Abstract
The status of experimental tests of general relativity and of theoretical frameworks for analyzing them is reviewed. Einstein’s equivalence principle (EEP) is well supported by experiments such as the Eötvös experiment, tests of special relativity, and the gravitational redshift experiment. Ongoing tests of EEP and of the inverse square law are searching for new interactions arising from unification or quantum gravity. Tests of general relativity at the postNewtonian level have reached high precision, including the light deflection, the Shapiro time delay, the perihelion advance of Mercury, and the Nordtvedt effect in lunar motion. Gravitational wave damping has been detected in an amount that agrees with general relativity to better than half a percent using the HulseTaylor binary pulsar, and other binary pulsar systems have yielded other tests, especially of strongfield effects. When direct observation of gravitational radiation from astrophysical sources begins, new tests of general relativity will be possible.
Introduction
At the time of the birth of general relativity (GR), experimental confirmation was almost a side issue. Einstein did calculate observable effects of general relativity, such as the perihelion advance of Mercury, which he knew to be an unsolved problem, and the deflection of light, which was subsequently verified. But compared to the inner consistency and elegance of the theory, he regarded such empirical questions as almost peripheral. Today, experimental gravitation is a major component of the field, characterized by continuing efforts to test the theory’s predictions, to search for gravitational imprints of highenergy particle interactions, and to detect gravitational waves from astronomical sources.
The modern history of experimental relativity can be divided roughly into four periods: Genesis, Hibernation, a Golden Era, and the Quest for Strong Gravity. The Genesis (1887–1919) comprises the period of the two great experiments which were the foundation of relativistic physics — the MichelsonMorley experiment and the Eötvös experiment — and the two immediate confirmations of GR — the deflection of light and the perihelion advance of Mercury. Following this was a period of Hibernation (1920–1960) during which theoretical work temporarily outstripped technology and experimental possibilities, and, as a consequence, the field stagnated and was relegated to the backwaters of physics and astronomy.
But beginning around 1960, astronomical discoveries (quasars, pulsars, cosmic background radiation) and new experiments pushed GR to the forefront. Experimental gravitation experienced a Golden Era (1960–1980) during which a systematic, worldwide effort took place to understand the observable predictions of GR, to compare and contrast them with the predictions of alternative theories of gravity, and to perform new experiments to test them. The period began with an experiment to confirm the gravitational frequency shift of light (1960) and ended with the reported decrease in the orbital period of the HulseTaylor binary pulsar at a rate consistent with the GR prediction of gravity wave energy loss (1979). The results all supported GR, and most alternative theories of gravity fell by the wayside (for a popular review, see [282]).
Since 1980, the field has entered what might be termed a Quest for Strong Gravity. Many of the remaining interesting weakfield predictions of the theory are extremely small and difficult to check, in some cases requiring further technological development to bring them into detectable range. The sense of a systematic assault on the weakfield predictions of GR has been supplanted to some extent by an opportunistic approach in which novel and unexpected (and sometimes inexpensive) tests of gravity have arisen from new theoretical ideas or experimental techniques, often from unlikely sources. Examples include the use of lasercooled atom and ion traps to perform ultraprecise tests of special relativity; the proposal of a “fifth” force, which led to a host of new tests of the weak equivalence principle; and recent ideas of large extra dimensions, which have motived new tests of the inverse square law of gravity at submillimeter scales.
Instead, much of the focus has shifted to experiments which can probe the effects of strong gravitational fields. The principal figure of merit that distinguishes strong from weak gravity is the quantity ϵ ∼ GM / (Rc^{2}), where G is the Newtonian gravitational constant, M is the characteristic mass scale of the phenomenon, R is the characteristic distance scale, and c is the speed of light. Near the event horizon of a nonrotating black hole, or for the expanding observable universe, ϵ ∼ 0.5; for neutron stars, ϵ ∼ 0.2. These are the regimes of strong gravity. For the solar system, ϵ < 10^{−5}; this is the regime of weak gravity. At one extreme are the strong gravitational fields associated with Planckscale physics. Will unification of the forces, or quantization of gravity at this scale leave observable effects accessible by experiment? Dramatically improved tests of the equivalence principle, of the inverse square law, or of local Lorentz invariance are being mounted, to search for or bound the imprinted effects of Planckscale phenomena. At the other extreme are the strong fields associated with compact objects such as black holes or neutron stars. Astrophysical observations and gravitational wave detectors are being planned to explore and test GR in the strongfield, highlydynamical regime associated with the formation and dynamics of these objects.
In this Living Review, we shall survey the theoretical frameworks for studying experimental gravitation, summarize the current status of experiments, and attempt to chart the future of the subject. We shall not provide complete references to early work done in this field but instead will refer the reader to the appropriate review articles and monographs, specifically to Theory and Experiment in Gravitational Physics [281], hereafter referred to as TEGP. Additional recent reviews in this subject are [276, 284, 286, 71, 98, 239]. References to TEGP will be by chapter or section, e.g., “TEGP 8.9 [281]”.
Tests of the Foundations of Gravitation Theory
The Einstein equivalence principle
The principle of equivalence has historically played an important role in the development of gravitation theory. Newton regarded this principle as such a cornerstone of mechanics that he devoted the opening paragraph of the Principia to it. In 1907, Einstein used the principle as a basic element in his development of general relativity. We now regard the principle of equivalence as the foundation, not of Newtonian gravity or of GR, but of the broader idea that spacetime is curved. Much of this viewpoint can be traced back to Robert Dicke, who contributed crucial ideas about the foundations of gravitation theory between 1960 and 1965. These ideas were summarized in his influential Les Houches lectures of 1964 [93], and resulted in what has come to be called the Einstein equivalence principle (EEP).
One elementary equivalence principle is the kind Newton had in mind when he stated that the property of a body called “mass” is proportional to the “weight”, and is known as the weak equivalence principle (WEP). An alternative statement of WEP is that the trajectory of a freely falling “test” body (one not acted upon by such forces as electromagnetism and too small to be affected by tidal gravitational forces) is independent of its internal structure and composition. In the simplest case of dropping two different bodies in a gravitational field, WEP states that the bodies fall with the same acceleration (this is often termed the Universality of Free Fall, or UFF).
The Einstein equivalence principle (EEP) is a more powerful and farreaching concept; it states that:

1.
WEP is valid.

2.
The outcome of any local nongravitational experiment is independent of the velocity of the freelyfalling reference frame in which it is performed.

3.
The outcome of any local nongravitational experiment is independent of where and when in the universe it is performed.
The second piece of EEP is called local Lorentz invariance (LLI), and the third piece is called local position invariance (LPI).
For example, a measurement of the electric force between two charged bodies is a local nongravitational experiment; a measurement of the gravitational force between two bodies (Cavendish experiment) is not.
The Einstein equivalence principle is the heart and soul of gravitational theory, for it is possible to argue convincingly that if EEP is valid, then gravitation must be a “curved spacetime” phenomenon, in other words, the effects of gravity must be equivalent to the effects of living in a curved spacetime. As a consequence of this argument, the only theories of gravity that can fully embody EEP are those that satisfy the postulates of “metric theories of gravity”, which are:

1.
Spacetime is endowed with a symmetric metric.

2.
The trajectories of freely falling test bodies are geodesics of that metric.

3.
In local freely falling reference frames, the nongravitational laws of physics are those written in the language of special relativity.
The argument that leads to this conclusion simply notes that, if EEP is valid, then in local freely falling frames, the laws governing experiments must be independent of the velocity of the frame (local Lorentz invariance), with constant values for the various atomic constants (in order to be independent of location). The only laws we know of that fulfill this are those that are compatible with special relativity, such as Maxwell’s equations of electromagnetism. Furthermore, in local freely falling frames, test bodies appear to be unaccelerated, in other words they move on straight lines; but such “locally straight” lines simply correspond to “geodesics” in a curved spacetime (TEGP 2.3 [281]).
General relativity is a metric theory of gravity, but then so are many others, including the BransDicke theory and its generalizations. Theories in which varying nongravitational constants are associated with dynamical fields that couple to matter directly are not metric theories. Neither, in this narrow sense, is superstring theory (see Section 2.3), which, while based fundamentally on a spacetime metric, introduces additional fields (dilatons, moduli) that can couple to material stressenergy in a way that can lead to violations, say, of WEP. It is important to point out, however, that there is some ambiguity in whether one treats such fields as EEPviolating gravitational fields, or simply as additional matter fields, like those that carry electromagnetism or the weak interactions. Still, the notion of curved spacetime is a very general and fundamental one, and therefore it is important to test the various aspects of the Einstein equivalence principle thoroughly. We first survey the experimental tests, and describe some of the theoretical formalisms that have been developed to interpret them. For other reviews of EEP and its experimental and theoretical significance, see [126, 162].
Tests of the weak equivalence principle
A direct test of WEP is the comparison of the acceleration of two laboratorysized bodies of different composition in an external gravitational field. If the principle were violated, then the accelerations of different bodies would differ. The simplest way to quantify such possible violations of WEP in a form suitable for comparison with experiment is to suppose that for a body with inertial mass m_{I}, the passive gravitational mass m_{P} is no longer equal to m_{I}, so that in a gravitational field g, the acceleration is given by m_{I}a = m_{P}g. Now the inertial mass of a typical laboratory body is made up of several types of massenergy: rest energy, electromagnetic energy, weakinteraction energy, and so on. If one of these forms of energy contributes to m_{P} differently than it does to m_{I}, a violation of WEP would result. One could then write
where E^{A} is the internal energy of the body generated by interaction A, η^{A} is a dimensionless parameter that measures the strength of the violation of WEP induced by that interaction, and c is the speed of light. A measurement or limit on the fractional difference in acceleration between two bodies then yields a quantity called the “Eötvös ratio” given by
where we drop the subscript “I” from the inertial masses. Thus, experimental limits on η place limits on the WEPviolation parameters η^{A}.
Many highprecision Eötvöstype experiments have been performed, from the pendulum experiments of Newton, Bessel, and Potter to the classic torsionbalance measurements of Eötvös [100], Dicke [94], Braginsky [43], and their collaborators. In the modern torsionbalance experiments, two objects of different composition are connected by a rod or placed on a tray and suspended in a horizontal orientation by a fine wire. If the gravitational acceleration of the bodies differs, and this difference has a component perpendicular to the suspension wire, there will be a torque induced on the wire, related to the angle between the wire and the direction of the gravitational acceleration g. If the entire apparatus is rotated about some direction with angular velocity ω, the torque will be modulated with period 2π/ω. In the experiments of Eötvös and his collaborators, the wire and g were not quite parallel because of the centripetal acceleration on the apparatus due to the Earth’s rotation; the apparatus was rotated about the direction of the wire. In the Dicke and Braginsky experiments, g was that of the Sun, and the rotation of the Earth provided the modulation of the torque at a period of 24 hr (TEGP 2.4 (a) [281]). Beginning in the late 1980s, numerous experiments were carried out primarily to search for a “fifth force” (see Section 2.3.1), but their null results also constituted tests of WEP. In the “freefall Galileo experiment” performed at the University of Colorado, the relative freefall acceleration of two bodies made of uranium and copper was measured using a laser interferometric technique. The “EötWash” experiments carried out at the University of Washington used a sophisticated torsion balance tray to compare the accelerations of various materials toward local topography on Earth, movable laboratory masses, the Sun and the galaxy [249, 19], and have reached levels of 3 × 10^{−13} [2]. The resulting upper limits on η are summarized in Figure 1 (TEGP 14.1 [281]; for a bibliography of experiments up to 1991, see [107]).
A number of projects are in the development or planning stage to push the bounds on η even lower. The project MICROSCOPE (MICROSatellite à Trainée Compensée pour l’Observation du Principe d’Équivalence) is designed to test WEP to 10^{−15}. It is being developed by the French space agency CNES for a possible launch in March, 2008, for a oneyear mission [59]. The dragcompensated satellite will be in a Sunsynchronous polar orbit at 700 km altitude, with a payload consisting of two differential accelerometers, one with elements made of the same material (platinum), and another with elements made of different materials (platinum and titanium).
Another, known as Satellite Test of the Equivalence Principle (STEP) [247], is under consideration as a possible joint effort of NASA and the European Space Agency (ESA), with the goal of a 10^{−18} test. STEP would improve upon MICROSCOPE by using cryogenic techniques to reduce thermal noise, among other effects. At present, STEP (along with a number of variants, called MiniSTEP and QuickSTEP) has not been approved by any agency beyond the level of basic design studies or supporting research and development. An alternative concept for a space test of WEP is GalileoGalilei [261], which uses a rapidly rotating differential accelerometer as its basic element. Its goal is a bound on η at the 10^{−13} level on the ground and 10^{−17} in space.
Tests of local Lorentz invariance
Although special relativity itself never benefited from the kind of “crucial” experiments, such as the perihelion advance of Mercury and the deflection of light, that contributed so much to the initial acceptance of GR and to the fame of Einstein, the steady accumulation of experimental support, together with the successful merger of special relativity with quantum mechanics, led to its being accepted by mainstream physicists by the late 1920s, ultimately to become part of the standard toolkit of every working physicist. This accumulation included

