selection: Applications to observations of gravitational waves from compact binary systems”, Phys. Rev. D, 83, 082002, (2011). [DOI], [ADS], [arXiv:1101.1391 [grqc]]. (Cited on pages 42 and 65.)ADSCrossRefGoogle Scholar
 [143]Deller, A.T., Verbiest, J.P.W., Tingay, S.J. and Bailes, M., “Extremely High Precision VLBI Astrometry of PSR J0437–4715 and Implications for Theories of Gravity”, Astrophys. J. Lett., 685, L67–L70 (2008). [DOI], [arXiv:0808.1594 [astroph]]. (Cited on page 25.)ADSCrossRefGoogle Scholar
 [144]Delsate, T., Cardoso, V. and Pani, P., “Anti de Sitter black holes and branes in dynamical ChernSimons gravity: perturbations, stability and the hydrodynamic modes”, J. High Energy Phys., 2011(06), 055, (2011). [DOI], [arXiv:1103.5756 [hepth]]. (Cited on page 23.)MathSciNetzbMATHCrossRefGoogle Scholar
 [145]Detweiler, S., “Pulsar timing measurements and the search for gravitational waves”, Astrophys. J., 234, 1100–1104, (1979). [DOI], [ADS]. (Cited on page 34.)ADSCrossRefGoogle Scholar
 [146]Detweiler, S.L., “Black Holes and Gravitational Waves. III. The Resonant Frequencies of Rotating Holes”, Astrophys. J., 239, 292–295, (1980). [DOI], [ADS]. (Cited on page 89.)ADSCrossRefGoogle Scholar
 [147]Detweiler, S.L., “KleinGordon Equation and Rotating Black Holes”, Phys. Rev. D, 22, 2323–2326, (1980). [DOI], [ADS]. (Cited on page 78.)ADSCrossRefGoogle Scholar
 [148]Dilkes, F.A., Duff, M.J., Liu, J.T. and Sati, H., “Quantum discontinuity between zero and infinitesimal graviton mass with a Lambda term”, Phys. Rev. Lett., 87, 041301, (2001). [DOI], [arXiv:hepth/0102093]. (Cited on page 63.)ADSMathSciNetCrossRefGoogle Scholar
 [149]Dirac, P.A.M., “The Cosmological Constants”, Nature, 139, 323, (1937). [DOI], [ADS]. (Cited on page 67.)ADSzbMATHCrossRefGoogle Scholar
 [150]Douchin, F. and Haensel, P., “A unified equation of state of dense matter and neutron star structure”, Astron. Astrophys., 380, 151–167, (2001). [DOI], [ADS], [arXiv:astroph/0111092]. (Cited on page 81.)ADSCrossRefGoogle Scholar
 [151]Drake, S.P. and Szekeres, P., “Uniqueness of the NewmanJanis Algorithm in Generating the KerrNewman Metric”, Gen. Relativ. Gravit., 32, 445–458, (2000). [DOI], [arXiv:grqc/9807001]. (Cited on page 87.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [152]Dreyer, O., Kelly, B.J., Krishnan, B., Finn, L.S., Garrison, D. and LopezAleman, R., “Blackhole spectroscopy: Testing general relativity through gravitationalwave observations”, Class. Quantum Grav., 21, 787–804, (2004). [DOI], [ADS], [arXiv:grqc/0309007]. (Cited on page 89.)ADSzbMATHCrossRefGoogle Scholar
 [153]Droz, S., Knapp, D.J., Poisson, E. and Owen, B.J., “Gravitational waves from inspiraling compact binaries: Validity of the stationary phase approximation to the Fourier transform”, Phys. Rev. D, 59, 124016, (1999). [DOI], [arXiv:grqc/9901076]. (Cited on pages 54 and 68.)ADSCrossRefGoogle Scholar
 [154]Dubeibe, F.L., Pachón, L.A. and SanabriaGómez, Jose D., “Chaotic dynamics around astrophysical objects with nonisotropic stresses”, Phys. Rev. D, 75, 023008, (2007). [DOI], [ADS], [arXiv:grqc/0701065]. (Cited on page 85.)ADSCrossRefGoogle Scholar
 [155]Dubovsky, S., Tinyakov, P. and Zaldarriaga, M., “Bumpy black holes from spontaneous Lorentz violation”, J. High Energy Phys., 2007(11), 083, (2007). [DOI], [arXiv:0706.0288 [hepth]]. (Cited on pages 12 and 87.)MathSciNetzbMATHCrossRefGoogle Scholar
 [156]Dunkley, J. et al. (WMAP Collaboration), “FiveYear Wilkinson Microwave Anisotropy Probe Observations: Likelihoods and Parameters from the WMAP Data”, Astrophys. J. Suppl. Ser., 180, 306–329, (2009). [DOI], [arXiv:0803.0586 [astroph]]. (Cited on page 69.)ADSCrossRefGoogle Scholar
 [157]Dvali, G., Gabadadze, G. and Porrati, M., “4D gravity on a brane in 5D Minkowski space”, Phys. Lett. B, 485, 208–214, (2000). [DOI], [arXiv:hepth/0005016]. (Cited on page 17.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [158]Dyda, S., Flanagan, É.É. and Kamionkowski, M., “Vacuum Instability in ChernSimons Gravity”, Phys. Rev. D, 86, 124031, (2012). [DOI], [ADS], [arXiv: 1208.4871 [grqc]]. (Cited on page 73.)ADSCrossRefGoogle Scholar
 [159]Dykla, J.J., Conserved quantities and the formation of black holes in the BransDicke Theory of Gravitation, Ph.D. thesis, (California Institute of Technology, Pasadena, CA, 1972). [ADS]. (Cited on pages 16 and 52.)Google Scholar
 [160]Eardley, D.M., “Observable effects of a scalar gravitational field in a binary pulsar”, Astrophys. J. Lett., 196, L59–L62 (1975). [DOI], [ADS]. (Cited on pages 51 and 52.)ADSCrossRefGoogle Scholar
 [161]Eardley, D.M., Lee, D.L. and Lightman, A.P., “GravitationalWave Observations as a Tool for Testing Relativistic Gravity”, Phys. Rev. D, 8, 3308–3321 (1973). [DOI], [ADS]. (Cited on pages 16, 22, and 27.)ADSCrossRefGoogle Scholar
 [162]Ellis, J.A., Siemens, X. and van Haasteren, R., “An Efficient Approximation to the Likelihood for Gravitational Wave Stochastic Background Detection Using Pulsar Timing Data”, Astrophys. J., 769, 63, (2013). [DOI], [ADS], [arXiv:1302.1903 [astroph.IM]]. (Cited on page 49.)ADSCrossRefGoogle Scholar
 [163]Emparan, R., Fabbri, A. and Kaloper, N., “Quantum black holes as holograms in AdS brane worlds”, J. High Energy Phys., 2002(08), 043, (2002). [DOI], [ADS], [arXiv:hepth/0206155]. (Cited on page 25.)MathSciNetzbMATHCrossRefGoogle Scholar
 [164]Faraoni, V., “Illusions of general relativity in BransDicke gravity”, Phys. Rev. D, 59, 084021, (1999). [DOI], [arXiv:grqc/9902083]. (Cited on pages 16 and 52.)ADSMathSciNetCrossRefGoogle Scholar
 [165]Faraoni, V. and Gunzig, E., “Einstein frame or Jordan frame?”, Int. J. Theor. Phys., 38, 217–225, (1999). [DOI], [arXiv:astroph/9910176]. (Cited on page 14.)MathSciNetzbMATHCrossRefGoogle Scholar
 [166]Faraoni, V., Gunzig, E. and Nardone, P., “Conformal transformations in classical gravitational theories and in cosmology”, Fundam. Cosmic Phys., 20, 121–175, (1999). [arXiv:grqc/9811047]. (Cited on page 14.)ADSGoogle Scholar
 [167]Feroz, F., Gair, J.R., Hobson, M.P. and Porter, E.K., “Use of the MultiNest algorithm for gravitational wave data analysis”, Class. Quantum Grav., 26, 215003, (2009). [DOI], [ADS], [arXiv:0904.1544 [grqc]]. (Cited on page 42.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [168]Ferrari, V., Gualtieri, L. and Maselli, A., “Tidal interaction in compact binaries: a postNewtonian affine framework”, Phys. Rev. D, 85, 044045, (2012). [DOI], [arXiv:1 111.6607 [grqc]]. (Cited on page 82.)ADSCrossRefGoogle Scholar