the classic MichelsonMorley experiment and its descendents [186, 237, 141, 46],

the IvesStillwell, RossiHall, and other tests of timedilation [136, 229, 103],

tests of the independence of the speed of light of the velocity of the source, using both binary Xray stellar sources and highenergy pions [44, 5],
In addition to these direct experiments, there was the Dirac equation of quantum mechanics and its prediction of antiparticles and spin; later would come the stunningly successful relativistic theory of quantum electrodynamics.
In 2005, on the 100th anniversary of the introduction of special relativity, one might ask “what is there to test?”. Special relativity has been so thoroughly integrated into the fabric of modern physics that its validity is rarely challenged, except by cranks and crackpots. It is ironic then, that during the past several years, a vigorous theoretical and experimental effort has been launched, on an international scale, to find violations of special relativity. The motivation for this effort is not a desire to repudiate Einstein, but to look for evidence of new physics “beyond” Einstein, such as apparent violations of Lorentz invariance that might result from certain models of quantum gravity. Quantum gravity asserts that there is a fundamental length scale given by the Planck length, L_{Pl} = (ħG/c^{3})^{1/2} = 1.6 × 10^{−33} cm, but since length is not an invariant quantity (LorentzFitzGerald contraction), then there could be a violation of Lorentz invariance at some level in quantum gravity. In brane world scenarios, while physics may be locally Lorentz invariant in the higher dimensional world, the confinement of the interactions of normal physics to our fourdimensional “brane” could induce apparent Lorentz violating effects. And in models such as string theory, the presence of additional scalar, vector, and tensor longrange fields that couple to matter of the standard model could induce effective violations of Lorentz symmetry. These and other ideas have motivated a serious reconsideration of how to test Lorentz invariance with better precision and in new ways.
A simple and useful way of interpreting some of these modern experiments, called the c^{2}formalism, is to suppose that the electromagnetic interactions suffer a slight violation of Lorentz invariance, through a change in the speed of electromagnetic radiation c relative to the limiting speed of material test particles (c_{0}, made to take the value unity via a choice of units), in other words, c ≠ 1 (see Section 2.2.3). Such a violation necessarily selects a preferred universal rest frame, presumably that of the cosmic background radiation, through which we are moving at about 370 km s^{−1} [167]. Such a Lorentznoninvariant electromagnetic interaction would cause shifts in the energy levels of atoms and nuclei that depend on the orientation of the quantization axis of the state relative to our universal velocity vector, and on the quantum numbers of the state. The presence or absence of such energy shifts can be examined by measuring the energy of one such state relative to another state that is either unaffected or is affected differently by the supposed violation. One way is to look for a shifting of the energy levels of states that are ordinarily equally spaced, such as the Zeemansplit 2J + 1 ground states of a nucleus of total spin J in a magnetic field; another is to compare the levels of a complex nucleus with the atomic hyperfine levels of a hydrogen maser clock. The magnitude of these “clock anisotropies” would be proportional to δ ≡c^{−2} − 1.
The earliest clock anisotropy experiments were the HughesDrever experiments, performed in the period 1959–60 independently by Hughes and collaborators at Yale University, and by Drever at Glasgow University, although their original motivation was somewhat different [131, 96]. The HughesDrever experiments yielded extremely accurate results, quoted as limits on the parameter δ ≡ c^{−2} − 1 in Figure 2. Dramatic improvements were made in the 1980s using lasercooled trapped atoms and ions [215, 163, 53]. This technique made it possible to reduce the broading of resonance lines caused by collisions, leading to improved bounds on δ shown in Figure 2 (experiments labelled NIST, U. Washington and Harvard, respectively).
Also included for comparison is the corresponding limit obtained from MichelsonMorley type experiments (for a review, see [127]). In those experiments, when viewed from the preferred frame, the speed of light down the two arms of the moving interferometer is c, while it can be shown using the electrodynamics of the c^{2} formalism, that the compensating LorentzFitzGerald contraction of the parallel arm is governed by the speed c_{0} = 1. Thus the MichelsonMorley experiment and its descendants also measure the coefficient c^{−2} − 1. One of these is the BrilletHall experiment [46], which used a FabryPerot laser interferometer. In a recent series of experiments, the frequencies of electromagnetic cavity oscillators in various orientations were compared with each other or with atomic clocks as a function of the orientation of the laboratory [297, 168, 190, 12, 248]. These placed bounds on c^{−2} − 1 at the level of better than a part in 10^{9}. Haugan and Lömmerzahl [125] have considered the bounds that MichelsonMorley type experiments could place on a modified electrodynamics involving a “vectorvalued” effective photon mass.
The c^{2} framework focusses exclusively on classical electrodynamics. It has recently been extended to the entire standard model of particle physics by Kostelecký and colleagues [63, 64, 155]. The “Standard Model Extension” (SME) has a large number of Lorentzviolating parameters, opening up many new opportunities for experimental tests (see Section 2.2.4). A variety of clock anisotropy experiments have been carried out to bound the electromagnetic parameters of the SME framework [154]. For example, the cavity experiments described above [297, 168, 190] placed bounds on the coefficients of the tensors \({{\tilde k}_{{\rm{e}} }}\) and \({{\tilde k}_{{\rm{o +}}}}\) (see Section 2.2.4 for definitions) at the levels of 10^{−14} and 10^{−10}, respectively. Direct comparisons between atomic clocks based on different nuclear species place bounds on SME parameters in the neutron and proton sectors, depending on the nature of the transitions involved. The bounds achieved range from 10^{−27} to 10^{−32} GeV.
Astrophysical observations have also been used to bound Lorentz violations. For example, if photons satisfy the Lorentz violating dispersion relation
where E_{Pl} = (ħc^{5}/G)^{1/2} is the Planck energy, then the speed of light υ_{γ} = ∂E/∂p would be given, to linear order in the f^{(n)} by
Such a Lorentzviolating dispersion relation could be a relic of quantum gravity, for instance. By bounding the difference in arrival time of highenergy photons from a burst source at large distances, one could bound contributions to the dispersion for n > 2. One limit, f^{(3)} < 128 comes from observations of 1 and 2 TeV gamma rays from the blazar Markarian 421 [30]. Another limit comes from birefringence in photon propagation: In many Lorentz violating models, different photon polarizations may propagate with different speeds, causing the plane of polarization of a wave to rotate. If the frequency dependence of this rotation has a dispersion relation similar to Equation (3), then by studying “polarization diffusion” of light from a polarized source in a given bandwidth, one can effectively place a bound f^{(3)} < 10^{−4} [119]. Other testable effects of Lorentz invariance violation include threshold effects in particle reactions, gravitational Cerenkov radiation, and neutrino oscillations.
Mattingly [182] gives a thorough and uptodate review of both the theoretical frameworks and the experimental results for tests of LLI.
Tests of local position invariance
The principle of local position invariance, the third part of EEP, can be tested by the gravitational redshift experiment, the first experimental test of gravitation proposed by Einstein. Despite the fact that Einstein regarded this as a crucial test of GR, we now realize that it does not distinguish between GR and any other metric theory of gravity, but is only a test of EEP. A typical gravitational redshift experiment measures the frequency or wavelength shift Z ≡ Δν/ν = − Δλ/λ between two identical frequency standards (clocks) placed at rest at different heights in a static gravitational field. If the frequency of a given type of atomic clock is the same when measured in a local, momentarily comoving freely falling frame (Lorentz frame), independent of the location or velocity of that frame, then the comparison of frequencies of two clocks at rest at different locations boils down to a comparison of the velocities of two local Lorentz frames, one at rest with respect to one clock at the moment of emission of its signal, the other at rest with respect to the other clock at the moment of reception of the signal. The frequency shift is then a consequence of the firstorder Doppler shift between the frames. The structure of the clock plays no role whatsoever. The result is a shift
where ΔU is the difference in the Newtonian gravitational potential between the receiver and the emitter. If LPI is not valid, then it turns out that the shift can be written
where the parameter α may depend upon the nature of the clock whose shift is being measured (see TEGP 2.4 (c) [281] for details).
The first successful, highprecision redshift measurement was the series of PoundRebkaSnider experiments of 1960–1965 that measured the frequency shift of gammaray photons from ^{57}Fe as they ascended or descended the Jefferson Physical Laboratory tower at Harvard University. The high accuracy achieved — one percent — was obtained by making use of the Mossbauer effect to produce a narrow resonance line whose shift could be accurately determined. Other experiments since 1960 measured the shift of spectral lines in the Sun’s gravitational field and the change in rate of atomic clocks transported aloft on aircraft, rockets and satellites. Figure 3 summarizes the important redshift experiments that have been performed since 1960 (TEGP 2.4 (c) [281]).
After almost 50 years of inconclusive or contradictory measurements, the gravitational redshift of solar spectral lines was finally measured reliably. During the early years of GR, the failure to measure this effect in solar lines was siezed upon by some as reason to doubt the theory. Unfortunately, the measurement is not simple. Solar spectral lines are subject to the “limb effect”, a variation of spectral line wavelengths between the center of the solar disk and its edge or “limb”; this effect is actually a Doppler shift caused by complex convective and turbulent motions in the photosphere and lower chromosphere, and is expected to be minimized by observing at the solar limb, where the motions are predominantly transverse. The secret is to use strong, symmetrical lines, leading to unambiguous wavelength measurements. Successful measurements were finally made in 1962 and 1972 (TEGP 2.4 (c) [281]). In 1991, LoPresto et al. [172] measured the solar shift in agreement with LPI to about 2 percent by observing the oxygen triplet lines both in absorption in the limb and in emission just off the limb.
The most precise standard redshift test to date was the VessotLevine rocket experiment that took place in June 1976 [264]. A hydrogenmaser clock was flown on a rocket to an altitude of about 10,000 km and its frequency compared to a similar clock on the ground. The experiment took advantage of the masers’ frequency stability by monitoring the frequency shift as a function of altitude. A sophisticated data acquisition scheme accurately eliminated all effects of the firstorder Doppler shift due to the rocket’s motion, while tracking data were used to determine the payload’s location and the velocity (to evaluate the potential difference ΔU, and the special relativistic time dilation). Analysis of the data yielded a limit α < 2 × 10^{−4}.
A “null” redshift experiment performed in 1978 tested whether the relative rates of two different clocks depended upon position. Two hydrogen maser clocks and an ensemble of three superconductingcavity stabilized oscillator (SCSO) clocks were compared over a 10day period. During the period of the experiment, the solar potential U/c^{2} changed sinusoidally with a 24hour period by 3 × 10^{−13} because of the Earth’s rotation, and changed linearly at 3 × 10^{−12} per day because the Earth is 90 degrees from perihelion in April. However, analysis of the data revealed no variations of either type within experimental errors, leading to a limit on the LPI violation parameter α^{H} − α^{SCSO} < 2 × 10^{−2} [258]. This bound has been improved using more stable frequency standards, such as atomic fountain clocks [120, 216, 23]. The current bound, from comparing a Cesium atomic fountain with a Hydrogen maser for a year, is α^{H} − α^{Cs} < 2.1 × 10^{−5} [23].
The varying gravitational redshift of Earthbound clocks relative to the highly stable millisecond pulsar PSR 1937+21, caused by the Earth’s motion in the solar gravitational field around the EarthMoon center of mass (amplitude 4000 km), was measured to about 10 percent [251]. Two measurements of the redshift using stable oscillator clocks on spacecraft were made at the one percent level: One used the Voyager spacecraft in Saturn’s gravitational field [158], while another used the Galileo spacecraft in the Sun’s field [160].
The gravitational redshift could be improved to the 10^{−10} level using an array of laser cooled atomic clocks on board a spacecraft which would travel to within four solar radii of the Sun [180].
Modern advances in navigation using Earthorbiting atomic clocks and accurate timetransfer must routinely take gravitational redshift and timedilation effects into account. For example, the Global Positioning System (GPS) provides absolute positional accuracies of around 15 m (even better in its military mode), and 50 nanoseconds in time transfer accuracy, anywhere on Earth. Yet the difference in rate between satellite and ground clocks as a result of relativistic effects is a whopping 39 microseconds per day (46 µs from the gravitational redshift, and −7 µs from time dilation). If these effects were not accurately accounted for, GPS would fail to function at its stated accuracy. This represents a welcome practical application of GR! (For the role of GR in GPS, see [15, 16]; for a popular essay, see [287].)
Local position invariance also refers to position in time. If LPI is satisfied, the fundamental constants of nongravitational physics should be constants in time. Table 1 shows current bounds on cosmological variations in selected dimensionless constants. For discussion and references to early work, see TEGP 2.4 (c) [281] or [97]. For a comprehensive recent review both of experiments and of theoretical ideas that underly proposals for varying constants, see [262].
Experimental bounds on varying constants come in two types: bounds on the present rate of variation, and bounds on the difference between today’s value and a value in the distant past. The main example of the former type is the clock comparison test, in which highly stable atomic clocks of different fundamental type are intercompared over periods ranging from months to years (variants of the null redshift experiment). If the frequencies of the clocks depend differently on the electromagnetic fine structure constant α_{EM}, the electronproton mass ratio m_{e}/m_{P}, or the gyromagnetic ratio of the proton g_{P}, for example, then a limit on a drift of the fractional frequency difference translates into a limit on a drift of the constant(s). The dependence of the frequencies on the constants may be quite complex, depending on the atomic species involved. The most recent experiments have exploited the techniques of laser cooling and trapping, and of atom fountains, in order to achieve extreme clock stability, and compared the Rubidium87 hyperfine transition [181], the Mercury199 ion electric quadrupole transition [31], the atomic Hydrogen 1S–2S transition [111], or an optical transition in Ytterbium171 [209], against the groundstate hyperfine transition in Cesium133. These experiments show that, today, \({{\dot \alpha}_{{\rm{EM}}}}/{\alpha _{{\rm{EM}}}} < 3 \times {10^{ 15}}{\rm{y}}{{\rm{r}}^{ 1}}\).
The second type of bound involves measuring the relics of or signal from a process that occurred in the distant past and comparing the inferred value of the constant with the value measured in the laboratory today. One subtype uses astronomical measurements of spectral lines at large redshift, while the other uses fossils of nuclear processes on Earth to infer values of constants early in geological history.
Earlier comparisons of spectral lines of different atoms or transitions in distant galaxies and quasars produced bounds α_{EM} or g_{p}(m_{e}/m_{P}) on the order of a part in 10 per Hubble time [298]. Dramatic improvements in the precision of astronomical and laboratory spectroscopy, in the ability to model the complex astronomical environments where emission and absorption lines are produced, and in the ability to reach large redshift have made it possible to improve the bounds significantly. In fact, in 1999, Webb et al. [269, 193] announced that measurements of absorption lines in Mg, Al, Si, Cr, Fe, Ni, and Zn in quasars in the redshift range 0.5 < Z < 3.5 indicated a smaller value of α_{EM} in earlier epochs, namely Δα_{EM}/α_{EM} = (−0.72 ± 0.18) × 10^{−5}, corresponding to \({{\dot \alpha}_{{\rm{EM}}}}/{\alpha _{{\rm{EM}}}} = (6.4 \pm 1.4) \times {10^{ 16}}{\rm{y}}{{\rm{r}}^{{\rm{ 1}}}}\) (assuming a linear drift with time). Measurements by other groups have so far failed to confirm this nonzero effect [242, 51, 219]; a recent analysis of Mg absorption systems in quasars at \(0.4 < Z < 2.3\,{\rm{gave}}\,{{\dot \alpha}_{{\rm{EM}}}}/{\alpha _{{\rm{EM}}}} = ( 0.6 \pm 0.6) \times {10^{ 16}}{\rm{y}}{{\rm{r}}^{ 1}}\) [242].
Another important set of bounds arises from studies of the “Oklo” phenomenon, a group of natural, sustained ^{235}U fission reactors that occurred in the Oklo region of Gabon, Africa, around 1.8 billion years ago. Measurements of ore samples yielded an abnormally low value for the ratio of two isotopes of Samarium, ^{149}Sm/^{147}Sm. Neither of these isotopes is a fission product, but ^{149}Sm can be depleted by a flux of neutrons. Estimates of the neutron fluence (integrated dose) during the reactors’ “on” phase, combined with the measured abundance anomaly, yield a value for the neutron crosssection for ^{149}Sm 1.8 billion years ago that agrees with the modern value. However, the capture crosssection is extremely sensitive to the energy of a lowlying level (E ∼ 0.1 eV), so that a variation in the energy of this level of only 20 meV over a billion years would change the capture crosssection from its present value by more than the observed amount. This was first analyzed in 1976 by Shlyakter [241]. Recent reanalyses of the Oklo data [72, 116, 210] lead to a bound on \({{\dot \alpha}_{{\rm{EM}}}}\) at the level of around 5 × 10^{−17} yr^{−1}.
In a similar manner, recent reanalyses of decay rates of ^{187}Re in ancient meteorites (4.5 billion years old) gave the bound \({{\dot \alpha}_{{\rm{EM}}}}/{\alpha _{{\rm{EM}}}} < 3.4 \times {10^{ 16}}{\rm{y}}{{\rm{r}}^{{\rm{ 1}}}}\) [205].
Theoretical frameworks for analyzing EEP
Schiff’s conjecture
Because the three parts of the Einstein equivalence principle discussed above are so very different in their empirical consequences, it is tempting to regard them as independent theoretical principles. On the other hand, any complete and selfconsistent gravitation theory must possess sufficient mathematical machinery to make predictions for the outcomes of experiments that test each principle, and because there are limits to the number of ways that gravitation can be meshed with the special relativistic laws of physics, one might not be surprised if there were theoretical connections between the three subprinciples. For instance, the same mathematical formalism that produces equations describing the free fall of a hydrogen atom must also produce equations that determine the energy levels of hydrogen in a gravitational field, and thereby the ticking rate of a hydrogen maser clock. Hence a violation of EEP in the fundamental machinery of a theory that manifests itself as a violation of WEP might also be expected to show up as a violation of local position invariance. Around 1960, Schiff conjectured that this kind of connection was a necessary feature of any selfconsistent theory of gravity. More precisely, Schiff’s conjecture states that any complete, selfconsistent theory of gravity that embodies WEP necessarily embodies EEP. In other words, the validity of WEP alone guarantees the validity of local Lorentz and position invariance, and thereby of EEP.
If Schiff’s conjecture is correct, then Eötvös experiments may be seen as the direct empirical foundation for EEP, hence for the interpretation of gravity as a curvedspacetime phenomenon. Of course, a rigorous proof of such a conjecture is impossible (indeed, some special counterexamples are known [204, 194, 62]), yet a number of powerful “plausibility” arguments can be formulated.
The most general and elegant of these arguments is based upon the assumption of energy conservation. This assumption allows one to perform very simple cyclic gedanken experiments in which the energy at the end of the cycle must equal that at the beginning of the cycle. This approach was pioneered by Dicke, Nordtvedt, and Haugan (see, e.g., [124]). A system in a quantum state A decays to state B, emitting a quantum of frequency ν. The quantum falls a height H in an external gravitational field and is shifted to frequency ν′, while the system in state B falls with acceleration g_{B}. At the bottom, state A is rebuilt out of state B, the quantum of frequency ν′, and the kinetic energy m_{BgB}H that state B has gained during its fall. The energy left over must be exactly enough, m_{AgA}H, to raise state A to its original location. (Here an assumption of local Lorentz invariance permits the inertial masses m_{A} and m_{B} to be identified with the total energies of the bodies.) If g_{A} and g_{B} depend on that portion of the internal energy of the states that was involved in the quantum transition from A to B according to
(violation of WEP), then by conservation of energy, there must be a corresponding violation of LPI in the frequency shift of the form (to lowest order in hν/mc^{2})
Haugan generalized this approach to include violations of LLI [124] (TEGP 2.5 [281]).
Box 1. The THϵμ formalism

Coordinate system and conventions:
x^{0} = t: time coordinate associated with the static nature of the static spherically symmetric (SSS) gravitational field; x = (x, y, z): isotropic quasiCartesian spatial coordinates; spatial vector and gradient operations as in Cartesian space.

Matter and field variables:

m_{0a}: rest mass of particle a.

e_{a}: charge of particle a.

\(x_a^\mu (t)\): world line of particle a.

\(\upsilon _a^\mu = dx_a^\mu/dt\): coordinate velocity of particle a.

A_{μ} =: electromagnetic vector potential; E = ∇A_{0} − ∂A/∂t, B = ∇ × A.


Gravitational potential:
U(x).