 [169]Fierz, M. and Pauli, W., “On relativistic wave equations for particles of arbitrary spin in an electromagnetic field”, Proc. R. Soc. London, Ser. A, 173, 211–232, (1939). [DOI], [ADS]. (Cited on pages 17, 19, 63, 64, and 65.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [170]Figueras, P., Lucietti, J. and Wiseman, T., “Ricci solitons, Ricci flow, and strongly coupled CFT in the Schwarzschild Unruh or Boulware vacua”, Class. Quantum Grav., 28, 215018, (2011). [DOI], [arXiv:1104.4489 [hepth]]. (Cited on page 25.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [171]Figueras, P. and Tunyasuvunakool, S., “CFTs in rotating black hole backgrounds”, Class. Quantum Grav., 30, 125015, (2013). [DOI], [arXiv:1304.1162 [hepth]]. (Cited on page 25.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [172]Figueras, P. and Wiseman, T., “Gravity and large black holes in RandallSundrum II braneworlds”, Phys. Rev. Lett., 107, 081101, (2011). [DOI], [arXiv:1105.2558 [hepth]]. (Cited on page 25.)ADSCrossRefGoogle Scholar
 [173]Finn, L.S. and Chernoff, D.F., “Observing binary inspiral in gravitational radiation: One interferometer”, Phys. Rev. D, 47, 2198–2219, (1993). [DOI], [arXiv:grqc/9301003]. (Cited on page 39.)ADSCrossRefGoogle Scholar
 [174]Finn, L.S. and Sutton, P.J., “Bounding the mass of the graviton using binary pulsar observations”, Phys. Rev. D, 65, 044022, (2002). [DOI], [ADS], [arXiv:grqc/0109049]. (Cited on pages 19, 39, 64, and 65.)ADSzbMATHCrossRefGoogle Scholar
 [175]Flanagan, É.É. and Hinderer, T., “Constraining neutron star tidal Love numbers with gravitational wave detectors”, Phys. Rev. D, 77, 021502, (2008). [DOI], [ADS], [arXiv:0709.1915 [astroph]]. (Cited on page 81.)ADSCrossRefGoogle Scholar
 [176]Fradkin, E.S. and Tseytlin, A.A., “Quantum string theory effective action”, Nucl. Phys. B, 261, 1–27, (1985). [DOI], [ADS]. (Cited on page 16.)ADSMathSciNetCrossRefGoogle Scholar
 [177]Freire, P.C.C. et al., “The relativistic pulsarwhite dwarf binary PSR J1738+0333 — II. The most stringent test of scalartensor gravity”, Mon. Not. R. Astron. Soc., 423, 3328–3343, (2012). [DOI], [ADS], [arXiv: 1205.1450 [astroph.GA]]. (Cited on pages 16 and 17.)ADSCrossRefGoogle Scholar
 [178]Friedberg, R., Lee, T.D. and Pang, Y., “Minisoliton stars”, Phys. Rev. D, 35, 3640–3657, (1987). [DOI], [ADS]. (Cited on page 90.)ADSCrossRefGoogle Scholar
 [179]Friedberg, R., Lee, T.D. and Pang, Y., “Scalar soliton stars and black holes”, Phys. Rev. D, 35, 3658–3677, (1987). [DOI], [ADS]. (Cited on page 90.)ADSCrossRefGoogle Scholar
 [180]Frolov, A.V. and Guo, J.Q., “Small Cosmological Constant from Running Gravitational Coupling”, arXiv, eprint, (2011). [ADS], [arXiv:1101.4995 [astroph.CO]]. (Cited on pages 24 and 26.)Google Scholar
 [181]Fujii, Y. and Maeda, K.I., The ScalarTensor Theory of Gravitation, Cambridge Monographs on Mathematical Physics, (Cambridge University Press, Cambridge; New York, 2003). [Google Books]. (Cited on page 14.)zbMATHGoogle Scholar
 [182]Gair, J.R., Li, C. and Mandel, I., “Observable properties of orbits in exact bumpy spacetimes”, Phys. Rev. D, 77, 024035, (2008). [DOI], [ADS], [arXiv:0708.0628 [grqc]]. (Cited on pages 8 and 85.)ADSCrossRefGoogle Scholar
 [183]Gair, J.R., Vallisneri, M., Larson, S.L. and Baker, J.G., “Testing General Relativity with LowFrequency, SpaceBased GravitationalWave Detectors”, Living Rev. Relativity, 16, lrr20137 (2013). [DOI], [ADS], [arXiv:1212.5575 [grqc]]. URL (accessed 10 October 2013): http://www.livingreviews.org/lrr20137. (Cited on page 8.)
 [184]Gair, J.R. and Yunes, N., “Approximate waveforms for extrememassratio inspirals in modified gravity spacetimes”, Phys. Rev. D, 84, 064016, (2011). [DOI], [ADS], [arXiv:1106.6313 [grqc]]. (Cited on pages 8, 84, and 88.)ADSCrossRefGoogle Scholar
 [185]Gambini, R., Rastgoo, S. and Pullin, J., “Small Lorentz violations in quantum gravity: do they lead to unacceptably large effects?”, Class. Quantum Grav., 28, 155005, (2011). [DOI], [arXiv:1106.1417 [grqc]]. (Cited on page 19.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [186]Garattini, R., “Modified dispersion relations and noncommutative geometry lead to a finite Zero Point Energy”, in Kouneiher, J., Barbachoux, C., Masson, T. and Vey, D., eds., Frontiers of Fundamental Physics: The Eleventh International Symposium, Paris, France, 6–9 July 2010, AIP Conference Proceedings, 1446, pp. 298–310, (American Institute of Physics, Melville, NY, 2011). [DOI], [ADS], [arXiv:1102.0117 [grqc]]. (Cited on page 19.)Google Scholar
 [187]Garattini, R. and Mandanici, G., “Modified dispersion relations lead to a finite zero point gravitational energy”, Phys. Rev. D, 83, 084021, (2011). [DOI], [arXiv:1102.3803 [grqc]]. (Cited on page 19.)ADSCrossRefGoogle Scholar
 [188]Garattini, R. and Mandanici, G., “Particle propagation and effective spacetime in gravity’s rainbow”, Phys. Rev. D, 85, 023507, (2012). [DOI], [arXiv:1109.6563 [grqc]]. (Cited on page 19.)ADSCrossRefGoogle Scholar
 [189]Garay, L.J. and GarcíaBellido, J., “JordanBransDicke quantum wormholes and Coleman’s mechanism”, Nucl. Phys. B, 400, 416–434, (1993). [DOI], [arXiv:grqc/9209015]. (Cited on page 16.)ADSzbMATHCrossRefGoogle Scholar
 [190]Garfinkle, D., Pretorius, F. and Yunes, N., “Linear stability analysis and the speed of gravitational waves in dynamical ChernSimons modified gravity”, Phys. Rev. D, 82, 041501, (2010). [DOI], [arXiv:1007.2429 [grqc]]. (Cited on pages 23 and 58.)ADSCrossRefGoogle Scholar
 [191]Gasperini, M. and Ungarelli, C., “Detecting a relic background of scalar waves with LIGO”, Phys. Rev. D, 64, 064009, (2001). [DOI], [arXiv:grqc/0103035]. (Cited on page 45.)ADSCrossRefGoogle Scholar
 [192]Gates Jr, S.J., Ketov, S.V. and Yunes, N., “Seeking the Loop Quantum Gravity BarberoImmirzi Parameter and Field in 4D, \({\mathcal N} = 1\) Supergravity”, Phys. Rev. D, 80, 065003, (2009). [DOI], [arXiv:0906.4978 [hepth]]. (Cited on pages 20 and 28.)ADSMathSciNetCrossRefGoogle Scholar
 [193]Gehrels, N. et al. (Swift team), “The Swift GammaRay Burst Mission”, in Fenimore, E. and Galassi, M., eds., GammaRay Bursts: 30 Years of Discovery, GammaRay Burst Symposium, Santa Fe, NM, USA, 8–12 September 2003, AIP Conference Proceedings, 727, pp. 637–641, (American Institute of Physics, Melville, NY, 2004). [DOI], [arXiv:astroph/0405233]. (Cited on page 72.)Google Scholar