Arbitrary functions:
$$T(U),H(U),\epsilon (U),\mu (U);\,{\rm{EEP}}\,{\rm{is}}\,{\rm{satisfied}}\,{\rm{if}}\,\epsilon = \mu = {(H/T)^{1/2}}{\rm{for}}\,{\rm{all}}\,U.$$ 
Action:
$$I =  \sum\limits_a {{m_{{0a}}}\int {{{(T  H\upsilon_a^2)}^{1/2}}dt}} + \sum\limits_a {{e_a}\int {{A_\mu}(x_a^\nu)\upsilon_a^\mu dt + {{(8\pi)}^{ 1}}\int {(\epsilon {E^2}  {\mu ^{ 1}}{B^2})} {d^4}x.}}$$ 
Nonmetric parameters:
$${\Gamma _0} =  c_0^2{\partial \over {\partial U}}\ln {[\epsilon {(T/H)^{1/2}}]_0},\quad {\Lambda _0} =  c_0^2{\partial \over {\partial U}}\ln {[\mu {(T/H)^{1/2}}]_0},\quad {\Upsilon _0} = 1  {(T{H^{ 1}}\epsilon \mu)_0},$$where c_{0} = (T_{0}/H_{0})^{1/2} and subscript “0” refers to a chosen point in space. If EEP is satisfied, Γ_{0} ≡ Λ_{0} ≡ ϒ_{0} ≡ 0.
The THϵμ formalism
The first successful attempt to prove Schiff’s conjecture more formally was made by Lightman and Lee [166]. They developed a framework called the THϵμ formalism that encompasses all metric theories of gravity and many nonmetric theories (see Box 1). It restricts attention to the behavior of charged particles (electromagnetic interactions only) in an external static spherically symmetric (SSS) gravitational field, described by a potential U. It characterizes the motion of the charged particles in the external potential by two arbitrary functions T(U) and H(U), and characterizes the response of electromagnetic fields to the external potential (gravitationally modified Maxwell equations) by two functions ϵ(U) and μ(U). The forms of T, H, ϵ, and μ vary from theory to theory, but every metric theory satisfies
for all U. This consequence follows from the action of electrodynamics with a “minimal” or metric coupling:
where the variables are defined in Box 1, and where F_{μν} ≡ A_{ν,μ} − A_{μ,ν}. By identifying g_{00} = T and g_{ij} = Hδ_{ij} in a SSS field, F_{i0} = E_{i} and F_{ij} = ϵ_{ijk}B_{k}, one obtains Equation (9). Conversely, every theory within this class that satisfies Equation (9) can have its electrodynamic equations cast into “metric” form. In a given nonmetric theory, the functions T, H, ϵ, and μ will depend in general on the full gravitational environment, including the potential of the Earth, Sun, and Galaxy, as well as on cosmological boundary conditions. Which of these factors has the most influence on a given experiment will depend on the nature of the experiment.
Lightman and Lee then calculated explicitly the rate of fall of a “test” body made up of interacting charged particles, and found that the rate was independent of the internal electromagnetic structure of the body (WEP) if and only if Equation (9) was satisfied. In other words, WEP ⇒ EEP and Schiff’s conjecture was verified, at least within the restrictions built into the formalism.
Certain combinations of the functions T, H, ϵ, and μ reflect different aspects of EEP. For instance, position or Udependence of either of the combinations ϵ(T/H)^{1/2} and μ(T/H)^{1/2} signals violations of LPI, the first combination playing the role of the locally measured electric charge or fine structure constant. The “nonmetric parameters” Γ_{0} and Λ_{0} (see Box 1) are measures of such violations of EEP. Similarly, if the parameter ϒ_{0} ≡ 1 − (TH^{−1}ϵμ)_{0} is nonzero anywhere, then violations of LLI will occur. This parameter is related to the difference between the speed of light c, and the limiting speed of material test particles c_{0}, given by
In many applications, by suitable definition of units, c_{0} can be set equal to unity. If EEP is valid, Γ_{0} ≡ Λ_{0} ≡ ϒ_{0} = 0 everywhere.
The rate of fall of a composite spherical test body of electromagnetically interacting particles then has the form
where \(E_{\rm{B}}^{{\rm{ES}}}\) and \(E_{\rm{B}}^{{\rm{MS}}}\) are the electrostatic and magnetostatic binding energies of the body, given by
where r_{ab} = x_{a} − x_{b}, n_{ab} = (x_{a} − x_{b})/r_{ab}, and the angle brackets denote an expectation value of the enclosed operator for the system’s internal state. Eötvös experiments place limits on the WEPviolating terms in Equation (13), and ultimately place limits on the nonmetric parameters Γ_{0} < 2 × 10^{−10} and Λ_{0} < 3 × 10^{−6}. (We set ϒ_{0} = 0 because of very tight constraints on it from tests of LLI; see Figure 2, where δ = − ϒ.) These limits are sufficiently tight to rule out a number of nonmetric theories of gravity thought previously to be viable (TEGP 2.6 (f) [281]).
The TH_{ϵμ} formalism also yields a gravitationally modified Dirac equation that can be used to determine the gravitational redshift experienced by a variety of atomic clocks. For the redshift parameter α (see Equation (6)), the results are (TEGP 2.6 (c) [281]):
The redshift is the standard one (α = 0), independently of the nature of the clock if and only if Γ_{0} ≡ Λ_{0} ≡ 0. Thus the VessotLevine rocket redshift experiment sets a limit on the parameter combination 3Γ_{0} − Λ_{0} (see Figure 3); the nullredshift experiment comparing hydrogenmaser and SCSO clocks sets a limit on \(\vert{\alpha _{\rm{H}}}  {\alpha _{{\rm{SCSO}}}}\vert = {3 \over 2}\vert{\Gamma _0}  {\Lambda _0}\vert\). Alvarez and Mann [7, 6, 8, 9, 10] extended the THϵμ formalism to permit analysis of such effects as the Lamb shift, anomalous magnetic moments and nonbaryonic effects, and placed interesting bounds on EEP violations.
The c^{2} formalism
The THϵμ formalism can also be applied to tests of local Lorentz invariance, but in this context it can be simplified. Since most such tests do not concern themselves with the spatial variation of the functions T, H, ϵ, and μ, but rather with observations made in moving frames, we can treat them as spatial constants. Then by rescaling the time and space coordinates, the charges and the electromagnetic fields, we can put the action in Box 1 into the form (TEGP 2.6 (a) [281])
where c^{2} ≡ H_{0}/(T_{0}ϵ_{0}μ_{0}) = (1 − ϒ_{0})^{−1}. This amounts to using units in which the limiting speed c_{0} of massive test particles is unity, and the speed of light is c. If c ≠ 1, LLI is violated; furthermore, the form of the action above must be assumed to be valid only in some preferred universal rest frame. The natural candidate for such a frame is the rest frame of the microwave background.
The electrodynamical equations which follow from Equation (17) yield the behavior of rods and clocks, just as in the full THϵμ formalism. For example, the length of a rod which moves with velocity V relative to the rest frame in a direction parallel to its length will be observed by a rest observer to be contracted relative to an identical rod perpendicular to the motion by a factor \(1  {V^2}/2 + {\mathcal O}({V^4})\). Notice that c does not appear in this expression, because only electrostatic interactions are involved, and c appears only in the magnetic sector of the action (17). The energy and momentum of an electromagnetically bound body moving with velocity V relative to the rest frame are given by
where \({M_R} = {M_0}  E_{\rm{B}}^{{\rm{ES}}}\), M_{0} is the sum of the particle rest masses, \(E_{\rm{B}}^{{\rm{ES}}}\) is the electrostatic binding energy of the system (see Equation (14) with \(T_0^{1/2}{H_{{0^\epsilon}_0}}^{ 1} = 1\)), and
where
Note that (c^{−2} − 1) corresponds to the parameter δ plotted in Figure 2.
The electrodynamics given by Equation (17) can also be quantized, so that we may treat the interaction of photons with atoms via perturbation theory. The energy of a photon is ħ times its frequency ω, while its momentum is ħω/c. Using this approach, one finds that the difference in round trip travel times of light along the two arms of the interferometer in the MichelsonMorley experiment is given by L_{0}(υ^{2}/c)(c^{−2} − 1). The experimental null result then leads to the bound on (c^{−2} − 1) shown on Figure 2. Similarly the anisotropy in energy levels is clearly illustrated by the tensorial terms in Equations (18, 20); by evaluating \(\tilde E_{\rm{B}}^{{\rm{ES}}ij}\) for each nucleus in the various HughesDrevertype experiments and comparing with the experimental limits on energy differences, one obtains the extremely tight bounds also shown on Figure 2.
The behavior of moving atomic clocks can also be analyzed in detail, and bounds on (c^{−2} − 1) can be placed using results from tests of time dilation and of the propagation of light. In some cases, it is advantageous to combine the c^{2} framework with a “kinematical” viewpoint that treats a general class of boost transformations between moving frames. Such kinematical approaches have been discussed by Robertson, Mansouri and Sexl, and Will (see [279]).
For example, in the “JPL” experiment, in which the phases of two hydrogen masers connected by a fiberoptic link were compared as a function of the Earth’s orientation, the predicted phase difference as a function of direction is, to first order in V, the velocity of the Earth through the cosmic background,
where \(\tilde \phi = 2\pi \nu L\), ν is the maser frequency, L = 21 km is the baseline, and where n and n_{0} are unit vectors along the direction of propagation of the light at a given time and at the initial time of the experiment, respectively. The observed limit on a diurnal variation in the relative phase resulted in the bound c^{−2} − 1 < 3 × 10^{−4}. Tighter bounds were obtained from a “twophoton absorption” (TPA) experiment, and a 1960s series of “Mössbauerrotor” experiments, which tested the isotropy of time dilation between a gamma ray emitter on the rim of a rotating disk and an absorber placed at the center [279].
The Standard Model Extension (SME)
Kostelecký and collaborators developed a useful and elegant framework for discussing violations of Lorentz symmetry in the context of the standard model of particle physics [63, 64, 155]. Called the Standard Model Extension (SME), it takes the standard SU(3) × SU(2) × U(1) field theory of particle physics, and modifies the terms in the action by inserting a variety of tensorial quantities in the quark, lepton, Higgs, and gauge boson sectors that could explicitly violate LLI. SME extends the earlier classical THϵμ and c^{2} frameworks, and the χ − g framework of Ni [194] to quantum field theory and particle physics. The modified terms split naturally into those that are odd under CPT (i.e. that violate CPT) and terms that are even under CPT. The result is a rich and complex framework, with many parameters to be analyzed and tested by experiment. Such details are beyond the scope of this review; for a review of SME and other frameworks, the reader is referred to the Living Review by Mattingly [182].
Here we confine our attention to the electromagnetic sector, in order to link the SME with the c^{2} framework discussed above. In the SME, the Lagrangian for a scalar particle ϕ charge e interacting with electrodynamics takes the form
where D_{μ}ϕ = ∂_{μ}ϕ + ieA_{μ}ϕ, where (k_{ϕ})^{μν} is a real symmetric tracefree tensor, and where (k_{F})^{μναβ} is a tensor with the symmetries of the Riemann tensor, and with vanishing double trace. It has 19 independent components. There could also be a CPTodd term in \({\mathcal L}\) of the form (k_{A})^{μϵ}_{μναβ}A^{ν}F^{αβ}, but because of a variety of preexisting theoretical and experimental constraints, it is generally set to zero.
The tensor (k_{F})^{μανβ} can be decomposed into “electric”, “magnetic”, and “oddparity” components, by defining
In many applications it is useful to use the further decomposition
The first expression is a single number, the next three are symmetric tracefree matrices, and the final is an antisymmetric matrix, accounting thereby for the 19 components of the original tensor (k_{F})^{μανβ}.
In the rest frame of the universe, these tensors have some form that is established by the global nature of the solutions of the overarching theory being used. In a frame that is moving relative to the universe, the tensors will have components that depend on the velocity of the frame, and on the orientation of the frame relative to that velocity.
In the case where the theory is rotationally symmetric in the preferred frame, the tensors (k_{ϕ})^{μν} and (k_{F})^{μναβ} can be expressed in the form
where [] around indices denote antisymmetrization, and where u^{μ} is the fourvelocity of an observer at rest in the preferred frame. With this assumption, all the tensorial quantities in Equation (24) vanish in the preferred frame, and, after suitable rescalings of coordinates and fields, the action (22) can be put into the form of the c^{2} framework, with
EEP, particle physics, and the search for new interactions
Thus far, we have discussed EEP as a principle that strictly divides the world into metric and nonmetric theories, and have implied that a failure of EEP might invalidate metric theories (and thus general relativity). On the other hand, there is mounting theoretical evidence to suggest that EEP is likely to be violated at some level, whether by quantum gravity effects, by effects arising from string theory, or by hitherto undetected interactions. Roughly speaking, in addition to the pure Einsteinian gravitational interaction, which respects EEP, theories such as string theory predict other interactions which do not. In string theory, for example, the existence of such EEPviolating fields is assured, but the theory is not yet mature enough to enable a robust calculation of their strength relative to gravity, or a determination of whether they are long range, like gravity, or short range, like the nuclear and weak interactions, and thus too short range to be detectable.
In one simple example [92], one can write the Lagrangian for the lowenergy limit of a stringinspired theory in the socalled “Einstein frame”, in which the gravitational Lagrangian is purely general relativistic:
where \({{\tilde g}_{\mu \nu}}\) is the nonphysical metric, \({{\tilde R}_{\mu \nu}}\) is the Ricci tensor derived from it, φ is a dilaton field, and \({\tilde G}\), U and \({\tilde M}\) are functions of φ. The Lagrangian includes that for the electromagnetic field F_{μν}, and that for particles, written in terms of Dirac spinors \({\tilde \psi}\). This is not a metric representation because of the coupling of φ to matter via \(\tilde M(\varphi)\) and U(φ). A conformal transformation \({{\tilde g}_{\mu \nu}} = F(\varphi){g_{\mu \nu}},\tilde \psi = F{(\varphi)^{ 3/4}}\psi\), puts the Lagrangian in the form (“Jordan” frame)
One may choose \(F(\varphi) = {\rm{const}}./\tilde M{(\varphi)^2}\) so that the particle Lagrangian takes the metric form (no explicit coupling to φ), but the electromagnetic Lagrangian will still couple nonmetrically to U(φ). The gravitational Lagrangian here takes the form of a scalartensor theory (see Section 3.3.2). But the nonmetric electromagnetic term will, in general, produce violations of EEP. For examples of specific models, see [254, 85]. Another class of nonmetric theories are included in the “varying speed of light (VSL)” theories; for a detailed review, see [178].
On the other hand, whether one views such effects as a violation of EEP or as effects arising from additional “matter” fields whose interactions, like those of the electromagnetic field, do not fully embody EEP, is to some degree a matter of semantics. Unlike the fields of the standard model of electromagnetic, weak and strong interactions, which couple to properties other than massenergy and are either short range or are strongly screened, the fields inspired by string theory could be long range (if they remain massless by virtue of a symmetry, or at best, acquire a very small mass), and can couple to massenergy, and thus can mimic gravitational fields. Still, there appears to be no way to make this precise.
As a result, EEP and related tests are now viewed as ways to discover or place constraints on new physical interactions, or as a branch of “nonaccelerator particle physics”, searching for the possible imprints of highenergy particle effects in the lowenergy realm of gravity. Whether current or proposed experiments can actually probe these phenomena meaningfully is an open question at the moment, largely because of a dearth of firm theoretical predictions.
The “fifth” force
On the phenomenological side, the idea of using EEP tests in this way may have originated in the middle 1980s, with the search for a “fifth” force. In 1986, as a result of a detailed reanalysis of Eötvös’ original data, Fischbach et al. [108] suggested the existence of a fifth force of nature, with a strength of about a percent that of gravity, but with a range (as defined by the range λ of a Yukawa potential, e^{−r/λ}/r) of a few hundred meters. This proposal dovetailed with earlier hints of a deviation from the inversesquare law of Newtonian gravitation derived from measurements of the gravity profile down deep mines in Australia, and with emerging ideas from particle physics suggesting the possible presence of very lowmass particles with gravitationalstrength couplings. During the next four years numerous experiments looked for evidence of the fifth force by searching for compositiondependent differences in acceleration, with variants of the Eötvös experiment or with freefall Galileotype experiments. Although two early experiments reported positive evidence, the others all yielded null results. Over the range between one and 10^{4} meters, the null experiments produced upper limits on the strength of a postulated fifth force between 10^{−3} and 10^{−6} of the strength of gravity. Interpreted as tests of WEP (corresponding to the limit of infiniterange forces), the results of two representative experiments from this period, the freefall Galileo experiment and the early EötWash experiment, are shown in Figure 1. At the same time, tests of the inversesquare law of gravity were carried out by comparing variations in gravity measurements up tall towers or down mines or boreholes with gravity variations predicted using the inverse square law together with Earth models and surface gravity data mathematically “continued” up the tower or down the hole. Despite early reports of anomalies, independent tower, borehole, and seawater measurements ultimately showed no evidence of a deviation. Analyses of orbital data from planetary range measurements, lunar laser ranging (LLR), and laser tracking of the LAGEOS satellite verified the inversesquare law to parts in 10^{8} over scales of 10^{3} to 10^{5} km, and to parts in 10^{9} over planetary scales of several astronomical units [250]. A consensus emerged that there was no credible experimental evidence for a fifth force of nature, of a type and range proposed by Fischbach et al. For reviews and bibliographies of this episode, see [107, 109, 110, 4, 278].
Shortrange modifications of Newtonian gravity
Although the idea of an intermediaterange violation of Newton’s gravitational law was dropped, new ideas emerged to suggest the possibility that the inversesquare law could be violated at very short ranges, below the centimeter range of existing laboratory verifications of the 1/r^{2} behavior. One set of ideas [13, 11, 221, 220] posited that some of the extra spatial dimensions that come with string theory could extend over macroscopic scales, rather than being rolled up at the Planck scale of 10^{−33} cm, which was then the conventional viewpoint. On laboratory distances large compared to the relevant scale of the extra dimension, gravity would fall off as the inverse square, whereas on short scales, gravity would fall off as 1/R^{2+n}, where n is the number of large extra dimensions. Many models favored n = 1 or n = 2. Other possibilities for effective modifications of gravity at short range involved the exchange of light scalar particles.
Following these proposals, many of the highprecision, lownoise methods that were developed for tests of WEP were adapted to carry out laboratory tests of the inverse square law of Newtonian gravitation at millimeter scales and below. The challenge of these experiments has been to distinguish gravitationlike interactions from electromagnetic and quantum mechanical (Casimir) effects. No deviations from the inverse square law have been found to date at distances between 10 µm and 10 mm[171, 130, 129, 52, 170]. For a comprehensive review of both the theory and the experiments, see [3].
Tests of PostNewtonian Gravity
Metric theories of gravity and the strong equivalence principle
Universal coupling and the metric postulates
The empirical evidence supporting the Einstein equivalence principle, discussed in the previous Section 2, supports the conclusion that the only theories of gravity that have a hope of being viable are metric theories, or possibly theories that are metric apart from very weak or shortrange nonmetric couplings (as in string theory). Therefore for the remainder of this review, we shall turn our attention exclusively to metric theories of gravity, which assume that

1.
there exists a symmetric metric,

2.
test bodies follow geodesics of the metric, and

3.
in local Lorentz frames, the nongravitational laws of physics are those of special relativity.
The property that all nongravitational fields should couple in the same manner to a single gravitational field is sometimes called “universal coupling”. Because of it, one can discuss the metric as a property of spacetime itself rather than as a field over spacetime. This is because its properties may be measured and studied using a variety of different experimental devices, composed of different nongravitational fields and particles, and, because of universal coupling, the results will be independent of the device. Thus, for instance, the proper time between two events is a characteristic of spacetime and of the location of the events, not of the clocks used to measure it.
Consequently, if EEP is valid, the nongravitational laws of physics may be formulated by taking their special relativistic forms in terms of the Minkowski metric η and simply “going over” to new forms in terms of the curved spacetime metric g, using the mathematics of differential geometry. The details of this “going over” can be found in standard textbooks (see [189, 270], TEGP 3.2. [281]).
The strong equivalence principle
In any metric theory of gravity, matter and nongravitational fields respond only to the spacetime metric g. In principle, however, there could exist other gravitational fields besides the metric, such as scalar fields, vector fields, and so on. If, by our strict definition of metric theory, matter does not couple to these fields, what can their role in gravitation theory be? Their role must be that of mediating the manner in which matter and nongravitational fields generate gravitational fields and produce the metric; once determined, however, the metric alone acts back on the matter in the manner prescribed by EEP.
What distinguishes one metric theory from another, therefore, is the number and kind of gravitational fields it contains in addition to the metric, and the equations that determine the structure and evolution of these fields. From this viewpoint, one can divide all metric theories of gravity into two fundamental classes: “purely dynamical” and “priorgeometric”.
By “purely dynamical metric theory” we mean any metric theory whose gravitational fields have their structure and evolution determined by coupled partial differential field equations. In other words, the behavior of each field is influenced to some extent by a coupling to at least one of the other fields in the theory. By “prior geometric” theory, we mean any metric theory that contains “absolute elements”, fields or equations whose structure and evolution are given a priori, and are independent of the structure and evolution of the other fields of the theory. These “absolute elements” typically include flat background metrics η or cosmic time coordinates t.
General relativity is a purely dynamical theory since it contains only one gravitational field, the metric itself, and its structure and evolution are governed by partial differential equations (Einstein’s equations). BransDicke theory and its generalizations are purely dynamical theories; the field equation for the metric involves the scalar field (as well as the matter as source), and that for the scalar field involves the metric. Rosen’s bimetric theory is a priorgeometric theory: It has a nondynamical, Riemannflat background metric η, and the field equations for the physical metric g involve η.
By discussing metric theories of gravity from this broad point of view, it is possible to draw some general conclusions about the nature of gravity in different metric theories, conclusions that are reminiscent of the Einstein equivalence principle, but that are subsumed under the name “strong equivalence principle”.
Consider a local, freely falling frame in any metric theory of gravity. Let this frame be small enough that inhomogeneities in the external gravitational fields can be neglected throughout its volume. On the other hand, let the frame be large enough to encompass a system of gravitating matter and its associated gravitational fields. The system could be a star, a black hole, the solar system, or a Cavendish experiment. Call this frame a “quasilocal Lorentz frame”. To determine the behavior of the system we must calculate the metric. The computation proceeds in two stages. First we determine the external behavior of the metric and gravitational fields, thereby establishing boundary values for the fields generated by the local system, at a boundary of the quasilocal frame “far” from the local system. Second, we solve for the fields generated by the local system. But because the metric is coupled directly or indirectly to the other fields of the theory, its structure and evolution will be influenced by those fields, and in particular by the boundary values taken on by those fields far from the local system. This will be true even if we work in a coordinate system in which the asymptotic form of g_{µν} in the boundary region between the local system and the external world is that of the Minkowski metric. Thus the gravitational environment in which the local gravitating system resides can influence the metric generated by the local system via the boundary values of the auxiliary fields. Consequently, the results of local gravitational experiments may depend on the location and velocity of the frame relative to the external environment. Of course, local nongravitational experiments are unaffected since the gravitational fields they generate are assumed to be negligible, and since those experiments couple only to the metric, whose form can always be made locally Minkowskian at a given spacetime event. Local gravitational experiments might include Cavendish experiments, measurement of the acceleration of massive selfgravitating bodies, studies of the structure of stars and planets, or analyses of the periods of “gravitational clocks”. We can now make several statements about different kinds of metric theories.

A theory which contains only the metric g yields local gravitational physics which is independent of the location and velocity of the local system. This follows from the fact that the only field coupling the local system to the environment is g, and it is always possible to find a coordinate system in which g takes the Minkowski form at the boundary between the local system and the external environment (neglecting inhomogeneities in the external gravitational field). Thus the asymptotic values of g_{μν} are constants independent of location, and are asymptotically Lorentz invariant, thus independent of velocity. General relativity is an example of such a theory.

A theory which contains the metric g and dynamical scalar fields φ_{A} yields local gravitational physics which may depend on the location of the frame but which is independent of the velocity of the frame. This follows from the asymptotic Lorentz invariance of the Minkowski metric and of the scalar fields, but now the asymptotic values of the scalar fields may depend on the location of the frame. An example is BransDicke theory, where the asymptotic scalar field determines the effective value of the gravitational constant, which can thus vary as φ varies. On the other hand, a form of velocity dependence in local physics can enter indirectly if the asymptotic values of the scalar field vary with time cosmologically. Then the rate of variation of the gravitational constant could depend on the velocity of the frame.