 [194]Geroch, R., “Multipole moments. I. Flat space”, J. Math. Phys., 11, 1955–1961, (1970). [DOI], [ADS]. (Cited on page 84.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [195]Geroch, R., “Multipole moments. II. Curved space”, J. Math. Phys., 11, 2580–2588, (1970). [DOI], [ADS]. (Cited on page 84.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [196]Glampedakis, K. and Babak, S., “Mapping spacetimes with LISA: Inspiral of a testbody in a ‘quasiKerr’ field”, Class. Quantum Grav., 23, 4167–4188, (2006). [DOI], [ADS], [arXiv:grqc/0510057]. (Cited on pages 8, 84, and 87.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [197]Goenner, H., “Some remarks on the genesis of scalartensor theories”, Gen. Relativ. Gravit., 44, 2077–2097, (2012). [DOI], [arXiv: 1204.3455 [grqc]]. (Cited on page 14.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [198]Goldberger, W.D. and Rothstein, I.Z., “Effective field theory of gravity for extended objects”, Phys. Rev. D, 73, 104029, (2006). [DOI], [arXiv:hepth/0409156]. (Cited on page 51.)ADSMathSciNetCrossRefGoogle Scholar
 [199]Goldberger, W.D. and Rothstein, I.Z., “Towers of gravitational theories”, Gen. Relativ. Gravit., 38, 1537–1546, (2006). [DOI], [ADS], [arXiv:hepth/0605238]. (Cited on page 51.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [200]Goldhaber, A.S. and Nieto, M.M., “Mass of the graviton”, Phys. Rev. D, 9, 1119–1121, (1974). [DOI], [ADS]. (Cited on page 18.)ADSCrossRefGoogle Scholar
 [201]Gossan, S., Veitch, J. and Sathyaprakash, B.S., “Bayesian model selection for testing the nohair theorem with black hole ringdowns”, Phys. Rev. D, 85, 124056, (2012). [DOI], [ADS], [arXiv:1111.5819 [grqc]]. (Cited on page 89.)ADSCrossRefGoogle Scholar
 [202]Gralla, S.E., “Motion of small bodies in classical field theory”, Phys. Rev. D, 81, 084060, (2010). [DOI], [arXiv:1002.5045 [grqc]]. (Cited on page 52.)ADSMathSciNetCrossRefGoogle Scholar
 [203]Green, M.B., Schwarz, J.H. and Witten, E., Superstring Theory. Vol 1: Introduction, Cambridge Monographs on Mathematical Physics, (Cambridge University Press, Cambridge; New York, 1987). (Cited on pages 20 and 22.)zbMATHGoogle Scholar
 [204]Green, M.B., Schwarz, J.H. and Witten, E., Superstring Theory. Vol 2: Loop Amplitudes, Anomalies and Phenomenology, Cambridge Monographs on Mathematical Physics, (Cambridge University Press, Cambridge; New York, 1987). (Cited on pages 20, 22, and 28.)zbMATHGoogle Scholar
 [205]Gregory, P.C., Bayesian Logical Data Analysis for the Physical Sciences: A Comparative Approach with ‘Mathematica’ Support, (Cambridge University Press, Cambridge; New York, 2005). [ADS], [Google Books]. (Cited on page 39.)zbMATHCrossRefGoogle Scholar
 [206]Groenewold, H.J., “On the principles of elementary quantum mechanics”, Physica, 12, 405–460, (1946). [DOI], [ADS]. (Cited on page 26.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [207]Grumiller, D. and Yunes, N., “How do black holes spin in ChernSimons modified gravity?”, Phys. Rev. D, 77, 044015, (2008). [DOI], [ADS], [arXiv:0711.1868 [grqc]]. (Cited on pages 22, 29, and 30.)ADSMathSciNetCrossRefGoogle Scholar
 [208]Guenther, D.B., Krauss, L.M. and Demarque, P., “Testing the Constancy of the Gravitational Constant Using Helioseismology”, Astrophys. J., 498, 871–876, (1998). [DOI], [ADS]. (Cited on page 25.)ADSCrossRefGoogle Scholar
 [209]Guéron, E. and Letelier, P.S., “Chaos in pseudoNewtonian black holes with halos”, Astron. Astrophys., 368, 716–720, (2001). [DOI], [ADS], [arXiv:astroph/0101140]. (Cited on page 85.)ADSzbMATHCrossRefGoogle Scholar
 [210]Guéron, E. and Letelier, P.S., “Geodesic chaos around quadrupolar deformed centers of attraction”, Phys. Rev. E, 66, 046611, (2002). [DOI], [ADS]. (Cited on page 85.)ADSMathSciNetCrossRefGoogle Scholar
 [211]Gümrükçüoğlu, A.E., Kuroyanagi, S., Lin, C., Mukohyama, S. and Tanahashi, N., “Gravitational wave signal from massive gravity”, Class. Quantum Grav., 29, 235026, (2012). [DOI], [arXiv: 1208.5975 [hepth]]. (Cited on page 18.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [212]Gürsel, Y. and Tinto, M., “Near optimal solution to the inverse problem for gravitationalwave bursts”, Phys. Rev. D, 40, 3884–3938, (1989). [DOI], [ADS]. (Cited on pages 43, 45, and 81.)ADSCrossRefGoogle Scholar
 [213]Hansen, R.O., “Multipole moments of stationary spacetimes”, J. Math. Phys., 15, 46–52, (1974). [DOI], [ADS]. (Cited on page 84.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [214]Harada, T., “Stability analysis of spherically symmetric star in scalar — tensor theories of gravity”, Prog. Theor. Phys., 98, 359–379, (1997). [DOI], [arXiv:grqc/9706014]. (Cited on page 16.)ADSCrossRefGoogle Scholar
 [215]Harada, T., “Neutron stars in scalar tensor theories of gravity and catastrophe theory”, Phys. Rev. D, 57, 4802–4811, (1998). [DOI], [arXiv:grqc/9801049]. (Cited on page 16.)ADSMathSciNetCrossRefGoogle Scholar
 [216]Harada, T., Chiba, T., Nakao, K.I. and Nakamura, T., “Scalar gravitational wave from OppenheimerSnyder collapse in scalartensor theories of gravity”, Phys. Rev. D, 55, 2024–2037, (1997). [DOI], [arXiv:grqc/9611031]. (Cited on page 43.)ADSCrossRefGoogle Scholar
 [217]Harry, G.M. (LIGO Scientific Collaboration), “Advanced LIGO: The next generation of gravitational wave detectors”, Class. Quantum Grav., 27, 084006, (2010). [DOI], [ADS]. (Cited on page 8.)ADSMathSciNetCrossRefGoogle Scholar
 [218]Hartle, J.B. and Thorne, K.S., “Slowly Rotating Relativistic Stars. II. Models for Neutron Stars and Supermassive Stars”, Astrophys. J., 153, 807–834, (1968). [DOI], [ADS]. (Cited on page 87.)ADSCrossRefGoogle Scholar
 [219]Hassan, S.F. and Rosen, R.A., “Bimetric Gravity from Ghostfree Massive Gravity”, J. High Energy Phys., 2012(02), 126, (2012). [DOI], [arXiv:1109.3515 [hepth]]. (Cited on pages 17 and 30.)MathSciNetzbMATHCrossRefGoogle Scholar
 [220]Hassan, S.F. and Rosen, R.A., “Confirmation of the Secondary Constraint and Absence of Ghost in Massive Gravity and Bimetric Gravity”, J. High Energy Phys., 2012(04), 123, (2012). [DOI], [arXiv:11 11.2070 [hepth]]. (Cited on pages 17 and 30.)MathSciNetCrossRefGoogle Scholar
 [221]Hastings, W.K., “Monte Carlo sampling methods using Markov chains and their applications”, Biometrika, 57, 97–109, (1970). [DOI]. (Cited on page 41.)MathSciNetzbMATHCrossRefGoogle Scholar