A theory which contains the metric g and additional dynamical vector or tensor fields or priorgeometric fields yields local gravitational physics which may have both location and velocitydependent effects.
These ideas can be summarized in the strong equivalence principle (SEP), which states that:

1.
WEP is valid for selfgravitating bodies as well as for test bodies.

2.
The outcome of any local test experiment is independent of the velocity of the (freely falling) apparatus.

3.
The outcome of any local test experiment is independent of where and when in the universe it is performed.
The distinction between SEP and EEP is the inclusion of bodies with selfgravitational interactions (planets, stars) and of experiments involving gravitational forces (Cavendish experiments, gravimeter measurements). Note that SEP contains EEP as the special case in which local gravitational forces are ignored.
The above discussion of the coupling of auxiliary fields to local gravitating systems indicates that if SEP is strictly valid, there must be one and only one gravitational field in the universe, the metric g. These arguments are only suggestive however, and no rigorous proof of this statement is available at present. Empirically it has been found that almost every metric theory other than GR introduces auxiliary gravitational fields, either dynamical or prior geometric, and thus predicts violations of SEP at some level (here we ignore quantumtheory inspired modifications to GR involving “R^{2}” terms). The one exception is Nordström’s 1913 conformallyflat scalar theory [195], which can be written purely in terms of the metric; the theory satisfies SEP, but unfortunately violates experiment by predicting no deflection of light. General relativity seems to be the only viable metric theory that embodies SEP completely. In Section 3.6, we shall discuss experimental evidence for the validity of SEP.
The parametrized postNewtonian formalism
Despite the possible existence of longrange gravitational fields in addition to the metric in various metric theories of gravity, the postulates of those theories demand that matter and nongravitational fields be completely oblivious to them. The only gravitational field that enters the equations of motion is the metric g. The role of the other fields that a theory may contain can only be that of helping to generate the spacetime curvature associated with the metric. Matter may create these fields, and they plus the matter may generate the metric, but they cannot act back directly on the matter. Matter responds only to the metric.
Thus the metric and the equations of motion for matter become the primary entities for calculating observable effects, and all that distinguishes one metric theory from another is the particular way in which matter and possibly other gravitational fields generate the metric.
The comparison of metric theories of gravity with each other and with experiment becomes particularly simple when one takes the slowmotion, weakfield limit. This approximation, known as the postNewtonian limit, is sufficiently accurate to encompass most solarsystem tests that can be performed in the foreseeable future. It turns out that, in this limit, the spacetime metric g predicted by nearly every metric theory of gravity has the same structure. It can be written as an expansion about the Minkowski metric (η_{μν} = diag(−1,1,1,1)) in terms of dimensionless gravitational potentials of varying degrees of smallness. These potentials are constructed from the matter variables (see Box 2) in imitation of the Newtonian gravitational potential
The “order of smallness” is determined according to the rules U ∼ υ ∼ Π ∼ p/ρ ϵ, υ^{i} ∼ d/dt/ddx ∼ ϵ^{1/2}, and so on (we use units in which G = c = 1; see Box 2).
A consistent postNewtonian limit requires determination of g_{00} correct through \({\mathcal O}({\epsilon ^2})\), g_{0i} through \({\mathcal O}({\epsilon ^{3/2}})\), and g_{ij} through \({\mathcal O}(\epsilon)\) (for details see TEGP 4.1 [281]). The only way that one metric theory differs from another is in the numerical values of the coefficients that appear in front of the metric potentials. The parametrized postNewtonian (PPN) formalism inserts parameters in place of these coefficients, parameters whose values depend on the theory under study. In the current version of the PPN formalism, summarized in Box 2, ten parameters are used, chosen in such a manner that they measure or indicate general properties of metric theories of gravity (see Table 2). Under reasonable assumptions about the kinds of potentials that can be present at postNewtonian order (basically only Poissonlike potentials), one finds that ten PPN parameters exhaust the possibilities.
The parameters γ and β are the usual EddingtonRobertsonSchiff parameters used to describe the “classical” tests of GR, and are in some sense the most important; they are the only nonzero parameters in GR and scalartensor gravity. The parameter ξ is nonzero in any theory of gravity that predicts preferredlocation effects such as a galaxyinduced anisotropy in the local gravitational constant G_{l} (also called “Whitehead” effects); α_{1}, α_{2}, α_{3} measure whether or not the theory predicts postNewtonian preferredframe effects; α_{3} ζ_{1}, ζ_{2}, ζ_{3}, ζ_{4} measure whether or not the theory predicts violations of global conservation laws for total momentum. In Table 2 we show the values these parameters take

1.
in GR,

2.
in any theory of gravity that possesses conservation laws for total momentum, called “semiconservative” (any theory that is based on an invariant action principle is semiconservative), and

3.
in any theory that in addition possesses six global conservation laws for angular momentum, called “fully conservative” (such theories automatically predict no postNewtonian preferredframe effects).
Semiconservative theories have five free PPN parameters (γ, β, ξ, α_{1}, α_{2}) while fully conservative theories have three (γ, β, ξ).
The PPN formalism was pioneered by Kenneth Nordtvedt [197], who studied the postNewtonian metric of a system of gravitating point masses, extending earlier work by Eddington, Robertson and Schiff (TEGP 4.2 [281]). Will [274] generalized the framework to perfect fluids. A general and unified version of the PPN formalism was developed by Will and Nordtvedt. The canonical version, with conventions altered to be more in accord with standard textbooks such as [189], is discussed in detail in TEGP 4 [281]. Other versions of the PPN formalism have been developed to deal with point masses with charge, fluid with anisotropic stresses, bodies with strong internal gravity, and postpostNewtonian effects (TEGP 4.2, 14.2 [281]).
Box 2. The Parametrized PostNewtonian formalism

Coordinate system:
The framework uses a nearly globally Lorentz coordinate system in which the coordinates are (t, x^{1},x^{2},x^{3}). Threedimensional, Euclidean vector notation is used throughout. All coordinate arbitrariness (“gauge freedom”) has been removed by specialization of the coordinates to the standard PPN gauge (TEGP 4.2 [281]). Units are chosen so that G = c = 1, where G is the physically measured Newtonian constant far from the solar system.

Matter variables:

ρ: density of rest mass as measured in a local freely falling frame momentarily comoving with the gravitating matter.

υ^{i}= (dx^{i}/dt): coordinate velocity of the matter.

w^{i}: coordinate velocity of the PPN coordinate system relative to the mean restframe of the universe.

p: pressure as measured in a local freely falling frame momentarily comoving with the matter.

Π: internal energy per unit rest mass (it includes all forms of nonrestmass, nongravitational energy, e.g., energy of compression and thermal energy).


PPN parameters:
$$\gamma, \,\beta, \,\xi, \,{\alpha _1},\,{\alpha _2},\,{\alpha _3},\,{\varsigma _1},\,{\varsigma _2},\,{\varsigma _3},\,{\varsigma _4}.$$ 
Metric:
$$\begin{array}{*{20}c} {{g_{00}} =  1 + 2U  2\beta {U^2}  2\xi {\Phi _W} + (2\gamma + 2 + {\alpha _3} + {\zeta _1}  2\xi){\Phi _1} + 2(3\gamma  2\beta + 1 + {\zeta _2} + \xi){\Phi _2}}\\ {\quad \quad + 2(1 + {\zeta _3}){\Phi _3} + 2(3\gamma + 3{\zeta _4}  2\xi){\Phi _4}  ({\zeta _1}  2\xi){\mathcal A}  ({\alpha _1}  {\alpha _2}  {\alpha _3}){w^2}U  {\alpha _2}{w^i}{w^j}{U_{ij}}}\\ {\quad \quad + (2{\alpha _3}  {\alpha _1}){w^i}{V_i} + {\mathcal O}({\epsilon ^3}),}\\ {{g_{0i}} =  {1 \over 2}(4\gamma + 3 + {\alpha _1}  {\alpha _2} + {\zeta _1}  2\xi){V_i}  {1 \over 2}(1 + {\alpha _2}  {\zeta _1} + 2\xi){W_i}  {1 \over 2}({\alpha _1}  2{\alpha _2}){w^i}U}\\ {\quad \quad  {\alpha _2}{w^j}{U_{ij}} + {\mathcal O}({\epsilon ^{5/2}}),}\\ {{g_{ij}} = (1 + 2\gamma U){\delta _{ij}} + {\mathcal O}({\epsilon ^2}).}\\ \end{array}$$ 
Metric potentials:
$$\begin{array}{*{20}c}{U = \int {{{\rho \prime} \over {\vert {\rm{x}}  {\rm{x}}\prime \vert}}{d^3}x\prime ,\quad \quad \quad \quad \quad \quad \quad}} \\ {{U_{ij}} = \int {{\rho \prime {{(x  x\prime)}_i}{{(x  x\prime)}_j}} \over {\vert {\rm{x}}  {\rm{x}}\prime {\vert ^3}}}{d^3}x\prime ,\quad \quad \quad} \\ {\quad \quad \quad \quad \quad {\Phi _W} = \int {{\rho \prime \rho \prime \prime ({\rm{x}}  {\rm{x}}\prime)} \over {\vert {\rm{x}}  {\rm{x}}\prime {\vert ^3}}}\cdot\left({{{{\rm{x}}\prime  {\rm{x}}\prime \prime} \over {\vert {\rm{x}}  {\rm{x}}\prime \prime \vert}}  {{{\rm{x}}  {\rm{x}}\prime \prime} \over {\vert {\rm{x}}\prime  {\rm{x}}\prime \prime \vert}}} \right){d^3}x\prime {d^3}x\prime \prime ,} \\ {A = \int {{\rho \prime {{[{\rm{v}}\prime \cdot({\rm{x}}  {\rm{x}}\prime)]}^2}} \over {\vert {\rm{x}}  {\rm{x}}\prime {\vert ^3}}}{d^3}x\prime ,\quad \quad \quad \quad} \\ {{\Phi _1} = \int {{{\rho \prime \upsilon {\prime ^2}} \over {\vert {\rm{x}}  {\rm{x}}\prime \vert}}{d^3}x\prime ,\quad \quad \quad \quad \quad \quad}} \\ {{\Phi _2} = \int {{{\rho \prime U\prime} \over {\vert {\rm{x}}  {\rm{x}}\prime \vert}}{d^3}x\prime ,\quad \quad \quad \quad \quad \quad}} \\ {{\Phi _3} = \int {{{\rho \prime \Pi \prime} \over {\vert {\rm{x}}  {\rm{x}}\prime \vert}}} {d^3}x\prime ,\quad \quad \quad \quad \quad \quad} \\ {{\Phi _4} = \int {{{p\prime} \over {\vert {\rm{x}}  {\rm{x}}\prime \vert}}{d^3}x\prime ,\quad \quad \quad \quad \quad \quad}} \\ {{V_i} = \int {{{\rho \prime \upsilon {\prime _i}} \over {\vert {\rm{x}}  {\rm{x}}\prime \vert}}{d^3}x\prime ,\quad \quad \quad \quad \quad \quad}} \\ {\,\,\,\,{W_i} = \int {{\rho \prime [{\rm{v}}\prime \cdot({\rm{x}}  {\rm{x}}\prime)]{{(x  x\prime)}_i}} \over {\vert {\rm{x}}  {\rm{x}}\prime {\vert ^3}}}{d^3}x\prime .\quad \quad} \\ \end{array}$$ 
Stressenergy tensor (perfect fluid):
$$\begin{array}{*{20}c} {{T^{00}} = \rho (1 + \prod + {v^2} + 2U)}\\ {{T^{0i}} = \rho {v^i}\left({1 + \prod + {v^2} + 2U + {p \over p}} \right),}\\ {{T^{ij}} = \rho {v^i}{v^j}\left({1 + \prod + {v^2} + 2U + {p \over p}} \right) + p{\delta ^{ij}}(1  2\gamma U).}\\ \end{array}$$ 
Equations of motion:

Stressed matter: T^{μν;ν} = 0.

Test bodies: \({{{d^2}{x^\mu}} \over {d{\lambda ^2}}} + {\Gamma ^\mu}\nu \lambda {{d{x^\nu}} \over {d\lambda}}{{d{x^\lambda}} \over {d\lambda}} = 0\).

Maxwell’s equations: F^{μν;ν} = 4πJ^{μ}, F_{μν} = A_{ν;μ} − A_{μ;ν}.

Competing theories of gravity
One of the important applications of the PPN formalism is the comparison and classification of alternative metric theories of gravity. The population of viable theories has fluctuated over the years as new effects and tests have been discovered, largely through the use of the PPN framework, which eliminated many theories thought previously to be viable. The theory population has also fluctuated as new, potentially viable theories have been invented.
In this review, we shall focus on GR, the general class of scalartensor modifications of it, of which the JordanFierzBransDicke theory (BransDicke, for short) is the classic example, and vectortensor theories. The reasons are severalfold:

A full compendium of alternative theories circa 1981 is given in TEGP 5 [281].

Many alternative metric theories developed during the 1970s and 1980s could be viewed as “strawman” theories, invented to prove that such theories exist or to illustrate particular properties. Few of these could be regarded as wellmotivated theories from the point of view, say, of field theory or particle physics.

A number of theories fall into the class of “priorgeometric” theories, with absolute elements such as a flat background metric in addition to the physical metric. Most of these theories predict “preferredframe” effects, that have been tightly constrained by observations (see Section 3.6.2). An example is Rosen’s bimetric theory.

A large number of alternative theories of gravity predict gravitational wave emission substantially different from that of general relativity, in strong disagreement with observations of the binary pulsar (see Section 7).

Scalartensor modifications of GR have become very popular in unification schemes such as string theory, and in cosmological model building. Because the scalar fields could be massive, the potentials in the postNewtonian limit could be modified by Yukawalike terms.

Vectortensor theories have attracted recent attention, in the spirit of the SME (see Section 2.2.4), as models for violations of Lorentz invariance in the gravitational sector.
General relativity
The metric g is the sole dynamical field, and the theory contains no arbitrary functions or parameters, apart from the value of the Newtonian coupling constant G, which is measurable in laboratory experiments. Throughout this article, we ignore the cosmological constant Λ_{C}. We do this despite recent evidence, from supernova data, of an accelerating universe, which would indicate either a nonzero cosmological constant or a dynamical “dark energy” contributing about 70 percent of the critical density. Although Λ_{C} has significance for quantum field theory, quantum gravity, and cosmology, on the scale of the solarsystem or of stellar systems its effects are negligible, for the values of Ac inferred from supernova observations.
The field equations of GR are derivable from an invariant action principle δI = 0, where
where R is the Ricci scalar, and I_{m} is the matter action, which depends on matter fields ψ_{m} universally coupled to the metric g. By varying the action with respect to g_{μν}, we obtain the field equations
where T_{μν} is the matter energymomentum tensor. General covariance of the matter action implies the equations of motion T^{μν;ν} = 0; varying I_{m} with respect to ψ_{m} yields the matter field equations of the Standard Model. By virtue of the absence of priorgeometric elements, the equations of motion are also a consequence of the field equations via the Bianchi identities G^{μν;ν} = 0.
The general procedure for deriving the postNewtonian limit of metric theories is spelled out in TEGP 5.1 [281], and is described in detail for GR in TEGP 5.2 [281]. The PPN parameter values are listed in Table 3.
Scalartensor theories
These theories contain the metric g, a scalar field φ, a potential function V(φ), and a coupling function A(φ) (generalizations to more than one scalar field have also been carried out [73]). For some purposes, the action is conveniently written in a nonmetric representation, sometimes denoted the “Einstein frame”, in which the gravitational action looks exactly like that of GR:
where \(\tilde R \equiv {\tilde g^{\mu \nu}}{\tilde R_{\mu \nu}}\) is the Ricci scalar of the “Einstein” metric \({{\tilde g}_{\mu \nu}}\). (Apart from the scalar potential term V(φ), this corresponds to Equation (28) with \(\tilde G(\varphi) \equiv {(4\pi G)^{ 1}},U(\varphi) \equiv 1\), and \(\tilde M(\varphi) \propto A(\varphi).)\). This representation is a “nonmetric” one because the matter fields ψ_{m} couple to a combination of φ and \({{\tilde g}_{\mu \nu}}\). Despite appearances, however, it is a metric theory, because it can be put into a metric representation by identifying the “physical metric”
The action can then be rewritten in the metric form
where
The Einstein frame is useful for discussing general characteristics of such theories, and for some cosmological applications, while the metric representation is most useful for calculating observable effects. The field equations, postNewtonian limit and PPN parameters are discussed in TEGP 5.3 [281], and the values of the PPN parameters are listed in Table 3.
The parameters that enter the postNewtonian limit are
where ϕ_{0} is the value of ϕ today far from the system being studied, as determined by appropriate cosmological boundary conditions. In BransDicke theory (ω(ϕ) = ω_{bd} = const.), the larger the value of ω_{BD}, the smaller the effects of the scalar field, and in the limit ω_{BD} → ∞ (α_{0} → 0), the theory becomes indistinguishable from GR in all its predictions. In more general theories, the function ω(ϕ) could have the property that, at the present epoch, and in weakfield situations, the value of the scalar field ϕ_{0} is such that ω is very large and Λ is very small (theory almost identical to GR today), but that for past or future values of ϕ, or in strongfield regions such as the interiors of neutron stars, ω and Λ could take on values that would lead to significant differences from GR. It is useful to point out that all versions of scalartensor gravity predict that γ ≤ 1 (see Table 3).
Damour and EspositoFarèse [73] have adopted an alternative parametrization of scalartensor theories, in which one expands ln A(φ) about a cosmological background field value φ_{0}:
A precisely linear coupling function produces BransDicke theory, with \(\alpha _0^2 = 1/(2{\omega _{{\rm{BD}}}} + 3)\), or \(1/(2 + {\omega _{{\rm{BD}}}}) = 2\alpha _0^2/(1 + \alpha _0^2)\). The function ln A(φ) acts as a potential for the scalar field φ within matter, and, if β_{0} > 0, then during cosmological evolution, the scalar field naturally evolves toward the minimum of the potential, i.e. toward α_{0} ≈ 0, ω → ∞, or toward a theory close to, though not precisely GR [80, 81]. Estimates of the expected relic deviations from GR today in such theories depend on the cosmological model, but range from 10^{−5} to a few times 10^{−7} for γ − 1.
Negative values of β_{0} correspond to a “locally unstable” scalar potential (the overall theory is still stable in the sense of having no tachyons or ghosts). In this case, objects such as neutron stars can experience a “spontaneous scalarization”, whereby the interior values of φ can take on values very different from the exterior values, through nonlinear interactions between strong gravity and the scalar field, dramatically affecting the stars’ internal structure and leading to strong violations of SEP. On the other hand, in the case β_{0} < 0, one must confront that fact that, with an unstable φ potential, cosmological evolution would presumably drive the system away from the peak where α_{0} ≈ 0, toward parameter values that could be excluded by solar system experiments.
Scalar fields coupled to gravity or matter are also ubiquitous in particlephysicsinspired models of unification, such as string theory [254, 176, 85, 82, 83]. In some models, the coupling to matter may lead to violations of EEP, which could be tested or bounded by the experiments described in Section 2.1. In many models the scalar field could be massive; if the Compton wavelength is of macroscopic scale, its effects are those of a “fifth force”. Only if the theory can be cast as a metric theory with a scalar field of infinite range or of range long compared to the scale of the system in question (solar system) can the PPN framework be strictly applied. If the mass of the scalar field is sufficiently large that its range is microscopic, then, on solarsystem scales, the scalar field is suppressed, and the theory is essentially equivalent to general relativity.
Vectortensor theories
These theories contain the metric g and a dynamical, typically timelike, fourvector field u^{μ}. In some models, the fourvector is unconstrained, while in others, called EinsteinÆther theories it is constrained to be timelike with unit norm. The most general action for such theories that is quadratic in derivatives of the vector is given by
where
The coefficients c_{i} are arbitrary. In the unconstrained theories, λ ≡ 0 and ω is arbitrary. In the constrained theories, λ is a Lagrange multiplier, and by virtue of the constraint u_{μ}u^{μ} = −1, the factor ωu_{μ}u^{μ} in front of the Ricci scalar can be absorbed into a rescaling of G; equivalently, in the constrained theories, we can set ω = 0. Note that the possible term u^{μuν}R_{μν} can be shown under integration by parts to be equivalent to a linear combination of the terms involving c_{2} and c_{3}.
Unconstrained theories were studied during the 1970s as “strawman” alternatives to GR. In addition to having up to four arbitrary parameters, they also left the magnitude of the vector field arbitrary, since it satisfies a linear homogenous vacuum field equation of the form \({\mathcal L}{u^\mu} = 0({c_4} = 0\) in all such cases studied). Indeed, this latter fact was one of most serious defects of these theories. Each theory studied corresponds to a special case of the action (39), all with λ ≡ 0:

General vectortensor theory; ω, τ, ϵ, η (see TEGP 5.4 [281])
The gravitational Lagrangian for this class of theories had the form R+ωu_{μ}u^{μ} R+ηu^{μ}u^{ν} R_{μν} −ϵF_{μν}F^{μν} + τ∇_{μ}u_{ν}∇^{μ}u^{ν}, where F_{μν} = ∇_{μ}u_{ν}∇ − ∇_{ν}u_{μ}, corresponding to the values c_{1} = 2ϵ − τ, c_{2} = −η, c_{1} + c_{2} + c_{3} = −τ, c_{4} = 0. In these theories γ, β, α_{1}, and α_{2} are complicated functions of the parameters and of u^{2} = −u^{μ}u_{μ}, while the rest vanish.

WillNordtvedt theory (see [290])
This is the special case c_{1} = −1, c_{2} = c_{3} = c_{4} = 0. In this theory, the PPN parameters are given by γ = β = 1, α_{2} = u^{2}/(1 + u^{2}/2), and zero for the rest.