 [222]Hawking, S.W., “Gravitational Radiation from Colliding Black Holes”, Phys. Rev. Lett., 26, 1344–1346, (1971). [DOI], [ADS]. (Cited on pages 16 and 83.)ADSCrossRefGoogle Scholar
 [223]Hawking, S.W., “Black Holes in General Relativity”, Commun. Math. Phys., 25, 152–166, (1972). [DOI], [ADS]. (Cited on page 83.)ADSMathSciNetCrossRefGoogle Scholar
 [224]Hawking, S.W., “Black Holes in the BransDicke Theory of Gravitation”, Commun. Math. Phys., 25, 167–171, (1972). [DOI], [ADS]. (Cited on pages 16 and 52.)ADSMathSciNetCrossRefGoogle Scholar
 [225]Hawking, S.W. and Ellis, G.F.R., The Large Scale Structure of SpaceTime, Cambridge Monographs on Mathematical Physics, (Cambridge University Press, Cambridge; New York, 1973). [ADS], [Google Books]. (Cited on page 12.)zbMATHCrossRefGoogle Scholar
 [226]Hawking, S.W. and Hartle, J.B., “Energy and angular momentum flow into a black hole”, Commun. Math. Phys., 27, 283–290, (1972). [DOI], [ADS]. (Cited on page 83.)ADSMathSciNetCrossRefGoogle Scholar
 [227]Hawking, S.W. and Israel, W., eds., Three Hundred Years of Gravitation, (Cambridge University Press, Cambridge; New York, 1987). [Google Books]. (Cited on page 81.)zbMATHGoogle Scholar
 [228]Hayama, K. and Nishizawa, A., “Modelindependent test of gravity with a network of groundbased gravitationalwave detectors”, Phys. Rev. D, 87, 062003, (2013). [DOI], [ADS], [arXiv: 1208.4596 [grqc]]. (Cited on pages 43, 44, and 81.)ADSCrossRefGoogle Scholar
 [229]Hayasaki, K., Yagi, K., Tanaka, T. and Mineshige, S., “Gravitational wave diagnosis of a circumbinary disk”, Phys. Rev. D, 87, 044051, (2013). [DOI], [arXiv:1201.2858 [astroph.CO]]. (Cited on pages 69 and 80.)ADSCrossRefGoogle Scholar
 [230]Healy, J., Bode, T., Haas, R., Pazos, E., Laguna, P., Shoemaker, D.M. and Yunes, N., “Late Inspiral and Merger of Binary Black Holes in ScalarTensor Theories of Gravity”, arXiv, eprint, (2011). [ADS], [arXiv:1112.3928 [grqc]]. (Cited on pages 16 and 52.)Google Scholar
 [231]Hellings, R.W. and Downs, G.S., “Upper limits on the isotropic gravitational radiation background from pulsar timing analysis”, Astrophys. J. Lett., 265, L39–L42 (1983). [DOI], [ADS]. (Cited on page 48.)ADSCrossRefGoogle Scholar
 [232]Hinterbichler, K., “Theoretical Aspects of Massive Gravity”, Rev. Mod. Phys., 84, 671–710, (2012). [DOI], [arXiv:1105.3735 [hepth]]. (Cited on page 17.)ADSCrossRefGoogle Scholar
 [233]Hořava, P., “Membranes at quantum criticality”, J. High Energy Phys., 2009(03), 020, (2009). [DOI], [arXiv:0812.4287 [hepth]]. (Cited on page 19.)ADSMathSciNetCrossRefGoogle Scholar
 [234]Hořava, P., “Quantum gravity at a Lifshitz point”, Phys. Rev. D, 79, 084008, (2009). [DOI], [arXiv:0901.3775 [hepth]]. (Cited on pages 19 and 30.)ADSMathSciNetCrossRefGoogle Scholar
 [235]Horbatsch, M.W. and Burgess, C.P., “SemiAnalytic Stellar Structure in ScalarTensor Gravity”, J. Cosmol. Astropart. Phys., 2011(08), 027, (2011). [DOI], [arXiv: 1006.4411 [grqc]]. (Cited on pages 16 and 53.)CrossRefGoogle Scholar
 [236]Horbatsch, M.W. and Burgess, C.P., “Cosmic blackhole hair growth and quasar OJ287”, J. Cosmol. Astropart. Phys., 2012(05), 010, (2012). [DOI], [ADS], [arXiv: 1111.4009 [grqc]]. (Cited on page 25.)CrossRefGoogle Scholar
 [237]Hoyle, C.D., Kapner, D.J., Heckel, B.R., Adelberger, E.G., Gundlach, J.H., Schmidt, U. and Swanson, H.E., “Submillimeter tests of the gravitational inversesquare law”, Phys. Rev. D, 70, 042004, (2004). [DOI], [arXiv:hepph/0405262]. (Cited on page 21.)ADSCrossRefGoogle Scholar
 [238]Hughes, S.A., “Evolution of circular, nonequatorial orbits of Kerr black holes due to gravitationalwave emission”, Phys. Rev. D, 61, 084004, (2000). [DOI], [arXiv:grqc/9910091]. Errata: 10.1103/PhysRevD.63.049902, 10.1103/PhysRevD.65.069902, 10.1103/PhysRevD.67.089901, 10.1103/PhysRevD.78.109902. (Cited on page 91.)ADSMathSciNetCrossRefGoogle Scholar
 [239]Hughes, S.A., “Evolution of circular, nonequatorial orbits of Kerr black holes due to gravitationalwave emission. II. Inspiral trajectories and gravitational waveforms”, Phys. Rev. D, 64, 064004, (2001). [DOI], [arXiv:grqc/0104041]. (Cited on page 91.)ADSCrossRefGoogle Scholar
 [240]Huwyler, C., Klein, A. and Jetzer, P., “Testing general relativity with LISA including spin precession and higher harmonics in the waveform”, Phys. Rev. D, 86, 084028, (2012). [DOI], [ADS], [arXiv:1108.1826 [grqc]]. (Cited on page 74.)ADSCrossRefGoogle Scholar
 [241]Isaacson, R.A., “Gravitational Radiation in the Limit of High Frequency. I. The Linear Approximation and Geometrical Optics”, Phys. Rev., 166, 1263–1271, (1968). [DOI], [ADS]. (Cited on page 62.)ADSCrossRefGoogle Scholar
 [242]Isaacson, R.A., “Gravitational Radiation in the Limit of High Frequency. II. Nonlinear Terms and the Effective Stress Tensor”, Phys. Rev., 166, 1272–1279, (1968). [DOI], [ADS]. (Cited on page 62.)ADSCrossRefGoogle Scholar
 [243]Israel, W., “Event Horizons in Static Vacuum SpaceTimes”, Phys. Rev., 164, 1776–1779, (1967). [DOI], [ADS]. (Cited on page 83.)ADSCrossRefGoogle Scholar
 [244]Israel, W., “Event Horizons in Static Electrovac SpaceTimes”, Commun. Math. Phys., 8, 245–260, (1968). [DOI], [ADS]. (Cited on pages 16 and 83.)ADSMathSciNetCrossRefGoogle Scholar
 [245]Jackiw, R. and Pi, S.Y., “ChernSimons modification of general relativity”, Phys. Rev. D, 68, 104012, (2003). [DOI], [ADS], [arXiv:grqc/0308071]. (Cited on pages 21, 23, 29, and 69.)ADSMathSciNetCrossRefGoogle Scholar
 [246]Jacobson, T., “Primordial black hole evolution in tensor scalar cosmology”, Phys. Rev. Lett., 83, 2699–2702, (1999). [DOI], [arXiv:astroph/9905303]. (Cited on pages 25 and 53.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [247]Jacobson, T., “Einsteinæther gravity: A status report”, in From Quantum to Emergent Gravity: Theory and Phenomenology, June 11–15 2007, Trieste, Italy, Proceedings of Science, PoS(QGPh)020, (SISSA, Trieste, 2008). [arXiv:0801.1547 [grqc]]. URL (accessed 15 April 2013): http://pos.sissa.it/cgibin/reader/conf.cgi?confid=43. (Cited on page 30.)Google Scholar
 [248]Jaranowski, P. and Królak, A., “GravitationalWave Data Analysis. Formalism and Sample Applications: The Gaussian Case”, Living Rev. Relativity, 15, lrr20124 (2012). [DOI], [ADS]. URL (accessed 15 April 2013): http://www.livingreviews.org/lrr20124. (Cited on page 39.)
 [249]Jofré, P., Reisenegger, A. and Fernández, R., “Constraining a Possible Time Variation of the Gravitational Constant through ‘Gravitochemical Heating’ of Neutron Stars”, Phys. Rev. Lett., 97, 131102, (2006). [DOI], [arXiv:astroph/0606708]. (Cited on page 25.)ADSCrossRefGoogle Scholar
 [250]Johannsen, T. and Psaltis, D., “Testing the NoHair Theorem with Observations in the Electromagnetic Spectrum. I. Properties of a QuasiKerr Spacetime”, Astrophys. J., 716, 187–197, (2010). [DOI], [arXiv:1003.3415 [astroph.HE]]. (Cited on page 84.)ADSCrossRefGoogle Scholar
 [251]Johannsen, T. and Psaltis, D., “Testing the NoHair Theorem with Observations in the Electromagnetic Spectrum. II. Black Hole Images”, Astrophys. J., 718, 446–454, (2010). [DOI], [arXiv: 1005.1931 [astroph.HE]]. (Cited on page 84.)ADSCrossRefGoogle Scholar
 [252]Johannsen, T. and Psaltis, D., “Metric for rapidly spinning black holes suitable for strongfield tests of the nohair theorem”, Phys. Rev. D, 83, 124015, (2011). [DOI], [arXiv:1105.3191 [grqc]]. (Cited on page 87.)ADSCrossRefGoogle Scholar
 [253]Johannsen, T. and Psaltis, D., “Testing the NoHair Theorem with Observations in the Electromagnetic Spectrum. III. QuasiPeriodic Variability”, Astrophys. J., 726, 11, (2011). [DOI], [arXiv:1010.1000 [astroph.HE]]. (Cited on page 84.)ADSCrossRefGoogle Scholar
 [254]Johannsen, T. and Psaltis, D., “Testing the NoHair Theorem with Observations in the Electromagnetic Spectrum. IV. Relativistically Broadened Iron Lines”, Astrophys. J., 773, 57, (2013). [DOI], [ADS], [arXiv:1202.6069 [astroph.HE]]. (Cited on page 84.)ADSCrossRefGoogle Scholar
 [255]Johannsen, T., Psaltis, D. and McClintock, J.E., “Constraints on the Size of Extra Dimensions from the Orbital Evolution of BlackHole XRay Binaries”, Astrophys. J., 691, 997–1004, (2009). [DOI], [ADS], [arXiv:0803.1835 [astroph]]. (Cited on pages 23, 24, and 68.)ADSCrossRefGoogle Scholar
 [256]Kamaretsos, I., Hannam, M., Husa, S. and Sathyaprakash, B.S., “Blackhole hair loss: Learning about binary progenitors from ringdown signals”, Phys. Rev. D, 85, 024018, (2012). [DOI], [ADS], [arXiv:1107.0854 [grqc]]. (Cited on page 89.)ADSCrossRefGoogle Scholar
 [257]Kanti, P., Mavromatos, N.E., Rizos, J., Tamvakis, K. and Winstanley, E., “Dilatonic black holes in higher curvature string gravity”, Phys. Rev. D, 54, 5049–5058, (1996). [DOI], [arXiv:hepth/9511071]. (Cited on page 58.)ADSMathSciNetCrossRefGoogle Scholar
 [258]Kanti, P., Mavromatos, N.E., Rizos, J., Tamvakis, K. and Winstanley, E., “Dilatonic black holes in higher curvature string gravity: II. Linear stability”, Phys. Rev. D, 57, 6255–6264, (1998). [DOI], [arXiv:hepth/9703192]. (Cited on page 58.)ADSMathSciNetCrossRefGoogle Scholar
 [259]Kapner, D.J., Cook, T.S., Adelberger, E.G., Gundlach, J.H., Heckel, B.R., Hoyle, C.D. and Swanson, H.E., “Tests of the Gravitational InverseSquare Law below the DarkEnergy Length Scale”, Phys. Rev. Lett., 98, 021101, (2007). [DOI], [arXiv:hepph/0611184]. (Cited on pages 21, 23, and 25.)ADSCrossRefGoogle Scholar
 [260]Kaspi, V.M., Taylor, J.H. and Ryba, M.F., “Highprecision timing of millisecond pulsars. III. Longterm monitoring of PSRs B1855+09 and B1937+21”, Astrophys. J., 428, 713, (1994). [DOI], [ADS]. (Cited on page 25.)ADSCrossRefGoogle Scholar
 [261]Kehagias, A. and Sfetsos, K., “Deviations from the 1/r^{2} Newton law due to extra dimensions”, Phys. Lett. B, 472, 39–44, (2000). [DOI], [arXiv:hepph/9905417]. (Cited on page 63.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [262]Keppel, D. and Ajith, P., “Constraining the mass of the graviton using coalescing blackhole binaries”, Phys. Rev. D, 82, 122001, (2010). [DOI], [ADS], [arXiv: 1004.0284 [grqc]]. (Cited on pages 56, 64, and 65.)ADSCrossRefGoogle Scholar
 [263]Kesden, M., Gair, J.R. and Kamionkowski, M., “Gravitationalwave signature of an inspiral into a supermassive horizonless object”, Phys. Rev. D, 71, 044015, (2005). [DOI], [ADS], [arXiv:astroph/0411478]. (Cited on pages 8 and 91.)ADSCrossRefGoogle Scholar
 [264]Kim, H., “New black hole solutions in BransDicke theory of gravity”, Phys. Rev. D, 60, 024001, (1999). [DOI], [arXiv:grqc/9811012]. (Cited on page 52.)ADSMathSciNetCrossRefGoogle Scholar
 [265]Klein, A., Cornish, N. and Yunes, N., “Gravitational Waveforms for Precessing, Quasicircular Binaries via Multiple Scale Analysis and Uniform Asymptotics: The Near Spin Alignment Case”, arXiv, eprint, (2013). [arXiv:1305.1932 [grqc]]. (Cited on page 55.)Google Scholar
 [266]Kocsis, B., Haiman, Z. and Menou, K., “Premerger Localization of Gravitational Wave Standard Sirens with LISA: Triggered Search for an Electromagnetic Counterpart”, Astrophys. J., 684, 870–887, (2008). [DOI], [ADS], [arXiv:0712.1144 [astroph]]. (Cited on page 19.)ADSCrossRefGoogle Scholar
 [267]Kocsis, B., Yunes, N. and Loeb, A., “Observable signatures of extreme massratio inspiral black hole binaries embedded in thin accretion disks”, Phys. Rev. D, 84, 024032, (2011). [DOI], [ADS], [arXiv:1104.2322 [astroph.GA]]. (Cited on pages 69, 80, and 86.)ADSCrossRefGoogle Scholar
 [268]Kodama, H. and Yoshino, H., “Axiverse and Black Hole”, Int. J. Mod. Phys.: Conf. Ser., 7, 84–115, (2012). [DOI], [ADS], [arXiv:1108.1365 [hepth]]. (Cited on page 22.)Google Scholar
 [269]Kogan, I. I., Mouslopoulos, S. and Papazoglou, A., “The m → 0 limit for massive graviton in dS_{4} and AdS_{4}: How to circumvent the van DamVeltmanZakharov discontinuity”, Phys. Lett. B, 503, 173–180, (2001). [DOI], [arXiv:hepth/0011138]. (Cited on page 18.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [270]Kolmogorov, A.N., “O sohranenii uslovnoperiodicheskih dvizhenij pri malom izmenenii funkcii Gamil’tona”, Dokl. Akad. Nauk. SSSR, 98, 527–530, (1954). On conservation of conditionally periodic motions for a small change in Hamilton’s function. (Cited on page 85.)MathSciNetGoogle Scholar