HellingsNordtvedt theory; ω (see [128])
This is the special case c_{1} = 2, c_{2} = 2ω, c_{1} + c_{2} + c_{3} = 0 = c_{4}. Here γ, β, α_{1} and α_{2} are complicated functions of the parameters and of u^{2}, while the rest vanish.
The EinsteinÆther theories were motivated in part by a desire to explore possibilities for violations of Lorentz invariance in gravity, in parallel with similar studies in matter interactions, such as the SME. The general class of theories was analyzed by Jacobson and collaborators [137, 183, 138, 99, 113], motivated in part by [156].
Analyzing the postNewtonian limit, they were able to infer values of the PPN parameters γ and β as follows [113]:
where c_{123} = c_{1} + c_{2} + c_{3}, c_{13} = c_{1} − c_{3}, c_{14} = c_{1} − c_{4}, subject to the constraints \({c_{123}} \neq 0,{c_{14}} \neq 2,2{c_1}  c_1^2 + c_3^2 \neq 0\).
By requiring that gravitational wave modes have real (as opposed to imaginary) frequencies, one can impose the bounds c_{1}/(c_{1} + c_{4}) ≥ 0 and (c_{1} + c_{2} + c_{3})/(c_{1} + c_{4}) ≥ 0. Considerations of positivity of energy impose the constraints c_{1} > 0, c_{1} + c_{4} > 0 and c_{1} + c_{2} + c_{3} > 0.
Tests of the parameter γ
With the PPN formalism in hand, we are now ready to confront gravitation theories with the results of solarsystem experiments. In this section we focus on tests of the parameter γ, consisting of the deflection of light and the time delay of light.
The deflection of light
A light ray (or photon) which passes the Sun at a distance d is deflected by an angle
(TEGP 7.1 [281]), where M_{⊙} is the mass of the Sun and Φ is the angle between the EarthSun line and the incoming direction of the photon (see Figure 4). For a grazing ray, d ≈ d_{⊙}, Φ ≈ 0, and
independent of the frequency of light. Another, more useful expression gives the change in the relative angular separation between an observed source of light and a nearby reference source as both rays pass near the Sun:
where d and d_{r} are the distances of closest approach of the source and reference rays respectively, Φ_{r} is the angular separation between the Sun and the reference source, and χ is the angle between the Sunsource and the Sunreference directions, projected on the plane of the sky (see Figure 4). Thus, for example, the relative angular separation between the two sources may vary if the line of sight of one of them passes near the Sun (d ∼ R_{⊙}, d_{r} ≫ d, χ varying with time).
It is interesting to note that the classic derivations of the deflection of light that use only the corpuscular theory of light (Cavendish 1784, von Soldner 1803 [277]), or the principle of equivalence (Einstein 1911), yield only the “1/2” part of the coefficient in front of the expression in Equation (46). But the result of these calculations is the deflection of light relative to local straight lines, as established for example by rigid rods; however, because of space curvature around the Sun, determined by the PPN parameter γ, local straight lines are bent relative to asymptotic straight lines far from the Sun by just enough to yield the remaining factor “γ/2”. The first factor “1/2” holds in any metric theory, the second “γ/2” varies from theory to theory. Thus, calculations that purport to derive the full deflection using the equivalence principle alone are incorrect.
The prediction of the full bending of light by the Sun was one of the great successes of Einstein’s GR. Eddington’s confirmation of the bending of optical starlight observed during a solar eclipse in the first days following World War I helped make Einstein famous. However, the experiments of Eddington and his coworkers had only 30 percent accuracy, and succeeding experiments were not much better: The results were scattered between one half and twice the Einstein value (see Figure 5), and the accuracies were low.
However, the development of radiointerferometery, and later of verylongbaseline radio interferometry (VLBI), produced greatly improved determinations of the deflection of light. These techniques now have the capability of measuring angular separations and changes in angles to accuracies better than 100 microarcseconds. Early measurements took advantage of a series of heavenly coincidences: Each year, groups of strong quasistellar radio sources pass very close to the Sun (as seen from the Earth), including the group 3C273, 3C279, and 3C48, and the group 0111+02, 0119+11, and 0116+08. As the Earth moves in its orbit, changing the lines of sight of the quasars relative to the Sun, the angular separation δθ between pairs of quasars varies (see Equation (48)). The time variation in the quantities d, d_{r}, χ, and Φ_{r} in Equation (48) is determined using an accurate ephemeris for the Earth and initial directions for the quasars, and the resulting prediction for δθ as a function of time is used as a basis for a leastsquares fit of the measured δθ, with one of the fitted parameters being the coefficient \({1 \over 2}(1 + \gamma)\). A number of measurements of this kind over the period 1969–1975 yielded an accurate determination of the coefficient \({1 \over 2}(1 + \gamma)\). A 1995 VLBI measurement using 3C273 and 3C279 yielded (1 + γ)/2 = 0.9996 ± 0.0017 [164].
In recent years, transcontinental and intercontinental VLBI observations of quasars and radio galaxies have been made primarily to monitor the Earth’s rotation (“VLBI” in Figure 5). These measurements are sensitive to the deflection of light over almost the entire celestial sphere (at 90° from the Sun, the deflection is still 4 milliarcseconds). A 2004 analysis of almost 2 million VLBI observations of 541 radio sources, made by 87 VLBI sites yielded (1 + γ)/2 = 0.99992 ± 0.00023, or equivalently, γ − 1 = (−1.7 ± 4.5) × 10^{−4} [240].
Analysis of observations made by the Hipparcos optical astrometry satellite yielded a test at the level of 0.3 percent [115]. A VLBI measurement of the deflection of light by Jupiter was reported; the predicted deflection of about 300 microarcseconds was seen with about 50 percent accuracy [257]. The results of lightdeflection measurements are summarized in Figure 5.
The time delay of light
A radar signal sent across the solar system past the Sun to a planet or satellite and returned to the Earth suffers an additional nonNewtonian delay in its roundtrip travel time, given by (see Figure 4)
where x_{e} (x⊕) are the vectors, and r_{e} (r_{⊕}) are the distances from the Sun to the source (Earth), respectively (TEGP 7.2 [281]). For a ray which passes close to the Sun,
where d is the distance of closest approach of the ray in solar radii, and r is the distance of the planet or satellite from the Sun, in astronomical units.
In the two decades following Irwin Shapiro’s 1964 discovery of this effect as a theoretical consequence of GR, several highprecision measurements were made using radar ranging to targets passing through superior conjunction. Since one does not have access to a “Newtonian” signal against which to compare the roundtrip travel time of the observed signal, it is necessary to do a differential measurement of the variations in roundtrip travel times as the target passes through superior conjunction, and to look for the logarithmic behavior of Equation (50). In order to do this accurately however, one must take into account the variations in roundtrip travel time due to the orbital motion of the target relative to the Earth. This is done by using radarranging (and possibly other) data on the target taken when it is far from superior conjunction (i.e. when the timedelay term is negligible) to determine an accurate ephemeris for the target, using the ephemeris to predict the PPN coordinate trajectory x_{e}(t) near superior conjunction, then combining that trajectory with the trajectory of the Earth x_{⊕}(t) to determine the Newtonian roundtrip time and the logarithmic term in Equation (50). The resulting predicted roundtrip travel times in terms of the unknown coefficient \({1 \over 2}(1 + \gamma)\) are then fit to the measured travel times using the method of leastsquares, and an estimate obtained for \({1 \over 2}(1 + \gamma)\).
The targets employed included planets, such as Mercury or Venus, used as passive reflectors of the radar signals (“passive radar”), and artificial satellites, such as Mariners 6 and 7, Voyager 2, the Viking Mars landers and orbiters, and the Cassini spacecraft to Saturn, used as active retransmitters of the radar signals (“active radar”).
The results for the coefficient \({1 \over 2}(1 + \gamma)\) of all radar timedelay measurements performed to date (including a measurement of the oneway time delay of signals from the millisecond pulsar PSR 1937+21) are shown in Figure 5 (see TEGP 7.2 [281] for discussion and references). The 1976 Viking experiment resulted in a 0.1 percent measurement [222].
A significant improvement was reported in 2003 from Doppler tracking of the Cassini spacecraft while it was on its way to Saturn [29], with a result γ − 1 = (2.1 ± 2.3) × 10^{−5}. This was made possible by the ability to do Doppler measurements using both Xband (7175 MHz) and Kaband (34316 MHz) radar, thereby significantly reducing the dispersive effects of the solar corona. In addition, the 2002 superior conjunction of Cassini was particularly favorable: With the spacecraft at 8.43 astronomical units from the Sun, the distance of closest approach of the radar signals to the Sun was only 1.6 R_{⊕}.
From the results of the Cassini experiment, we can conclude that the coefficient \({1 \over 2}(1 + \gamma)\) must be within at most 0.0012 percent of unity. Scalartensor theories must have ω > 40000 to be compatible with this constraint.
Shapiro time delay and the speed of gravity
In 2001, Kopeikin [147] suggested that a measurement of the time delay of light from a quasar as the light passed by the planet Jupiter could be used to measure the speed of the gravitational interaction. He argued that, since Jupiter is moving relative to the solar system, and since gravity propagates with a finite speed, the gravitational field experienced by the light ray should be affected by gravity’s speed, since the field experienced at one time depends on the location of the source a short time earlier, depending on how fast gravity propagates. According to his calculations, there should be a post^{1/2}Newtonian correction to the normal Shapiro timedelay formula (49) which depends on the velocity of Jupiter and on the velocity of gravity. On September 8, 2002, Jupiter passed almost in front of a quasar, and Kopeikin and Fomalont made precise measurements of the Shapiro delay with picosecond timing accuracy, and claimed to have measured the correction term to about 20 percent [112, 153, 148, 149].
However, several authors pointed out that this 1.5PN effect does not depend on the speed of propagation of gravity, but rather only depends on the speed of light [14, 288, 232, 49, 233]. Intuitively, if one is working to only first order in υ/c, then all that counts is the uniform motion of the planet, Jupiter (its acceleration about the Sun contributes a higherorder, unmeasurably small effect). But if that is the case, then the principle of relativity says that one can view things from the rest frame of Jupiter. In this frame, Jupiter’s gravitational field is static, and the speed of propagation of gravity is irrelevant. A detailed postNewtonian calculation of the effect was done using a variant of the PPN framework, in a class of theories in which the speed of gravity could be different from that of light [288], and found explicitly that, at first order in υ/c, the effect depends on the speed of light, not the speed of gravity, in line with intuition. Effects dependent upon the speed of gravity show up only at higher order in υ/c. Kopeikin gave a number of arguments in opposition to this interpretation [149, 151, 150, 152]. On the other hand, the υ/c correction term does show a dependence on the PPN parameter α_{1}, which could be nonzero in theories of gravity with a differing speed c_{g} of gravity (see Equation (7) of [288]). But existing tight bounds on α_{1} from other experiments (see Table 4) already far exceed the capability of the Jupiter VLBI experiment.
The perihelion shift of Mercury
The explanation of the anomalous perihelion shift of Mercury’s orbit was another of the triumphs of GR. This had been an unsolved problem in celestial mechanics for over half a century, since the announcement by Le Verrier in 1859 that, after the perturbing effects of the planets on Mercury’s orbit had been accounted for, and after the effect of the precession of the equinoxes on the astronomical coordinate system had been subtracted, there remained in the data an unexplained advance in the perihelion of Mercury. The modern value for this discrepancy is 43 arcseconds per century. A number of ad hoc proposals were made in an attempt to account for this excess, including, among others, the existence of a new planet Vulcan near the Sun, a ring of planetoids, a solar quadrupole moment and a deviation from the inversesquare law of gravitation, but none was successful. General relativity accounted for the anomalous shift in a natural way without disturbing the agreement with other planetary observations.
The predicted advance per orbit \(\Delta \tilde \omega\), including both relativistic PPN contributions and the Newtonian contribution resulting from a possible solar quadrupole moment, is given by
where m ≡ m_{1} + m_{2} and μ ≡ m_{1}m_{2}/m are the total mass and reduced mass of the twobody system respectively; p ≡ a(1 − e^{2}) is the semilatus rectum of the orbit, with the semimajor axis a and the eccentricity e; R is the mean radius of the oblate body; and J_{2} is a dimensionless measure of its quadrupole moment, given by J_{2} = (C − A)/m_{1}R^{2}, where C and A are the moments of inertia about the body’s rotation and equatorial axes, respectively (for details of the derivation see TEGP 7.3 [281]). We have ignored preferredframe and galaxyinduced contributions to \(\Delta \tilde \omega\); these are discussed in TEGP 8.3 [281].
The first term in Equation (51) is the classical relativistic perihelion shift, which depends upon the PPN parameters γ and β. The second term depends upon the ratio of the masses of the two bodies; it is zero in any fully conservative theory of gravity (α_{1} ≡ α_{2} ≡ α_{3} ≡ ζ_{2} ≡ 0); it is also negligible for Mercury, since µ/m ≈ m_{Merc}/M_{⊙} ≈ 2 × 10^{−7}. We shall drop this term henceforth.
The third term depends upon the solar quadrupole moment J_{2}. For a Sun that rotates uniformly with its observed surface angular velocity, so that the quadrupole moment is produced by centrifugal flattening, one may estimate J_{2} to be ∼ 1 × 10^{−7}. This actually agrees reasonably well with values inferred from rotating solar models that are in accord with observations of the normal modes of solar oscillations (helioseismology); the latest inversions of helioseismology data give J_{2} = (2.2 ± 0.1) × 10^{−7} [207, 211, 230, 184]. Substituting standard orbital elements and physical constants for Mercury and the Sun we obtain the rate of perihelion shift \({\dot \tilde \omega}\), in seconds of arc per century,
Now, the measured perihelion shift of Mercury is known accurately: After the perturbing effects of the other planets have been accounted for, the excess shift is known to about 0.1 percent from radar observations of Mercury between 1966 and 1990 [238]. Analysis of data taken since 1990 could improve the accuracy. The solar oblateness effect is smaller than the observational error, so we obtain the PPN bound 2γ − β − 1 < 3 × 10^{−3}.
Tests of the strong equivalence principle
The next class of solarsystem experiments that test relativistic gravitational effects may be called tests of the strong equivalence principle (SEP). In Section 3.1.2 we pointed out that many metric theories of gravity (perhaps all except GR) can be expected to violate one or more aspects of SEP. Among the testable violations of SEP are a violation of the weak equivalence principle for gravitating bodies that leads to perturbations in the EarthMoon orbit, preferredlocation and preferredframe effects in the locally measured gravitational constant that could produce observable geophysical effects, and possible variations in the gravitational constant over cosmological timescales.
The Nordtvedt effect and the lunar Eötvös experiment
In a pioneering calculation using his early form of the PPN formalism, Nordtvedt [196] showed that many metric theories of gravity predict that massive bodies violate the weak equivalence principle — that is, fall with different accelerations depending on their gravitational selfenergy. Dicke [228] argued that such an effect would occur in theories with a spatially varying gravitational constant, such as scalartensor gravity. For a spherically symmetric body, the acceleration from rest in an external gravitational potential U has the form
where E_{g} is the negative of the gravitational selfenergy of the body (E_{g} > 0). This violation of the massivebody equivalence principle is known as the “Nordtvedt effect”. The effect is absent in GR (η_{N} = 0) but present in scalartensor theory (η_{N}= 1/(2 + ω) +4Λ). The existence of the Nordtvedt effect does not violate the results of laboratory Eötvös experiments, since for laboratorysized objects E_{g}/m ≤ 10^{−27}, far below the sensitivity of current or future experiments. However, for astronomical bodies, E_{g}/m may be significant (3.6 × 10^{−6} for the Sun, 10^{−8} for Jupiter, 4.6 × 10^{−10} for the Earth, 0.2 × 10^{−10} for the Moon). If the Nordtvedt effect is present (η_{N} ≠ 0) then the Earth should fall toward the Sun with a slightly different acceleration than the Moon. This perturbation in the EarthMoon orbit leads to a polarization of the orbit that is directed toward the Sun as it moves around the EarthMoon system, as seen from Earth. This polarization represents a perturbation in the EarthMoon distance of the form
where ω_{0} and ω_{s} are the angular frequencies of the orbits of the Moon and Sun around the Earth (see TEGP 8.1 [281] for detailed derivations and references; for improved calculations of the numerical coefficient, see [201, 89]).
Since August 1969, when the first successful acquisition was made of a laser signal reflected from the Apollo 11 retroreflector on the Moon, the LLR experiment has made regular measurements of the roundtrip travel times of laser pulses between a network of observatories and the lunar retroreflectors, with accuracies that are at the 50 ps (1 cm) level, and that may soon approach 5 ps (1 mm). These measurements are fit using the method of leastsquares to a theoretical model for the lunar motion that takes into account perturbations due to the Sun and the other planets, tidal interactions, and postNewtonian gravitational effects. The predicted roundtrip travel times between retroreflector and telescope also take into account the librations of the Moon, the orientation of the Earth, the location of the observatories, and atmospheric effects on the signal propagation. The “Nordtvedt” parameter η_{N} along with several other important parameters of the model are then estimated in the leastsquares method.