 [271]Komatsu, E. et al. (WMAP Collaboration), “Fiveyear Wilkinson Microwave Anisotropy Probe observations: cosmological interpretation”, Astrophys. J. Suppl. Ser., 180, 330–376, (2009). [DOI], [ADS], [arXiv:0803.0547 [astroph]]. (Cited on page 69.)ADSCrossRefGoogle Scholar
 [272]Konno, K., Matsuyama, T. and Tanda, S., “Rotating Black Hole in Extended ChernSimons Modified Gravity”, Prog. Theor. Phys., 122, 561–568, (2009). [DOI], [ADS], [arXiv:0902.4767 [grqc]]. (Cited on pages 23, 57, and 58.)ADSzbMATHCrossRefGoogle Scholar
 [273]Kramer, M. and Wex, N., “The double pulsar system: A unique laboratory for gravity”, Class. Quantum Grav., 26, 073001, (2009). [DOI], [ADS]. (Cited on page 82.)ADSzbMATHCrossRefGoogle Scholar
 [274]Kramer, M. et al., “Tests of General Relativity from Timing the Double Pulsar”, Science, 314, 97–102, (2006). [DOI], [ADS], [arXiv:astroph/0609417]. (Cited on pages 5 and 82.)ADSCrossRefGoogle Scholar
 [275]Kusenko, A., “Solitons in the supersymmetric extensions of the standard model”, Phys. Lett. B, 405, 108–113, (1997). [DOI], [arXiv:hepph/9704273]. (Cited on page 90.)ADSMathSciNetCrossRefGoogle Scholar
 [276]Kusenko, A., “Supersymmetric Qballs: Theory and Cosmology”, in Nath, P., ed., Particles, Strings And Cosmology (PASCOS 98), Proceedings of the Sixth International Symposium, Boston, Massachusetts, 22–29 March 1998, pp. 540–543, (World Scientific, Singapore; Hackensack, NJ, 1999). [arXiv:hepph/9806529]. (Cited on page 90.)Google Scholar
 [277]Laguna, P., Larson, S.L., Spergel, D. and Yunes, N., “Integrated SachsWolfe Effect for Gravitational Radiation”, Astrophys. J. Lett., 715, L12–L15 (2009). [DOI], [arXiv:0905.1908 [grqc]]. (Cited on page 70.)ADSCrossRefGoogle Scholar
 [278]LanahanTremblay, N. and Faraoni, V., “The Cauchy problem of f(R) gravity”, Class. Quantum Grav., 24, 5667–5679, (2007). [DOI], [ADS], [arXiv:0709.4414 [grqc]]. (Cited on page 16.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [279]Lang, R.N. and Hughes, S.A., “Measuring coalescing massive binary black holes with gravitational waves: The impact of spininduced precession”, Phys. Rev. D, 74, 122001, (2006). [DOI], [arXiv:grqc/0608062]. Errata: 10.1103/PhysRevD.75.089902, 10.1103/PhysRevD.77.109901. (Cited on page 56.)ADSCrossRefGoogle Scholar
 [280]Lang, R.N., Hughes, S.A. and Cornish, N.J., “Measuring parameters of massive black hole binaries with partially aligned spins”, Phys. Rev. D, 84, 022002, (2011). [DOI], [arXiv:1101.3591 [grqc]]. (Cited on page 56.)ADSCrossRefGoogle Scholar
 [281]Larson, S.L. and Hiscock, W.A., “Using binary stars to bound the mass of the graviton”, Phys. Rev. D, 61, 104008, (2000). [DOI], [ADS], [arXiv:grqc/9912102]. (Cited on pages 64 and 65.)ADSCrossRefGoogle Scholar
 [282]Lattimer, J.M. and Schutz, B.F., “Constraining the equation of state with moment of inertia measurements”, Astrophys. J., 629, 979–984, (2005). [DOI], [arXiv:astroph/0411470]. (Cited on page 82.)ADSCrossRefGoogle Scholar
 [283]Lattimer, J.M. and Swesty, F.D., “A generalized equation of state for hot, dense matter”, Nucl. Phys. A, 535, 331–376, (1991). [DOI], [ADS]. (Cited on page 81.)ADSCrossRefGoogle Scholar
 [284]Lee, K., Jenet, F.A. and Price, R.H., “Pulsar Timing as a Probe of NonEinsteinian Polarizations of Gravitational Waves”, Astrophys. J., 685, 1304–1319, (2008). [DOI], [ADS]. (Cited on pages 37, 45, 48, and 49.)ADSCrossRefGoogle Scholar
 [285]Lee, K., Jenet, F.A., Price, R.H., Wex, N. and Kramer, M., “Detecting Massive Gravitons Using Pulsar Timing Arrays”, Astrophys. J., 722, 1589–1597 (2010). [DOI], [ADS], [arXiv: 1008.2561 [astroph.HE]]. (Cited on page 49.)ADSCrossRefGoogle Scholar
 [286]Letelier, P.S. and Vieira, W.M., “Chaos and rotating black holes with halos”, Phys. Rev. D, 56, 8095–8098, (1997). [DOI], [arXiv:grqc/9712008]. (Cited on page 85.)ADSMathSciNetCrossRefGoogle Scholar
 [287]Letelier, P.S. and Vieira, W.M., “Chaos in black holes surrounded by gravitational waves”, Class. Quantum Grav., 14, 1249–1257, (1997). [DOI], [ADS], [arXiv:grqc/9706025]. (Cited on page 85.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [288]Letelier, P.S. and Vieira, W.M., “Chaos and TaubNUT related spacetimes”, Phys. Lett. A, 244, 324–328, (1998). [DOI], [arXiv:grqc/9712030]. (Cited on page 85.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [289]Li, C. and Lovelace, G., “Generalization of Ryan’s theorem: Probing tidal coupling with gravitational waves from nearly circular, nearly equatorial, extrememassratio inspirals”, Phys. Rev. D, 77, 064022, (2008). [DOI], [ADS], [arXiv:grqc/0702146]. (Cited on pages 8 and 84.)ADSCrossRefGoogle Scholar
 [290]Li, T.G.F. et al., “Towards a generic test of the strong field dynamics of general relativity using compact binary coalescence”, Phys. Rev. D, 85, 082003, (2012). [DOI], [ADS], [arXiv:1110.0530 [grqc]]. (Cited on pages 42, 74, and 79.)ADSCrossRefGoogle Scholar
 [291]Li, T.G.F. et al., “Towards a generic test of the strong field dynamics of general relativity using compact binary coalescence: Further investigations”, J. Phys.: Conf. Ser., 363, 012028, (2012). [DOI], [ADS], [arXiv:1 111.5274 [grqc]]. (Cited on page 42.)ADSGoogle Scholar
 [292]Lichtenberg, A.J. and Lieberman, M.A., Regular and Chaotic Dynamics, Applied Mathematical Sciences, 38, (Springer, Berlin, 1992), 2nd edition. (Cited on page 85.)zbMATHCrossRefGoogle Scholar