Numerous ongoing analyses of the data find no evidence, within experimental uncertainty, for the Nordtvedt effect [295, 296] (for earlier results see [95, 294, 192]). These results represent a limit on a possible violation of WEP for massive bodies of about 1.4 parts in 10^{13} (compare Figure 1).
However, at this level of precision, one cannot regard the results of LLR as a “clean” test of SEP until one eliminates the possibility of a compensating violation of WEP for the two bodies, because the chemical compositions of the Earth and Moon differ in the relative fractions of iron and silicates. To this end, the EötWash group carried out an improved test of WEP for laboratory bodies whose chemical compositions mimic that of the Earth and Moon. The resulting bound of 1.4 parts in 10^{13} [19, 2] from composition effects reduces the ambiguity in the LLR bound, and establishes the firm SEP test at the level of about 2 parts in 10^{13}. These results can be summarized by the Nordtvedt parameter bound η_{N} = (4.4 ± 4.5) × 10^{−4}.
In the future, the Apache Point Observatory for Lunar Laser ranging Operation (APOLLO) project, a joint effort by researchers from the Universities of Washington, Seattle, and California, San Diego, plans to use enhanced laser and telescope technology, together with a good, highaltitude site in New Mexico, to improve the LLR bound by as much as an order of magnitude [296].
In GR, the Nordtvedt effect vanishes; at the level of several centimeters and below, a number of nonnull general relativistic effects should be present [201].
Tests of the Nordtvedt effect for neutron stars have also been carried out using a class of systems known as wideorbit binary millisecond pulsars (WBMSP), which are pulsarwhitedwarf binary systems with small orbital eccentricities. In the gravitational field of the galaxy, a nonzero Nordtvedt effect can induce an apparent anomalous eccentricity pointed toward the galactic center [86], which can be bounded using statistical methods, given enough WBMSPs (see [243] for a review and references). Using data from 21 WBMSPs, including recently discovered highly circular systems, Stairs et al. [244] obtained the bound Δ < 5.6 × 10^{−3}, where Δ = η_{N}(E_{g}/M)_{NS}. Because (E_{g}/M)_{NS} ∼ 0.1 for typical neutron stars, this bound does not compete with the bound on η_{N} from LLR; on the other hand, it does test SEP in the strongfield regime because of the presence of the neutron stars.
Preferredframe and preferredlocation effects
Some theories of gravity violate SEP by predicting that the outcomes of local gravitational experiments may depend on the velocity of the laboratory relative to the mean rest frame of the universe (preferredframe effects) or on the location of the laboratory relative to a nearby gravitating body (preferredlocation effects). In the postNewtonian limit, preferredframe effects are governed by the values of the PPN parameters α_{1}, α_{2}, and α_{3}, and some preferredlocation effects are governed by ξ (see Table 2).
The most important such effects are variations and anisotropies in the locallymeasured value of the gravitational constant which lead to anomalous Earth tides and variations in the Earth’s rotation rate, anomalous contributions to the orbital dynamics of planets and the Moon, selfaccelerations of pulsars, and anomalous torques on the Sun that would cause its spin axis to be randomly oriented relative to the ecliptic (see TEGP 8.2, 8.3, 9.3, and 14.3 (c) [281]). An bound on α_{3} of 4 × 10^{−20} from the period derivatives of 21 millisecond pulsars was reported in [26, 244]; improved bounds on α_{1} were achieved using LLR data [191], and using observations of the circular binary orbit of the pulsar J2317+1439 [25]. Negative searches for these effects have produced strong constraints on the PPN parameters (see Table 4).
Constancy of the Newtonian gravitational constant
Most theories of gravity that violate SEP predict that the locally measured Newtonian gravitational constant may vary with time as the universe evolves. For the scalartensor theories listed in Table 3, the predictions for Ġ/G can be written in terms of time derivatives of the asymptotic scalar field. Where G does change with cosmic evolution, its rate of variation should be of the order of the expansion rate of the universe, i.e. Ġ/G ∼ H_{0}, where H_{0} is the Hubble expansion parameter and is given by H_{0} = 100 h km s^{−1} Mpc^{−1} = 1.02 × 10^{−10} h yr^{−1}, where current observations of the expansion of the universe give h ≈ 0.73 ± 0.03.
Several observational constraints can be placed on Ġ/G, one kind coming from bounding the present rate of variation, another from bounding a difference between the present value and a past value. The first type of bound typically comes from LLR measurements, planetary radarranging measurements, and pulsar timing data. The second type comes from studies of the evolution of the Sun, stars and the Earth, bigbang nucleosynthesis, and analyses of ancient eclipse data. Recent results are shown in Table 5.
The best limits on a current Ġ/G come from LLR measurements (for earlier results see [95, 294, 192]). These have largely supplanted earlier bounds from ranging to the 1976 Viking landers (see TEGP, 14.3 (c) [281]), which were limited by uncertain knowledge of the masses and orbits of asteroids. However, improvements in knowledge of the asteroid belt, combined with continuing radar observations of planets and spacecraft, notably the Mars Global Surveyor (1998–2003) and Mars Odyssey (2002–present), may enable a bound on Ġ/G at the level of a part in 10^{13} per year. For an initial analysis along these lines, see [212]. It has been suggested that radar observations of the planned 2012 BepiColombo Mercury orbiter mission over a twoyear integration with 6 cm rms accuracy in range could yield Δ(Ġ/G) < 10^{−13} yr^{−1}; an eightyear mission could improve this by a factor 15 [187, 17].
Although bounds on Ġ/G from solarsystem measurements can be correctly obtained in a phenomenological manner through the simple expedient of replacing G by G_{0} + Ġ_{0}(t − t_{0}) in Newton’s equations of motion, the same does not hold true for pulsar and binary pulsar timing measurements. The reason is that, in theories of gravity that violate SEP, such as scalartensor theories, the “mass” and moment of inertia of a gravitationally bound body may vary with variation in G. Because neutron stars are highly relativistic, the fractional variation in these quantities can be comparable to ΔG/G, the precise variation depending both on the equation of state of neutron star matter and on the theory of gravity in the strongfield regime. The variation in the moment of inertia affects the spin rate of the pulsar, while the variation in the mass can affect the orbital period in a manner that can subtract from the direct effect of a variation in G, given by Ṗ_{b}/P_{b} = −2Ġ/G [200]. Thus, the bounds quoted in Table 5 for the binary pulsar PSR 1913+16 and others [143] (see also [87]) are theorydependent and must be treated as merely suggestive.
In a similar manner, bounds from helioseismology and bigbang nucleosynthesis (BBN) assume a model for the evolution of G over the multibillion year time spans involved. For example, the concordance of predictions for light elements produced around 3 minutes after the big bang with the abundances observed indicate that G then was within 20 percent of G today. Assuming a powerlaw variation of G ∼ t^{−α} then yields a bound on Ġ/G today shown in Table 5.
Other tests of postNewtonian gravity
Search for gravitomagnetism
According to GR, moving or rotating matter should produce a contribution to the gravitational field that is the analogue of the magnetic field of a moving charge or a magnetic dipole. In particular, one can view the g_{0i} part of the PPN metric (see Box 2) as an analogue of the vector potential of electrodynamics. In a suitable gauge, and dropping the preferredframe terms, it can be written
At PN order, this contributes a Lorentztype acceleration v × B_{g} to the equation of motion, where the gravitomagnetic field B_{g} is given by B_{g} = ∇ × (g_{0i}e^{i}).
Gravitomagnetism plays a role in a variety of measured relativistic effects involving moving material sources, such as the EarthMoon system and binary pulsar systems. Nordtvedt [199, 198] has argued that, if the gravitomagnetic potential (55) were turned off, then there would be anomalous orbital effects in LLR and binary pulsar data.
Rotation also produces a gravitomagnetic effect, since for a rotating body, \({\bf{V}} =  {1 \over 2}{\bf{x}} \times {\bf{J}}/{r^3}\), where J is the angular momentum of the body. The result is a “dragging of inertial frames” around the body, also called the LenseThirring effect. A consequence is a precession of a gyroscope’s spin S according to
where n is a unit radial vector, and r is the distance from the center of the body (TEGP 9.1 [281]).
The Relativity Gyroscope Experiment (Gravity Probe B or GPB) at Stanford University, in collaboration with NASA and LockheedMartin Corporation [246], recently completed a space mission to detect this framedragging or LenseThirring precession, along with the “geodetic” precession (see Section 3.7.2). A set of four superconductingniobiumcoated, spherical quartz gyroscopes were flown in a polar Earth orbit (642 km mean altitude, 0.0014 eccentricity), and the precessions of the gyroscopes relative to a distant guide star (HR 8703, IM Pegasi) were measured. For the given orbit, the predicted secular angular precession of the gyroscopes is in a direction perpendicular to the orbital plane at a rate \({1 \over 2}(1 + \gamma + {1 \over 4}{\alpha _1}) \times 41 \times {10^{ 3}}\) arcsec yr^{−1}. The accuracy goal of the experiment is about 0.5 milliarcseconds per year. The spacecraft was launched on April 20, 2004, and the mission ended in September 2005, as scheduled, when the remaining liquid helium boiled off.
It is too early to know whether the relativistic precessions were measured in the amount predicted by GR, because an important calibration of the instrument exploits the effect of the aberration of starlight on the pointing of the onboard telescope toward the guide star, and completing this calibration required the full mission data set. In addition, part of the measured effect includes the motion of the guide star relative to distant inertial frames. This was measured before, during and after the mission separately by radio astronomers at Harvard/SAO and elsewhere using VLBI, and the results of those measurements were to be strictly embargoed until the GPB team has completed its analysis of the gyro data. Final results from the experiment are expected in 2006.
Another way to look for framedragging is to measure the precession of orbital planes of bodies circling a rotating body. One implementation of this idea is to measure the relative precession, at about 31 milliarcseconds per year, of the line of nodes of a pair of laserranged geodynamics satellites (LAGEOS), ideally with supplementary inclination angles; the inclinations must be supplementary in order to cancel the dominant (126 degrees per year) nodal precession caused by the Earth’s Newtonian gravitational multipole moments. Unfortunately, the two existing LAGEOS satellites are not in appropriately inclined orbits, and no concrete plans exist at present to launch a third satellite in a supplementary orbit. Nevertheless, Ciufolini and Pavlis [56] combined nodal precession data from LAGEOS I and II with improved models for the Earth’s multipole moments provided by two recent orbiting geodesy satellites, Europe’s CHAMP (Challenging Minisatellite Payload) and NASA’s GRACE (Gravity Recovery and Climate Experiment), and reported a 5–10 percent confirmation of GR. In earlier reports, Ciufolini et al. had reported tests at the the 20–30 percent level, without the benefit of the GRACE/CHAMP data [55, 57, 54]. Some authors stressed the importance of adequately assessing systematic errors in the LAGEOS data [226, 133].
Geodetic precession
A gyroscope moving through curved spacetime suffers a precession of its spin axis given by
where v is the velocity of the gyroscope, and U is the Newtonian gravitational potential of the source (TEGP 9.1 [281]). The EarthMoon system can be considered as a “gyroscope”, with its axis perpendicular to the orbital plane. The predicted precession is about 2 arcseconds per century, an effect first calculated by de Sitter. This effect has been measured to about 0.6 percent using LLR data [95, 294, 295].
For the GPB gyroscopes orbiting the Earth, the precession is 6.6 arcseconds per year. A goal of GPB is to measure this effect to 8 × 10^{−5}; if achieved, this could bound the parameter γ to a part in 10^{4}, not competitive with the Cassini bound.
Tests of postNewtonian conservation laws
Of the five “conservation law” PPN parameters ζ_{1}, ζ_{2}, ζ_{3}, ζ_{4}, and α_{3}, only three, ζ_{2}, ζ_{3}, and α_{3}, have been constrained directly with any precision; ζ_{1} is constrained indirectly through its appearance in the Nordtvedt effect parameter η_{N}, Equation (53). There is strong theoretical evidence that ζ_{4}, which is related to the gravity generated by fluid pressure, is not really an independent parameter — in any reasonable theory of gravity there should be a connection between the gravity produced by kinetic energy (ρυ^{2}), internal energy (ρΠ), and pressure (p). From such considerations, there follows [275] the additional theoretical constraint
A nonzero value for any of these parameters would result in a violation of conservation of momentum, or of Newton’s third law in gravitating systems. An alternative statement of Newton’s third law for gravitating systems is that the “active gravitational mass”, that is the mass that determines the gravitational potential exhibited by a body, should equal the “passive gravitational mass”, the mass that determines the force on a body in a gravitational field. Such an equality guarantees the equality of action and reaction and of conservation of momentum, at least in the Newtonian limit.
A classic test of Newton’s third law for gravitating systems was carried out in 1968 by Kreuzer, in which the gravitational attraction of fluorine and bromine were compared to a precision of 5 parts in 10^{5}.
A remarkable planetary test was reported by Bartlett and van Buren [22]. They noted that current understanding of the structure of the Moon involves an ironrich, aluminumpoor mantle whose center of mass is offset about 10 km from the center of mass of an aluminumrich, ironpoor crust. The direction of offset is toward the Earth, about 14° to the east of the EarthMoon line. Such a model accounts for the basaltic maria which face the Earth, and the aluminumrich highlands on the Moon’s far side, and for a 2 km offset between the observed center of mass and center of figure for the Moon. Because of this asymmetry, a violation of Newton’s third law for aluminum and iron would result in a momentum nonconserving selfforce on the Moon, whose component along the orbital direction would contribute to the secular acceleration of the lunar orbit. Improved knowledge of the lunar orbit through LLR, and a better understanding of tidal effects in the EarthMoon system (which also contribute to the secular acceleration) through satellite data, severely limit any anomalous secular acceleration, with the resulting limit
According to the PPN formalism, in a theory of gravity that violates conservation of momentum, but that obeys the constraint of Equation (58), the electrostatic binding energy E_{e} of an atomic nucleus could make a contribution to the ratio of active to passive mass of the form
The resulting limit on ζ_{3} from the lunar experiment is ζ_{3} < 1 × 10^{−8} (TEGP 9.2, 14.3 (d) [281]). Nordtvedt [203] has examined whether this bound could be improved by considering the asymmetric distribution of ocean water on Earth.
Another consequence of a violation of conservation of momentum is a selfacceleration of the center of mass of a binary stellar system, given by
where δm = m_{1} − m_{2}, a is the semimajor axis, and n_{P} is a unit vector directed from the center of mass to the point of periastron of m_{1} (TEGP 9.3 [281]). A consequence of this acceleration would be nonvanishing values for d^{2}P/dt^{2}, where P denotes the period of any intrinsic process in the system (orbit, spectra, pulsar periods). The observed upper limit on d^{2}P_{p}/dt^{2} of the binary pulsar PSR 1913+16 places a strong constraint on such an effect, resulting in the bound α_{3} + ζ_{2} < 4 × 10^{−5}. Since α_{3} has already been constrained to be much less than this (see Table 4), we obtain a strong solitary bound on ζ_{2} < 4 × 10^{−5} [280].
Prospects for improved PPN parameter values
A number of advanced experiments or space missions are under development or have been proposed which could lead to significant improvements in values of the PPN parameters, of J_{2} of the Sun, and of Ġ/G.
LLR at the Apache Point Observatory (APOLLO project) could improve bounds on the Nordvedt parameter to the level 3 × 10^{−5} and on Ġ/G to better than 10^{−13} yr^{−1} [296].
The proposed 2012 ESA BepiColumbo Mercury orbiter, in a twoyear experiment, with 6 cm range capability, could yield improvements in γ to 3 × 10^{−5}, in β to 3 × 10^{−4}, in α_{1} to 10^{−5}, in Ġ/G to 10^{−13} yr^{−1}, and in J_{2} to 3 × 10^{−8}. An eightyear mission could yield further improvements by factors of 2–5 in β, α_{1}, and J_{2}, and a further factor 15 in Ġ/G [187, 17].
GAIA is a highprecision astrometric orbiting telescope (a successor to Hipparcos), which could measure lightdeflection and γ to the 10^{−6} level [101]. It is planned for launch by ESA in the 2011 time frame.
LATOR (Laser Astrometric Test of Relativity) is a concept for a NASA mission in which two microsatellites orbit the Sun on Earthlike orbits near superior conjunction, so that their lines of sight are close to the Sun. Using optical tracking and an optical interferometer on the International Space Station, it may be possible to measure the deflection of light with sufficient accuracy to bound γ to a part in 10^{8} and J_{2} to a part in 10^{8}, and to measure the solar framedragging effect to one percent [259, 260].
Nordtvedt [202] has argued that “grand fits” of large solar system ranging data sets, including radar ranging to Mercury, Mars, and satellites, and laser ranging to the Moon, could yield substantially improved measurements of PPN parameters. A recent contribution in that direction is [212].
Strong Gravity and Gravitational Waves: A New Testing Ground
Strongfield systems in general relativity
Defining weak and strong gravity
In the solar system, gravity is weak, in the sense that the Newtonian gravitational potential and related variables (U(x, t) ∼ υ^{2} ∼ p/ρ ∼ ϵ) are much smaller than unity everywhere. This is the basis for the postNewtonian expansion and for the “parametrized postNewtonian” framework described in Section 3.2. “Strongfield” systems are those for which the simple 1PN approximation of the PPN framework is no longer appropriate. This can occur in a number of situations:

The system may contain strongly relativistic objects, such as neutron stars or black holes, near and inside which ϵ ∼ 1, and the postNewtonian approximation breaks down. Nevertheless, under some circumstances, the orbital motion may be such that the interbody potential and orbital velocities still satisfy ϵ ≪ 1 so that a kind of postNewtonian approximation for the orbital motion might work; however, the strongfield internal gravity of the bodies could (especially in alternative theories of gravity) leave imprints on the orbital motion.

The evolution of the system may be affected by the emission of gravitational radiation. The 1PN approximation does not contain the effects of gravitational radiation backreaction. In the expression for the metric given in Box 2, radiation backreaction effects do not occur until \({\mathcal O}({\epsilon ^{7/2}})\) in g_{00}, \({\mathcal O}({\epsilon ^3})\) in \({g_{{0_i}}}\), and \({\mathcal O}({\varepsilon ^{5/2}})\) in g_{ij}. Consequently, in order to describe such systems, one must carry out a solution of the equations substantially beyond 1PN order, sufficient to incorporate the leading radiation damping terms at 2.5PN order. In addition, the PPN metric described in Section 3.2 is valid in the near zone of the system, i.e. within one gravitational wavelength of the system’s center of mass. As such it cannot describe the gravitational waves seen by a detector.

The system may be highly relativistic in its orbital motion, so that U ∼ υ^{2} ∼ 1 even for the interbody field and orbital velocity. Systems like this include the late stage of the inspiral of binary systems of neutron stars or black holes, driven by gravitational radiation damping, prior to a merger and collapse to a final stationary state. Binary inspiral is one of the leading candidate sources for detection by a worldwide network of laser interferometric gravitational wave observatories nearing completion. A proper description of such systems requires not only equations for the motion of the binary carried to extraordinarily high PN orders (at least 3.5PN), but also requires equations for the farzone gravitational waveform measured at the detector, that are equally accurate to high PN orders beyond the leading “quadrupole” approximation.
Of course, some systems cannot be properly described by any postNewtonian approximation because their behavior is fundamentally controlled by strong gravity. These include the imploding cores of supernovae, the final merger of two compact objects, the quasinormalmode vibrations of neutron stars and black holes, the structure of rapidly rotating neutron stars, and so on. Phenomena such as these must be analyzed using different techniques. Chief among these is the full solution of Einstein’s equations via numerical methods. This field of “numerical relativity” is a rapidly growing and maturing branch of gravitational physics, whose description is beyond the scope of this review (see [165, 24] for reviews). Another is black hole perturbation theory (see [188, 146, 235] for reviews).
Compact bodies and the strong equivalence principle
When dealing with the motion and gravitational wave generation by orbiting bodies, one finds a remarkable simplification within GR. As long as the bodies are sufficiently wellseparated that one can ignore tidal interactions and other effects that depend upon the finite extent of the bodies (such as their quadrupole and higher multipole moments), then all aspects of their orbital behavior and gravitational wave generation can be characterized by just two parameters: mass and angular momentum. Whether their internal structure is highly relativistic, as in black holes or neutron stars, or nonrelativistic as in the Earth and Sun, only the mass and angular momentum are needed. Furthermore, both quantities are measurable in principle by examining the external gravitational field of the bodies, and make no reference whatsoever to their interiors.
Damour [70] calls this the “effacement” of the bodies’ internal structure. It is a consequence of the SEP, described in Section 3.1.2.
General relativity satisfies SEP because it contains one and only one gravitational field, the spacetime metric g_{μν}. Consider the motion of a body in a binary system, whose size is small compared to the binary separation. Surround the body by a region that is large compared to the size of the body, yet small compared to the separation. Because of the general covariance of the theory, one can choose a freelyfalling coordinate system which comoves with the body, whose spacetime metric takes the Minkowski form at its outer boundary (ignoring tidal effects generated by the companion). There is thus no evidence of the presence of the companion body, and the structure of the chosen body can be obtained using the field equations of GR in this coordinate system. Far from the chosen body, the metric is characterized by the mass and angular momentum (assuming that one ignores quadrupole and higher multipole moments of the body) as measured far from the body using orbiting test particles and gyroscopes. These asymptotically measured quantities are oblivious to the body’s internal structure. A black hole of mass m and a planet of mass m would produce identical spacetimes in this outer region.
The geometry of this region surrounding the one body must be matched to the geometry provided by the companion body. Einstein’s equations provide consistency conditions for this matching that yield constraints on the motion of the bodies. These are the equations of motion. As a result the motion of two planets of mass and angular momentum m_{1}, m_{2}, J_{1}, and J_{2} is identical to that of two black holes of the same mass and angular momentum (again, ignoring tidal effects).
This effacement does not occur in an alternative gravitional theory like scalartensor gravity. There, in addition to the spacetime metric, a scalar field ϕ is generated by the masses of the bodies, and controls the local value of the gravitational coupling constant (i.e. G is a function of ϕ). Now, in the local frame surrounding one of the bodies in our binary system, while the metric can still be made Minkowskian far away, the scalar field will take on a value ϕ_{0} determined by the companion body. This can affect the value of G inside the chosen body, alter its internal structure (specifically its gravitational binding energy) and hence alter its mass. Effectively, each body can be characterized by several mass functions m_{A}(ϕ), which depend on the value of the scalar field at its location, and several distinct masses come into play, such as inertial mass, gravitational mass, “radiation” mass, etc. The precise nature of the functions will depend on the body, specifically on its gravitational binding energy, and as a result, the motion and gravitational radiation may depend on the internal structure of each body. For compact bodies such as neutron stars and black holes these internal structure effects could be large; for example, the gravitational binding energy of a neutron star can be 10–20 percent of its total mass. At 1PN order, the leading manifestation of this phenomenon is the Nordtvedt effect.
This is how the study of orbiting systems containing compact objects provides strongfield tests of GR. Even though the strongfield nature of the bodies is effaced in GR, it is not in other theories, thus any result in agreement with the predictions of GR constitutes a kind of “null” test of strongfield gravity.
Motion and gravitational radiation in general relativity
The motion of bodies and the generation of gravitational radiation are longstanding problems that date back to the first years following the publication of GR, when Einstein calculated the gravitational radiation emitted by a laboratoryscale object using the linearized version of GR, and de Sitter calculated Nbody equations of motion for bodies in the 1PN approximation to GR. It has at times been controversial, with disputes over such issues as whether Einstein’s equations alone imply equations of motion for bodies (Einstein, Infeld, and Hoffman demonstrated explicitly that they do, using a matching procedure similar to the one described above), whether gravitational waves are real or are artifacts of general covariance (Einstein waffled; Bondi and colleagues proved their reality rigorously in the 1950s), and even over algebraic errors (Einstein erred by a factor of 2 in his first radiation calculation; Eddington found the mistake). Shortly after the discovery of the binary pulsar PSR 1913+16 in 1974, questions were raised about the foundations of the “quadrupole formula” for gravitational radiation damping (and in some quarters, even about its quantitative validity). These questions were answered in part by theoretical work designed to shore up the foundations of the quadrupole approximation, and in part (perhaps mostly) by the agreement between the predictions of the quadrupole formula and the observed rate of damping of the pulsar’s orbit (see Section 5.1). Damour [70] gives a thorough historical and technical review of this subject up to 1986.
The problem of motion and radiation in GR has received renewed interest since 1990, with proposals for construction of largescale laser interferometric gravitational wave observatories, such as the LIGO project in the US, VIRGO and GEO600 in Europe, and TAMA300 in Japan, and the realization that a leading candidate source of detectable waves would be the inspiral, driven by gravitational radiation damping, of a binary system of compact objects (neutron stars or black holes) [1, 256]. The analysis of signals from such systems will require theoretical predictions from GR that are extremely accurate, well beyond the leadingorder prediction of Newtonian or even postNewtonian gravity for the orbits, and well beyond the leadingorder formulae for gravitational waves.
This presented a major theoretical challenge: to calculate the motion and radiation of systems of compact objects to very high PN order, a formidable algebraic task, while addressing a number of issues of principle that have historically plagued this subject, sufficiently well to ensure that the results were physically meaningful. This challenge has been largely met, so that we may soon see a remarkable convergence between observational data and accurate predictions of gravitational theory that could provide new, strongfield tests of GR.
Here we give a brief overview of the problem of motion and gravitational radiation in GR.
Einstein’s equations in “relaxed” form
The Einstein equations G_{μν} = 8πT_{μν} are elegant and deceptively simple, showing geometry (in the form of the Einstein tensor G_{μν}, which is a function of spacetime curvature) being generated by matter (in the form of the material stressenergy tensor T_{μν}). However, this is not the most useful form for actual calculations. For postNewtonian calculations, a far more useful form is the socalled “relaxed” Einstein equations:
where □ ≡ −∂^{2}/∂t^{2} + ∇^{2} is the flatspacetime wave operator, h^{αβ} is a “gravitational tensor potential” related to the deviation of the spacetime metric from its Minkowski form by the formula h^{αβ} ≡ η^{αβ} − (−g)^{1/2} g^{αβ}, g is the determinant of g_{αβ}, and a particular coordinate system has been specified by the de Donder or harmonic gauge condition ∂h^{αβ}/∂x^{β} = 0 (summation on repeated indices is assumed). This form of Einstein’s equations bears a striking similarity to Maxwell’s equations for the vector potential A^{α} in Lorentz gauge: □A^{α} = −4πJ^{α}, ∂A^{α}/∂x^{α} = 0. There is a key difference, however: The source on the right hand side of Equation (62) is given by the “effective” stressenergy pseudotensor
where Λ^{αβ} is the nonlinear “field” contribution given by terms quadratic (and higher) in h^{αβ} and its derivatives (see [189], Eqs. (20.20, 20.21) for formulae). In GR, the gravitational field itself generates gravity, a reflection of the nonlinearity of Einstein’s equations, and in contrast to the linearity of Maxwell’s equations.
Equation (62) is exact, and depends only on the assumption that spacetime can be covered by harmonic coordinates. It is called “relaxed” because it can be solved formally as a functional of source variables without specifying the motion of the source, in the form
where the integration is over the past flatspacetime null cone \({\mathcal C}\) of the field point (t, x). The motion of the source is then determined either by the equation ∂τ^{αβ}/∂x^{β} = 0 (which follows from the harmonic gauge condition), or from the usual covariant equation of motion \({T^{\alpha \beta}}_{;\beta} = 0\), where the subscript; β denotes a covariant divergence. This formal solution can then be iterated in a slow motion (υ < 1) weakfield (∥h^{αβ}∥ < 1) approximation. One begins by substituting \(h_0^{\alpha \beta} = 0\) into the source τ^{αβ} in Equation (64), and solving for the first iterate \(h_1^{\alpha \beta}\), and then repeating the procedure sufficiently many times to achieve a solution of the desired accuracy. For example, to obtain the 1PN equations of motion, two iterations are needed (i.e. \(h_2^{\alpha \beta}\) must be calculated); likewise, to obtain the leading gravitational waveform for a binary system, two iterations are needed.
At the same time, just as in electromagnetism, the formal integral (64) must be handled differently, depending on whether the field point is in the far zone or the near zone. For field points in the far zone or radiation zone, x > ƛ > x′ (ƛ is the gravitational wavelength divided by 2π), the field can be expanded in inverse powers of R = x in a multipole expansion, evaluated at the “retarded time” t − R. The leading term in 1/R is the gravitational waveform. For field points in the near zone or induction zone, x ∼ x′ < ƛ, the field is expanded in powers of x − x′ about the local time t, yielding instantaneous potentials that go into the equations of motion.
However, because the source τ^{αβ} contains h^{αβ} itself, it is not confined to a compact region, but extends over all spacetime. As a result, there is a danger that the integrals involved in the various expansions will diverge or be illdefined. This consequence of the nonlinearity of Einstein’s equations has bedeviled the subject of gravitational radiation for decades. Numerous approaches have been developed to try to handle this difficulty. The “postMinkowskian” method of Blanchet, Damour, and Iyer [35, 36, 37, 76, 38, 33] solves Einstein’s equations by two different techniques, one in the near zone and one in the far zone, and uses the method of singular asymptotic matching to join the solutions in an overlap region. The method provides a natural “regularization” technique to control potentially divergent integrals (see [34] for a thorough review). The “Direct Integration of the Relaxed Einstein Equations” (DIRE) approach of Will, Wiseman, and Pati [291, 208] retains Equation (64) as the global solution, but splits the integration into one over the near zone and another over the far zone, and uses different integration variables to carry out the explicit integrals over the two zones. In the DIRE method, all integrals are finite and convergent. Itoh and Futamase have used an extension of the EinsteinInfeldHoffman matching approach combined with a specific method for taking a pointparticle limit [134], while Damour, Jaranowski, and Schäfer have pioneered an ADM Hamiltonian approach that focuses on the equations of motion [139, 140, 77, 79, 78].
These methods assume from the outset that gravity is sufficiently weak that ∥h^{αβ}∥ < 1 and harmonic coordinates exists everywhere, including inside the bodies. Thus, in order to apply the results to cases where the bodies may be neutron stars or black holes, one relies upon the SEP to argue that, if tidal forces are ignored, and equations are expressed in terms of masses and spins, one can simply extrapolate the results unchanged to the situation where the bodies are ultrarelativistic. While no general proof of this exists, it has been shown to be valid in specific circumstances, such as at 2PN order in the equations of motion, and for black holes moving in a Newtonian background field [70].
Methods such as these have resolved most of the issues that led to criticism of the foundations of gravitational radiation theory during the 1970s.
Equations of motion and gravitational waveform
Among the results of these approaches are formulae for the equations of motion and gravitational waveform of binary systems of compact objects, carried out to high orders in a PN expansion. Here we shall only state the key formulae that will be needed for this review. For example, the relative twobody equation of motion has the form
where m = m_{1} + m_{2} is the total mass, r = x_{1} − x_{2}, v = v_{1} − v_{2}, and \({{\bf{A}}_{n{\rm{PN}}}}\). The notation A_{nPN} indicates that the term is \({\mathcal O}({\epsilon ^n})\) relative to the Newtonian term \( {\rm{\hat n}}\). Explicit and unambiguous formulae for nonspinning bodies through 3.5PN order have been calculated by various authors (see [34] for a review). Here we quote only the first PN corrections and the leading radiationreaction terms at 2.5PN order:
where η = m_{1}m_{2}/(m_{1} + m_{2})^{2}. These terms are sufficient to analyze the orbit and evolution of the binary pulsar (see Section 5.1). For example, the 1PN terms are responsible for the periastron advance of an eccentric orbit, given by \(\dot \omega = 6\pi {f_{\rm{b}}}m/a(1  {e^2})\), where a and e are the semimajor axis and eccentricity of the orbit, respectively, and f_{b} is the orbital frequency, given to the needed order by Kepler’s third law 2πf_{b} = (m/a^{3})^{1/2}.
Another product is a formula for the gravitational field far from the system, written schematically in the form
where R is the distance from the source, and the variables are to be evaluated at retarded time t − R. The leading term is the socalled quadrupole formula
where I^{ij} is the quadrupole moment of the source, and overdots denote time derivatives. For a binary system this leads to
For binary systems, explicit formulae for the waveform through 2PN order have been derived (see [40] for a readytouse presentation of the waveform for circular orbits; see [34] for a full review).
Given the gravitational waveform, one can compute the rate at which energy is carried off by the radiation (schematically ∫ ḣḣ dΩ, the gravitational analog of the Poynting flux). The lowestorder quadrupole formula leads to the gravitational wave energy flux
This has been extended to 3.5PN order beyond the quadrupole formula (see [34] for a review). Formulae for fluxes of angular and linear momentum can also be derived. The 2.5PN radiationreaction terms in the equation of motion (65) result in a damping of the orbital energy that precisely balances the energy flux (71) determined from the waveform. Averaged over one orbit, this results in a rate of increase of the binary’s orbital frequency given by
where \({\mathcal M}\) is the socalled “chirp” mass, given by \({\mathcal M} = {\eta ^{3/5}}m\). Notice that by making precise measurements of the phase Φ(t) = 2π ∫^{t} f (t′) dt′ of either the orbit or the gravitational waves (for which f = 2f_{b} for the dominant component) as a function of the frequency, one in effect measures the “chirp” mass of the system.
These formalisms have also been generalized to include the leading effects of spinorbit and spinspin coupling between the bodies [145, 144, 289].
Another approach to gravitational radiation is applicable to the special limit in which one mass is much smaller than the other. This is the method of black hole perturbation theory. One begins with an exact background spacetime of a black hole, either the nonrotating Schwarzschild or the rotating Kerr solution, and perturbs it according to \({g_{\mu \nu}} = g_{\mu \nu}^{(0)} + {h_{\mu \nu}}\). The particle moves on a geodesic of the background spacetime, and a suitably defined source stressenergy tensor for the particle acts as a source for the gravitational perturbation and wave field h_{μν}. This method provides numerical results that are exact in υ, as well as analytical results expressed as series in powers of υ, both for nonrotating and for rotating black holes. For nonrotating holes, the analytical expansions have been carried to 5.5PN order, or ϵ^{5.5} beyond the quadrupole approximation. All results of black hole perturbation agree precisely with the m_{1} → 0 limit of the PN results, up to the highest PN order where they can be compared (for reviews see [188, 235]).
Gravitational wave detection
A gravitational wave detector can be modelled as a body of mass M at a distance L from a fiducial laboratory point, connected to the point by a spring of resonant frequency ω_{0} and quality factor Q. From the equation of geodesic deviation, the infinitesimal displacement ξ of the mass along the line of separation from its equilibrium position satisfies the equation of motion
whereF_{+}(θ,ϕ,ψ) and F_{×} (θ,ϕ,ψ) are “beampattern” factors that depend on the direction of the source (θ,ϕ) and on a polarization angle ψ, and h_{+}(t) and h_{×}(t) are gravitational waveforms corresponding to the two polarizations of the gravitational wave (for a review, see [255]). In a source coordinate system in which the x − y plane is the plane of the sky and the zdirection points toward the detector, these two modes are given by
where \(h_{{\rm{TT}}}^{ij}\) represent transversetraceless (TT) projections of the calculated waveform of Equation (68), given by
where \({{\hat N}^j}\) is a unit vector pointing toward the detector. The beam pattern factors depend on the orientation and nature of the detector. For a wave approaching along the laboratory zdirection, and for a mass whose location on the x − y plane makes an angle ϕ with the x axis, the beam pattern factors are given by F_{+} = cos 2ϕ and F× = sin 2ϕ. For a resonant cylinder oriented along the laboratory z axis, and for source direction (θ, ϕ), they are given by F_{+} = sin^{2} θ cos 2ψ, F_{×} = sin^{2} θ sin 2ψ (the angle ψ measures the relative orientation of the laboratory and source xaxes). For a laser interferometer with one arm along the laboratory xaxis, the other along the yaxis, and with ξ defined as the differential displacement along the two arms, the beam pattern functions are \({F_ +} = {1 \over 2}(1 + {\cos ^2}\theta)\cos 2\phi \cos 2\psi  \cos \theta \sin 2\phi \sin 2\psi\) and \({F_ \times} = {1 \over 2}(1 + {\cos ^2}\theta)\cos 2\phi \sin 2\psi + \cos \theta \sin 2\phi \cos 2\psi\).
The waveforms h_{+}(t) and h_{×} (t) depend on the nature and evolution of the source. For example, for a binary system in a circular orbit, with an inclination i relative to the plane of the sky, and the xaxis oriented along the major axis of the projected orbit, the quadrupole approximation of Equation (70) gives
where Φ_{b}(t) = 2π ∫^{t} f_{b}(t′) dt′ is the orbital phase.
Stellar System Tests of Gravitational Theory
The binary pulsar and general relativity
The 1974 discovery of the binary pulsar B1913+16 by Joseph Taylor and Russell Hulse during a routine search for new pulsars provided the first possibility of probing new aspects of gravitational theory: the effects of strong relativistic internal gravitational fields on orbital dynamics, and the effects of gravitational radiation reaction. For reviews of the discovery, see the published Nobel Prize lectures by Hulse and Taylor [132, 252]. For reviews of the current status of pulsars, including binary and millisecond pulsars, see [173, 243].
The system consists of a pulsar of nominal period 59 ms in a close binary orbit with an as yet unseen companion. The orbital period is about 7.75 hours, and the eccentricity is 0.617. From detailed analyses of the arrival times of pulses (which amounts to an integrated version of the Dopplershift methods used in spectroscopic binary systems), extremely accurate orbital and physical parameters for the system have been obtained (see Table 6). Because the orbit is so close (≈ 1R_{⊙}) and because there is no evidence of an eclipse of the pulsar signal or of mass transfer from the companion, it is generally agreed that the companion is compact. Evolutionary arguments suggest that it is most likely a dead pulsar, while B1913+16 is a “recycled” pulsar. Thus the orbital motion is very clean, free from tidal or other complicating effects. Furthermore, the data acquisition is “clean” in the sense that by exploiting the intrinsic stability of the pulsar clock combined with the ability to maintain and transfer atomic time accurately using GPS, the observers can keep track of pulse timeofarrival with an accuracy of 13 µs, despite extended gaps between observing sessions (including a severalyear gap in the middle 1990s for an upgrade of the Arecibo radio telescope). The pulsar has shown no evidence of “glitches” in its pulse period.
Three factors make this system an arena where relativistic celestial mechanics must be used: the relatively large size of relativistic effects [υ_{orbit} ≈ (m/r)^{1/2} ≈ 10^{−3}], a factor of 10 larger than the corresponding values for solarsystem orbits; the short orbital period, allowing secular effects to build up rapidly; and the cleanliness of the system, allowing accurate determinations of small effects. Because the orbital separation is large compared to the neutron stars’ compact size, tidal effects can be ignored. Just as Newtonian gravity is used as a tool for measuring astrophysical parameters of ordinary binary systems, so GR is used as a tool for measuring astrophysical parameters in the binary pulsar.
The observational parameters that are obtained from a leastsquares solution of the arrivaltime data fall into three groups:

1.
nonorbital parameters, such as the pulsar period and its rate of change (defined at a given epoch), and the position of the pulsar on the sky;

2.
five “Keplerian” parameters, most closely related to those appropriate for standard Newtonian binary systems, such as the eccentricity e, the orbital period P_{b}, and the semimajor axis of the pulsar projected along the line of sight, a_{p} sin i; and

3.
five “postKeplerian” parameters.
The five postKeplerian parameters are: \(\langle \dot \omega \rangle\), the average rate of periastron advance; γ′, the amplitude of delays in arrival of pulses caused by the varying effects of the gravitational redshift and time dilation as the pulsar moves in its elliptical orbit at varying distances from the companion and with varying speeds; Ṗ_{b}, the rate of change of orbital period, caused predominantly by gravitational radiation damping; and r and s = sin i, respectively the “range” and “shape” of the Shapiro time delay of the pulsar signal as it propagates through the curved spacetime region near the companion, where i is the angle of inclination of the orbit relative to the plane of the sky. An additional 14 relativistic parameters are measurable in principle [88].
In GR, the five postKeplerian parameters can be related to the masses of the two bodies and to measured Keplerian parameters by the equations (TEGP 12.1, 14.6 (a) [281])
where m_{1} and m_{2} denote the pulsar and companion masses, respectively. The formula for \(\langle \dot \omega \rangle\) ignores possible nonrelativistic contributions to the periastron shift, such as tidally or rotationally induced effects caused by the companion (for discussion of these effects, see TEGP 12.1 (c) [281]). The formula for Ṗ_{b} includes only quadrupole gravitational radiation; it ignores other sources of energy loss, such as tidal dissipation (TEGP 12.1 (f) [281]). Notice that, by virtue of Kepler’s third law, (2πf_{b})^{2} = m/a^{3}, (2πmf_{b})^{2/3} = m/a ∼ ϵ, thus the first two postKeplerian parameters can be seen as \({\mathcal O}(\epsilon)\), or 1PN corrections to the underlying variable, while the third is an \({\mathcal O}({\epsilon ^{5/2}})\), or 2.5PN correction. The current observed values for the Keplerian and postKeplerian parameters are shown in Table 6. The parameters r and s are not separately measurable with interesting accuracy for B1913+16 because the orbit’s 47^{°} inclination does not lead to a substantial Shapiro delay.
Because f_{b} and e are separately measured parameters, the measurement of the three postKeplerian parameters provide three constraints on the two unknown masses. The periastron shift measures the total mass of the system, Ṗ_{b} measures the chirp mass, and γ′ measures a complicated function of the masses. GR passes the test if it provides a consistent solution to these constraints, within the measurement errors.
From the intersection of the \(\langle \dot \omega \rangle\) and γ′ constraints we obtain the values m_{1} = 1.4414±0.0002M_{⊙} and m_{2} = 1.3867±0.0002M_{⊙}. The third of Equations (78) then predicts the value Ṗ_{b} = −2.40242± 0.00002 × 10^{−12}. In order to compare the predicted value for Ṗ_{b} with the observed value of Table 6, it is necessary to take into account the small effect of a relative acceleration between the binary pulsar system and the solar system caused by the differential rotation of the galaxy. This effect was previously considered unimportant when Ṗ_{b} was known only to 10 percent accuracy. Damour and Taylor [87] carried out a careful estimate of this effect using data on the location and proper motion of the pulsar, combined with the best information available on galactic rotation; the current value of this effect is \(\dot P_{\rm{b}}^{{\rm{gal}}} \simeq  (0.0128 \pm 0.0050) \times {10^{ 12}}\). Subtracting this from the observed Ṗ_{b} (see Table 6) gives the corrected \(\dot P_{\rm{b}}^{{\rm{corr}}} =  (2.4056 \pm 0.0051) \times {10^{ 12}}\), which agrees with the prediction within the errors. In other words,
The consistency among the measurements is displayed in Figure 6, in which the regions allowed by the three most precise constraints have a single common overlap. Uncertainties in the parameters that go into the galactic correction are now the limiting factor in the accuracy of the test of gravitational damping.
A third way to display the agreement with GR is by comparing the observed phase of the orbit with a theoretical template phase as a function of time. If f_{b} varies slowly in time, then to first order in a Taylor expansion, the orbital phase is given by Φ_{b} (t) = 2πf_{b0}t + πḟ_{b0}t^{2}. The time of periastron passage t_{P} is given by Φ(t_{P}) = 2πN, where N is an integer, and consequently, the periastron time will not grow linearly with N. Thus the cumulative difference between periastron time t_{P} and N/f_{b0}, the quantities actually measured in practice, should vary according to \({t_{\rm{P}}}  N/{f_{{\rm{b}}0}} =  {{\dot f}_{{\rm{b0}}}}{N^2}/2f_{{\rm{b0}}}^3 \approx  ({{\dot f}_{{\rm{b0}}}}/2{f_{{\rm{b0}}}}){t^2}\). Figure 7 shows the results: The dots are the data points, while the curve is the predicted difference using the measured masses and the quadrupole formula for ḟ_{b0} [272].
The consistency among the constraints provides a test of the assumption that the two bodies behave as “point” masses, without complicated tidal effects, obeying the general relativistic equations of motion including gravitational radiation. It is also a test of strong gravity, in that the highly relativistic internal structure of the neutron stars does not influence their orbital motion, as predicted by the SEP of GR.
Recent observations [157, 271] indicate variations in the pulse profile, which suggests that the pulsar is undergoing geodetic precession on a 300year timescale as it moves through the curved spacetime generated by its companion (see Section 3.7.2). The amount is consistent with GR, assuming that the pulsar’s spin is suitably misaligned with the orbital angular momentum. Unfortunately, the evidence suggests that the pulsar beam may precess out of our line of sight by 2025.
A zoo of binary pulsars
Nine relativistic binary neutron star systems with orbital periods less than a day are now known. While some are less interesting for testing relativity, some have yielded interesting tests, and others, notably the recently discovered “double pulsar” are likely to produce significant results in the future. Here we describe some of the more interesting or best studied cases; the parameters of the first four are listed in Table 7.