 [293]Lightman, A.P. and Lee, D.L., “New TwoMetric Theory of Gravity with Prior Geometry”, Phys. Rev. D, 8, 3293–3302, (1973). [DOI], [ADS]. (Cited on page 24.)ADSCrossRefGoogle Scholar
 [294]Littenberg, T.B. and Cornish, N.J., “Bayesian approach to the detection problem in gravitational wave astronomy”, Phys. Rev. D, 80, 063007, (2009). [DOI], [ADS], [arXiv:0902.0368 [grqc]]. (Cited on pages 39 and 41.)ADSCrossRefGoogle Scholar
 [295]Lue, A., Wang, L. and Kamionkowski, M., “Cosmological Signature of New ParityViolating Interactions”, Phys. Rev. Lett., 83, 1506–1509, (1999). [DOI], [arXiv:astroph/9812088]. (Cited on page 69.)ADSCrossRefGoogle Scholar
 [296]LukesGerakopoulos, G., “The nonintegrability of the ZipoyVoorhees metric”, Phys. Rev. D, 86, 044013, (2012). [DOI], [arXiv:1206.0660 [grqc]]. (Cited on page 85.)ADSCrossRefGoogle Scholar
 [297]LukesGerakopoulos, G., Apostolatos, T.A. and Contopoulos, G., “Observable signature of a background deviating from the Kerr metric”, Phys. Rev. D, 81, 124005, (2010). [DOI], [ADS], [arXiv:1003.3120 [grqc]]. (Cited on pages 8, 85, and 86.)ADSCrossRefGoogle Scholar
 [298]Lyne, A.G. et al., “A DoublePulsar System: A Rare Laboratory for Relativistic Gravity and Plasma Physics”, Science, 303, 1153–1157, (2004). [DOI], [arXiv:astroph/0401086]. (Cited on page 5.)ADSCrossRefGoogle Scholar
 [299]Maggiore, M. and Nicolis, A., “Detection strategies for scalar gravitational waves with interferometers and resonant spheres”, Phys. Rev. D, 62, 024004, (2000). [DOI], [ADS], [arXiv:grqc/9907055]. (Cited on page 45.)ADSCrossRefGoogle Scholar
 [300]Magueijo, J. and Smolin, L., “Lorentz Invariance with an Invariant Energy Scale”, Phys. Rev. Lett., 88, 190403, (2002). [DOI], [ADS], [arXiv:hepth/0112090]. (Cited on page 19.)ADSCrossRefGoogle Scholar
 [301]Maldacena, J.M., “The large N limit of superconformal field theories and supergravity”, Adv. Theor. Math. Phys., 2, 231–252, (1998). [ADS], [arXiv:hepth/9711200]. (Cited on page 25.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [302]Manko, V.S. and Novikov, I.D., “Generalizations of the Kerr and KerrNewman metrics possessing an arbitrary set of massmultipole moments”, Class. Quantum Grav., 9, 2477–2487, (1992). [DOI], [ADS]. (Cited on page 85.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [303]Marsh, D.J.E., Macaulay, E., Trebitsch, M. and Ferreira, P.G., “Ultralight axions: Degeneracies with massive neutrinos and forecasts for future cosmological observations”, Phys. Rev. D, 85, 103514, (2012). [DOI], [ADS], [arXiv:1110.0502 [astroph.CO]]. (Cited on page 22.)ADSCrossRefGoogle Scholar
 [304]Maselli, A., Cardoso, V., Ferrari, V., Gualtieri, L. and Pani, P., “Equationofstateindependent relations in neutron stars”, Phys. Rev. D, 88, 023007 (2013). [DOI], [ADS], [arXiv:1304.2052 [grqc]]. (Cited on page 82.)ADSCrossRefGoogle Scholar
 [305]Maselli, A., Gualtieri, L., Pannarale, F. and Ferrari, V., “On the validity of the adiabatic approximation in compact binary inspirals”, Phys. Rev. D, 86, 044032, (2012). [DOI], [arXiv:1205.7006 [grqc]]. (Cited on page 82.)ADSCrossRefGoogle Scholar
 [306]Mazur, P.O., “Proof of uniqueness of the KerrNewman black hole solution”, J. Phys. A: Math. Gen., 15, 3173–3180, (1982). [DOI], [ADS]. (Cited on page 83.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [307]Mazur, P.O. and Mottola, E., “Gravitational Condensate Stars: An Alternative to Black Holes”, arXiv, eprint, (2001). [ADS], [arXiv:grqc/0109035]. (Cited on page 91.)Google Scholar
 [308]McWilliams, S.T., “Constraining the Braneworld with Gravitational Wave Observations”, Phys. Rev. Lett., 104, 141601, (2010). [DOI], [ADS], [arXiv:0912.4744 [grqc]]. (Cited on page 69.)ADSCrossRefGoogle Scholar
 [309]Medved, A.J.M., Martin, D. and Visser, M., “Dirty black holes: quasinormal modes”, Class. Quantum Grav., 21, 1393–1406, (2004). [DOI], [arXiv:grqc/0310009]. (Cited on page 91.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [310]Medved, A.J.M., Martin, D. and Visser, M., “Dirty black holes: Quasinormal modes for ‘squeezed’ horizons”, Class. Quantum Grav., 21, 2393–2405, (2004). [DOI], [arXiv:grqc/0310097]. (Cited on page 91.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [311]Mercuri, S. and Taveras, V., “Interaction of the BarberoImmirzi field with matter and pseudoscalar perturbations”, Phys. Rev. D, 80, 104007, (2009). [DOI], [arXiv:0903.4407 [grqc]]. (Cited on pages 20 and 28.)ADSMathSciNetCrossRefGoogle Scholar
 [312]Merritt, D., Alexander, T., Mikkola, S. and Will, C.M., “Testing properties of the galactic center black hole using stellar orbits”, Phys. Rev. D, 81, 062002, (2010). [DOI], [arXiv:0911.4718 [astroph.GA]]. (Cited on page 84.)ADSCrossRefGoogle Scholar
 [313]Merritt, D., Alexander, T., Mikkola, S. and Will, C.M., “Stellar dynamics of extrememassratio inspirals”, Phys. Rev. D, 84, 044024, (2011). [DOI], [ADS], [arXiv:1102.3180 [astroph.CO]]. (Cited on page 84.)ADSCrossRefGoogle Scholar
 [314]Metropolis, N., “Summation of imprecise numbers”, Comput. Math. Appl., 6, 297–299, (1980). [DOI]. (Cited on page 41.)MathSciNetzbMATHCrossRefGoogle Scholar
 [315]Mirshekari, S. and Will, C.M., “Compact binary systems in scalartensor gravity: Equations of motion to 2.5 postNewtonian order”, Phys. Rev. D, 87, 084070, (2013). [DOI], [ADS], [arXiv:1301.4680 [grqc]]. (Cited on pages 16 and 52.)ADSCrossRefGoogle Scholar
 [316]Mirshekari, S., Yunes, N. and Will, C.M., “Constraining Generic Lorentz Violation and the Speed of the Graviton with Gravitational Waves”, Phys. Rev. D, 85, 024041, (2012). [DOI], [arXiv:1110.2720 [grqc]]. (Cited on pages 8, 18, and 66.)ADSCrossRefGoogle Scholar
 [317]Mishra, C.K., Arun, K.G., Iyer, B.R. and Sathyaprakash, B.S., “Parametrized tests of postNewtonian theory using Advanced LIGO and Einstein Telescope”, Phys. Rev. D, 82, 064010 (2010). [DOI], [ADS], [arXiv:1005.0304 [grqc]]. (Cited on page 74.)ADSCrossRefGoogle Scholar
 [318]Misner, C.W., Thorne, K.S. and Wheeler, J.A., Gravitation, (W.H. Freeman, San Francisco, 1973). [ADS]. (Cited on pages 9 and 83.)Google Scholar
 [319]Molina, C., Pani, P., Cardoso, V. and Gualtieri, L., “Gravitational signature of Schwarzschild black holes in dynamical ChernSimons gravity”, Phys. Rev. D, 81, 124021, (2010). [DOI], [ADS], [arXiv:1004.4007 [grqc]]. (Cited on page 23.)ADSCrossRefGoogle Scholar
 [320]Mora, T. and Will, C.M., “PostNewtonian diagnostic of quasiequilibrium binary configurations of compact objects”, Phys. Rev. D, 69, 104021, (2004). [DOI], [arXiv:grqc/0312082 [grqc]]. (Cited on page 81.)ADSCrossRefGoogle Scholar
 [321]Moser, J., “On Invariant Curves of AreaPreserving Mappings of an Annulus”, Nachr. Akad. Wiss. Goettingen II, Math.Phys. Kl., 1962, 1–20, (1962). (Cited on page 85.)MathSciNetzbMATHGoogle Scholar
 [322]Moyal, J.E. and Bartlett, M.S., “Quantum mechanics as a statistical theory”, Proc. Cambridge Philos. Soc., 45, 99–124 (1949). [DOI], [ADS]. (Cited on page 26.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [323]Nakao, K.I., Harada, T., Shibata, M., Kawamura, S. and Nakamura, T., “Response of interferometric detectors to scalar gravitational waves”, Phys. Rev. D, 63, 082001, (2001). [DOI], [ADS], [arXiv:grqc/0006079]. (Cited on page 45.)ADSCrossRefGoogle Scholar
 [324]Nelson, W., “Static solutions for fourth order gravity”, Phys. Rev. D, 82, 104026, (2010). [DOI], [ADS], [arXiv:1010.3986 [grqc]]. (Cited on page 28.)ADSCrossRefGoogle Scholar
 [325]Nelson, W., Ochoa, J. and Sakellariadou, M., “Constraining the Noncommutative Spectral Action via Astrophysical Observations”, Phys. Rev. Lett., 105, 101602, (2010). [DOI], [arXiv: 1005.4279 [hepth]]. (Cited on pages 27, 28, 61, 62, and 63.)ADSCrossRefGoogle Scholar
 [326]Nelson, W., Ochoa, J. and Sakellariadou, M., “Gravitational Waves in the Spectral Action of Noncommutative Geometry”, Phys. Rev. D, 82, 085021, (2010). [DOI], [arXiv:1005.4276 [hepth]]. (Cited on pages 27, 28, 61, 62, and 63.)ADSCrossRefGoogle Scholar
 [327]Newman, E.T. and Janis, A.I., “Note on the Kerr SpinningParticle Metric”, J. Math. Phys., 6, 915–917, (1965). [DOI], [ADS]. (Cited on page 87.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [328]Ni, W.T., “Solarsystem tests of the inflation model with a Weyl term”, arXiv, eprint, (2012). [ADS], [arXiv:1203.2465 [astroph.CO]]. (Cited on page 28.)Google Scholar
 [329]Nishizawa, A., Taruya, A., Hayama, K., Kawamura, S. and Sakagami, M.A., “Probing nontensorial polarizations of stochastic gravitationalwave backgrounds with groundbased laser interferometers”, Phys. Rev. D, 79, 082002, (2009). [DOI], [ADS], [arXiv:0903.0528 [astroph.CO]]. (Cited on pages 32, 45, 46, 48, 49, and 81.)ADSCrossRefGoogle Scholar
 [330]Nishizawa, A., Taruya, A. and Kawamura, S., “Cosmological test of gravity with polarizations of stochastic gravitational waves around 0.1–1 Hz”, Phys. Rev. D, 81, 104043, (2010). [DOI], [ADS], [arXiv:0911.0525 [grqc]]. (Cited on page 81.)ADSCrossRefGoogle Scholar
 [331]Nordtvedt Jr, K.L., “Equivalence Principle for Massive Bodies: II. Theory”, Phys. Rev., 169, 1017–1025, (1968). [DOI], [ADS]. (Cited on page 73.)ADSzbMATHCrossRefGoogle Scholar
 [332]Nordtvedt Jr, K.L. and Will, C.M., “Conservation Laws and Preferred Frames in Relativistic Gravity. II. Experimental Evidence to Rule Out PreferredFrame Theories of Gravity”, Astrophys. J., 177, 775–792, (1972). [DOI], [ADS]. (Cited on page 73.)ADSMathSciNetCrossRefGoogle Scholar
 [333]Novak, J., “Spherical neutron star collapse toward a black hole in a tensorscalar theory of gravity”, Phys. Rev. D, 57, 4789–4801, (1998). [DOI], [ADS], [arXiv:grqc/9707041]. (Cited on page 43.)ADSMathSciNetCrossRefGoogle Scholar
 [334]Novak, J. and Ibáñez, J.M., “Gravitational waves from the collapse and bounce of a stellar core in tensor scalar gravity”, Astrophys. J., 533, 392–405, (2000). [DOI], [ADS], [arXiv:astroph/9911298]. (Cited on page 43.)ADSCrossRefGoogle Scholar
 [335]O’Connor, E. and Ott, C.D., “A new opensource code for spherically symmetric stellar collapse to neutron stars and black holes”, Class. Quantum Grav., 27, 114103, (2010). [DOI], [arXiv:0912.2393 [astroph.HE]]. (Cited on page 81.)ADSMathSciNetzbMATHCrossRefGoogle Scholar
 [336]Ohashi, A., Tagoshi, H. and Sasaki, M., “PostNewtonian Expansion of Gravitational Waves from a Compact Star Orbiting a Rotating Black Hole in BransDicke Theory: Circular Orbit Case”, Prog. Theor. Phys., 96, 713–727, (1996). [DOI]. (Cited on page 16.)ADSCrossRefGoogle Scholar
 [337]Ostrogradski, M.V., “Mémoire sur les équations différentielles relatives au problème des isopérimètres”, Mem. Acad. St. Petersbourg, VI Ser., 4, 385–517, (1850). (Cited on page 13.)Google Scholar
 [338]Palenzuela, C., Lehner, L. and Liebling, S.L., “Orbital dynamics of binary boson star systems”, Phys. Rev. D, 77, 044036, (2008). [DOI], [arXiv:0706.2435 [grqc]]. (Cited on page 91.)ADSCrossRefGoogle Scholar
 [339]Palenzuela, C., Olabarrieta, I., Lehner, L. and Liebling, S.L., “Headon collisions of boson stars”, Phys. Rev. D, 75, 064005, (2007). [DOI], [arXiv:grqc/0612067]. (Cited on page 91.)ADSCrossRefGoogle Scholar
 [340]