B1534+12
This is a binary pulsar system in our galaxy [245, 243, 18]. Its pulses are significantly stronger and narrower than those of B1913+16, so timing measurements are more precise, reaching 3 µs accuracy. The orbital plane appears to be almost edgeon relative to the line of sight (i ≃ 80°); as a result the Shapiro delay is substantial, and separate values of the parameters r and s have been obtained with interesting accuracy. Assuming GR, one infers that the two masses are m_{1} = 1.335 ± 0.002 M_{⊙} and m_{2} = 1.344 ± 0.002 M_{⊙}. The rate of orbit decay Ṗ_{b} agrees with GR to about 15 percent, but the precision is limited by the poorly known distance to the pulsar, which introduces a significant uncertainty into the subtraction of galactic acceleration. Independently of Ṗ_{b}, measurement of the four other postKeplerian parameters gives two tests of strongfield gravity in the nonradiative regime [253].

B2127+11C
This system appears to be a clone of the HulseTaylor binary pulsar, with very similar values for orbital period and eccentricity (see Table 7). The inferred total mass of the system is 2.706 ± 0.011 M_{⊙}. But because the system is in the globular cluster M15 (NGC 7078), it suffers Doppler shifts resulting from local accelerations, either by the mean cluster gravitational field or by nearby stars, that are more difficult to estimate than was the case with the galactic system B1913+16. This makes a separate, precision measurement of the relativistic contribution to Ṗ_{b} essentially impossible.

J07373039A, B
This binary pulsar system, discovered in 2003 [48], was already remarkable for its extraordinarily short orbital period (0.1 days) and large periastron advance (16.88° yr^{−1}), but then the companion was also discovered to be a pulsar [175]. Because two projected semimajor axes can now be measured, one can obtain the mass ratio directly from the ratio of the two values of ap sin i, and thereby obtain the two masses by combining that ratio with the periastron advance, assuming GR. The results are m_{A} = 1.337±0.005 M_{⊙} and m_{B} = 1.250±0.005 M_{⊙}, where A denotes the primary (first) pulsar. From these values, one finds that the orbit is nearly edgeon, with sin i = 0.9991, a value which is completely consistent with that inferred from the Shapiro delay parameter (see Table 7). In fact, the five measured postKeplerian parameters plus the ratio of the projected semimajor axes give six constraints on the masses (assuming GR): All six overlap within their measurement errors. This system provides a unique opportunity for tight tests of strongfield and radiative effects in GR. Furthermore, it is likely that galactic proper motion effects will play a significantly smaller role in the interpretation of Ṗ_{b} measurements than they did in B1913+16.

J11416545
This is a case where the companion is probably a white dwarf [20, 18]. The masses of the pulsar and companion are 1.30±0.02 and 0.986±0.02 M_{⊙}, respectively. Ṗ_{b} has been measured to about 25 percent, consistent with the GR prediction. But because of the asymmetry in sensitivities (S_{NS} ∼ 0.2, S_{WD} ∼ 10^{−4}), there is the possibility, absent in the double neutronstar systems, to place a strong bound on scalartensor gravity (see Section 5.4).

J17562251
Discovered in 2004, this pulsar is in a binary system with a probable neutron star companion, with P_{b} = 7.67 hr, e = 0.18, and \(\dot \omega = 2.585 \pm 0.002\deg {\rm{y}}{{\rm{r}}^{{\rm{ 1}}}}\) [104].

J1906+0746
The discovery of this system was reported in late 2005 [174]. It is a young, 144ms pulsar in a relativistic orbit with P_{b} = 3.98 hr, e = 0.085, and \(\dot \omega = 7.57 \pm 0.03\deg {\rm{y}}{{\rm{r}}^{{\rm{ 1}}}}\).
Binary pulsars and alternative theories
Soon after the discovery of the binary pulsar it was widely hailed as a new testing ground for relativistic gravitational effects. As we have seen in the case of GR, in most respects, the system has lived up to, indeed exceeded, the early expectations.
In another respect, however, the system has only partially lived up to its promise, namely as a direct testing ground for alternative theories of gravity. The origin of this promise was the discovery that alternative theories of gravity generically predict the emission of dipole gravitational radiation from binary star systems. In GR, there is no dipole radiation because the “dipole moment” (center of mass) of isolated systems is uniform in time (conservation of momentum), and because the “inertial mass” that determines the dipole moment is the same as the mass that generates gravitational waves (SEP). In other theories, while the inertial dipole moment may remain uniform, the “gravity wave” dipole moment need not, because the mass that generates gravitational waves depends differently on the internal gravitational binding energy of each body than does the inertial mass (violation of SEP). Schematically, in a coordinate system in which the center of inertial mass is at the origin, so that m_{I,1}x_{1} + m_{I,2}x_{2} = 0, the dipole part of the retarded gravitational field would be given by
where v = v_{1} − v_{2} and η and m are defined using inertial masses. In theories that violate SEP, the difference between gravitational wave mass and inertial mass is a function of the internal gravitational binding energy of the bodies. This additional form of gravitational radiation damping could, at least in principle, be significantly stronger than the usual quadrupole damping, because it depends on fewer powers of the orbital velocity υ, and it depends on the gravitational binding energy per unit mass of the bodies, which, for neutron stars, could be as large as 20 percent (see TEGP 10 [281] for further details). As one fulfillment of this promise, Will and Eardley worked out in detail the effects of dipole gravitational radiation in the bimetric theory of Rosen, and, when the first observation of the decrease of the orbital period was announced in 1979, the Rosen theory suffered a terminal blow. A wide class of alternative theories also fails the binary pulsar test because of dipole gravitational radiation (TEGP 12.3 [281]).
On the other hand, the early observations of PSR 1913+16 already indicated that, in GR, the masses of the two bodies were nearly equal, so that, in theories of gravity that are in some sense “close” to GR, dipole gravitational radiation would not be a strong effect, because of the apparent symmetry of the system. The Rosen theory, and others like it, are not “close” to GR, except in their predictions for the weakfield, slowmotion regime of the solar system. When relativistic neutron stars are present, theories like these can predict strong effects on the motion of the bodies resulting from their internal highly relativistic gravitational structure (violations of SEP). As a consequence, the masses inferred from observations of the periastron shift and γ′ may be significantly different from those inferred using GR, and may be different from each other, leading to strong dipole gravitational radiation damping. By contrast, the BransDicke theory is “close” to GR, roughly speaking within 1/ω_{BD} of the predictions of the latter, for large values of the coupling constant ω_{BD}. Thus, despite the presence of dipole gravitational radiation, the binary pulsar provides at present only a weak test of BransDicke theory, not competitive with solarsystem tests.
Binary pulsars and scalartensor gravity
Making the usual assumption that both members of the system are neutron stars, and using the methods summarized in TEGP 10–12 [281], one can obtain formulas for the periastron shift, the gravitational redshift/secondorder Doppler shift parameter, and the rate of change of orbital period, analogous to Equations (78). These formulas depend on the masses of the two neutron stars, on their selfgravitational binding energy, represented by “sensitivities” s and κ*, and on the BransDicke coupling constant ω_{BD}. First, there is a modification of Kepler’s third law, given by
then, the predictions for \(\langle \dot \omega \rangle\), γ′ and Ṗ_{b} are
where \({\mathcal M} \equiv {\chi ^{3/5}}{{\mathcal G}^{ 4/5}}{\eta ^{3/5}}m\), and, to first order in ξ ≡ (2 + ω_{BD})^{−1}, we have
where F(e) is defined in Equation (72). The quantities s_{a} and \(k_{\rm{a}}^{\ast}\) are defined by
and measure the “sensitivity” of the mass m_{a} and moment of inertia I_{a} of each body to changes in the scalar field (reflected in changes in G) for a fixed baryon number N (see TEGP 11, 12 and 14.6 (c) [281] for further details). The quantity s_{a} is related to the gravitational binding energy. These sensitivities will depend on the neutronstar equation of state. Notice how the violation of SEP in BransDicke theory introduces complex structuredependent effects in everything from the Newtonian limit (modification of the effective coupling constant in Kepler’s third law) to gravitational radiation. In the limit ξ → 0, we recover GR, and all structure dependence disappears. The first term in Ṗ_{b} (see Equation (84)) is the combined effect of quadrupole and monopole gravitational radiation, while the second term is the effect of dipole radiation.
Unfortunately, because of the near equality of the neutron star masses in the binary pulsar, dipole radiation is suppressed, and the bounds obtained are not competitive with the Cassini bound on γ [293], except for those generalized scalartensor theories, with β_{0} < 0 [74]. Bounds on the parameters α_{0} and β_{0} from solar system, binary pulsar, and gravitational wave observations (see Sections 5.1 and 6.3) are found in [74].
Alternatively, a binary pulsar system with dissimilar objects, such as a white dwarf or black hole companion, would provide potentially more promising tests of dipole radiation. In this regard, the recently discovered binary pulsar J1141+6545, with an apparent white dwarf companion, may play an important role. Here one can treat s_{WD} ∼ 10^{−4} as negligible. Then, from Equation (84), it is straightforward to show that, if the timing reaches sufficient accuracy to determine Ṗ_{b} to an accuracy σ in agreement with the prediction of GR, then the resulting lower bound on ω_{BD} would be
Thus, for s_{NS} ∼ 0.2, a 4 percent measurement would already compete with the Cassini bound (for further details, see [118, 102]).
Gravitational Wave Tests of Gravitational Theory
Gravitational wave observatories
Some time in the next decade, a new opportunity for testing relativistic gravity will be realized, when a worldwide network of kilometerscale, laser interferometric gravitational wave observatories in the U.S. (LIGO project), Europe (VIRGO and GEO600 projects), and Japan (TAMA300 project) begins regular detection and analysis of gravitational wave signals from astrophysical sources. These broadband antennas will have the capability of detecting and measuring the gravitational waveforms from astronomical sources in a frequency band between about 10 Hz (the seismic noise cutoff) and 500 Hz (the photon counting noise cutoff), with a maximum sensitivity to strain at around 100 Hz of h ∼ Δl/l ∼ 10^{−22} (rms), for the kilometerscale LIGO/VIRGO projects. The most promising source for detection and study of the gravitational wave signal is the “inspiralling compact binary” — a binary system of neutron stars or black holes (or one of each) in the final minutes of a death spiral leading to a violent merger. Such is the fate, for example, of the HulseTaylor binary pulsar B1913+16 in about 300 Myr, or the “double pulsar” J07373039 in about 85 Myr. Given the expected sensitivity of the “advanced LIGO” (around 2010), which could see such sources out to many hundreds of megaparsecs, it has been estimated that from 40 to several hundred annual inspiral events could be detectable. Other sources, such as supernova core collapse events, instabilities in rapidly rotating newborn neutron stars, signals from nonaxisymmetric pulsars, and a stochastic background of waves, may be detectable (for reviews, see [1, 256]; for updates on the status of various projects, see [114, 45]).
A similar network of cryogenic resonantmass gravitational antennas have been in operation for many years, albeit at lower levels of sensitivity (h ∼ 10^{−19}). While modest improvements in sensitivity may be expected in the future, these resonant detectors are not expected to be competitive with the large interferometers, unless new designs involving masses of spherical, or nearly spherical shape come to fruition. These systems are primarily sensitive to waves in relatively narrow bands about frequencies in the hundreds to thousands of Hz range [206, 123, 32, 217], although future improvements in sensitivity and increases in bandwidth may be possible [61].
In addition, plans are being developed for an orbiting laser interferometer space antenna (LISA for short). Such a system, consisting of three spacecraft orbiting the sun in a triangular formation separated from each other by five million kilometers, would be sensitive primarily in the very low frequency band between 10^{−4} and 10^{−1} Hz, with peak strain sensitivity of order h ∼ 10^{−23} [90].
In addition to opening a new astronomical window, the detailed observation of gravitational waves by such observatories may provide the means to test general relativistic predictions for the polarization and speed of the waves, for gravitational radiation damping and for strongfield gravity.
Polarization of gravitational waves
A laser interferometric or resonant bar gravitational wave detector whose scale is small compared to the gravitational wavelength measures the local components of a symmetric 3 × 3 tensor which is composed of the “electric” components of the Riemann curvature tensor, R_{0i0j}, via the equation of geodesic deviation, given, for a pair of freely falling particles by ẍ^{i} = −R_{0i0j}x^{j}, where x^{i} denotes the spatial separation. In general there are six independent components, which can be expressed in terms of polarizations (modes with specific transformation properties under rotations and boosts). Three are transverse to the direction of propagation, with two representing quadrupolar deformations and one representing a monopolar “breathing” deformation. Three modes are longitudinal, with one an axially symmetric stretching mode in the propagation direction, and one quadrupolar mode in each of the two orthogonal planes containing the propagation direction. Figure 8 shows the displacements induced on a ring of freely falling test particles by each of these modes. General relativity predicts only the first two transverse quadrupolar modes (a) and (b) independently of the source; these correspond to the waveforms h_{+} and h_{×} discussed earlier (note the cos 2ϕ and sin 2ϕ dependences of the displacements). Massless scalartensor gravitational waves can in addition contain the transverse breathing mode (c). In massive scalartensor theories, the longitudinal mode (d) can also be present, but is suppressed relative to (c) by a factor (λ/λ_{C})^{2}, where λ is the wavelength of the radiation, and λ_{C} is the Compton wavelength of the massive scalar. More general metric theories predict additional longitudinal modes, up to the full complement of six (TEGP 10.2 [281]).
A suitable array of gravitational antennas could delineate or limit the number of modes present in a given wave. The strategy depends on whether or not the source direction is known. In general there are eight unknowns (six polarizations and two direction cosines), but only six measurables (R_{0i0j}). If the direction can be established by either association of the waves with optical or other observations, or by timeofflight measurements between separated detectors, then six suitably oriented detectors suffice to determine all six components. If the direction cannot be established, then the system is underdetermined, and no unique solution can be found. However, if one assumes that only transverse waves are present, then there are only three unknowns if the source direction is known, or five unknowns otherwise. Then the corresponding number (three or five) of detectors can determine the polarization. If distinct evidence were found of any mode other than the two transverse quadrupolar modes of GR, the result would be disastrous for GR. On the other hand, the absence of a breathing mode would not necessarily rule out scalartensor gravity, because the strength of that mode depends on the nature of the source.
Some of the details of implementing such polarization observations have been worked out for arrays of resonant cylindrical, diskshaped, spherical, and truncated icosahedral detectors (TEGP 10.2 [281], for recent reviews see [169, 266]); initial work has been done to assess whether the groundbased or spacebased laser interferometers (or combinations of the two types) could perform interesting polarization measurements [267, 47, 177, 117, 273]. Unfortunately for this purpose, the two LIGO observatories (in Washington and Louisiana states, respectively) have been constructed to have their respective arms as parallel as possible, apart from the curvature of the Earth; while this maximizes the joint sensitivity of the two detectors to gravitational waves, it minimizes their ability to detect two modes of polarization.
Gravitational radiation backreaction
In the binary pulsar, a test of GR was made possible by measuring at least three relativistic effects that depended upon only two unknown masses. The evolution of the orbital phase under the damping effect of gravitational radiation played a crucial role. Another situation in which measurement of orbital phase can lead to tests of GR is that of the inspiralling compact binary system. The key differences are that here gravitational radiation itself is the detected signal, rather than radio pulses, and the phase evolution alone carries all the information. In the binary pulsar, the first derivative of the binary frequency ḟ_{b} was measured; here the full nonlinear variation of f_{b} as a function of time is measured.
Broadband laser interferometers are especially sensitive to the phase evolution of the gravitational waves, which carry the information about the orbital phase evolution. The analysis of gravitational wave data from such sources will involve some form of matched filtering of the noisy detector output against an ensemble of theoretical “template” waveforms which depend on the intrinsic parameters of the inspiralling binary, such as the component masses, spins, and so on, and on its inspiral evolution. How accurate must a template be in order to “match” the waveform from a given source (where by a match we mean maximizing the crosscorrelation or the signaltonoise ratio)? In the total accumulated phase of the wave detected in the sensitive bandwidth, the template must match the signal to a fraction of a cycle. For two inspiralling neutron stars, around 16,000 cycles should be detected during the final few minutes of inspiral; this implies a phasing accuracy of 10^{−5} or better. Since υ ∼ 1/10 during the late inspiral, this means that correction terms in the phasing at the level of υ^{5} or higher are needed. More formal analyses confirm this intuition [67, 105, 68, 214].
Because it is a slowmotion system (υ ∼ 10^{−3}), the binary pulsar is sensitive only to the lowestorder effects of gravitational radiation as predicted by the quadrupole formula. Nevertheless, the first correction terms of order υ and υ^{2} to the quadrupole formula were calculated as early as 1976 [268] (see TEGP 10.3 [281]).
But for laser interferometric observations of gravitational waves, the bottom line is that, in order to measure the astrophysical parameters of the source and to test the properties of the gravitational waves, it is necessary to derive the gravitational waveform and the resulting radiation backreaction on the orbit phasing at least to 2PN order beyond the quadrupole approximation, and preferably to 3PN order.
For the special case of nonspinning bodies moving on quasicircular orbits (i.e. circular apart from a slow inspiral), the evolution of the gravitational wave frequency f = 2f_{b} through 2PN order has the form
where η = m_{1}m_{2}/m^{2}. The first term is the quadrupole contribution (compare Equation (72)), the second term is the 1PN contribution, the third term, with the coefficient 4π, is the “tail” contribution, and the fourth term is the 2PN contribution, first reported jointly by Blanchet et al. [40, 39, 291]. The 2.5PN, 3PN and 3.5PN contributions have also been calculated (see [34] for a review).
Similar expressions can be derived for the loss of angular momentum and linear momentum. Expressions for noncircular orbits have also been derived [121, 75]. These losses react back on the orbit to circularize it and cause it to inspiral. The result is that the orbital phase (and consequently the gravitational wave phase) evolves nonlinearly with time. It is the sensitivity of the broadband laser interferometric detectors to phase that makes the higherorder contributions to df/dt so observationally relevant.
If the coefficients of each of the powers of f in Equation (97) can be measured, then one again obtains more than two constraints on the two unknowns m_{1} and m_{2}, leading to the possibility to test GR. For example, Blanchet and Sathyaprakash [42, 41] have shown that, by observing a source with a sufficiently strong signal, an interesting test of the 4π coefficient of the “tail” term could be performed.
Another possibility involves gravitational waves from a small mass orbiting and inspiralling into a (possibly supermassive) spinning black hole. A general noncircular, nonequatorial orbit will precess around the hole, both in periastron and in orbital plane, leading to a complex gravitational waveform that carries information about the nonspherical, strongfield spacetime around the hole. According to GR, this spacetime must be the Kerr spacetime of a rotating black hole, uniquely specified by its mass and angular momentum, and consequently, observation of the waves could test this fundamental hypothesis of GR [231, 213].
Thirdly, the dipole gravitational radiation predicted by scalartensor theories will result in a modification of the gravitational radiation backreaction, and thereby of the phase evolution. Including only the leading quadrupole and dipole contributions, one obtains, in BransDicke theory,
where \({\mathcal M} = ({\chi ^{3/5}}{{\mathcal G}^{ 4/5}}){\eta ^{3/5}}m\), and b is the coefficient of the dipole term, given by \(b = (5/48)({\chi ^{ 1}}{{\mathcal G}^{4/3}})\xi {{\mathcal S}^2}\), where χ, \({\mathcal G},{\mathcal S}\) are given by Equations (94), and ξ = 1/(2 + ω_{BD}). Double neutron star systems are not promising because the small range of masses available near 1.4 M_{⊙} results in suppression of dipole radiation by symmetry. For black holes, s = 0.5 identically, consequently double black hole systems turn out to be observationally identical in the two theories. Thus mixed systems involving a neutron star and a black hole are preferred. However, a number of analyses of the capabilities of both groundbased and spacebased (LISA) observatories have shown that observing waves from neutronstarblackhole inspirals is not likely to bound scalartensor gravity at a level competitive with the Cassini bound or with future solarsystem improvements [283, 161, 236, 292, 27, 28].
Speed of gravitational waves
According to GR, in the limit in which the wavelength of gravitational waves is small compared to the radius of curvature of the background spacetime, the waves propagate along null geodesics of the background spacetime, i.e. they have the same speed c as light (in this section, we do not set c = 1). In other theories, the speed could differ from c because of coupling of gravitation to “background” gravitational fields. For example, in the Rosen bimetric theory with a flat background metric η, gravitational waves follow null geodesics of η, while light follows null geodesics of g (TEGP 10.1 [281]).
Another way in which the speed of gravitational waves could differ from c is if gravitation were propagated by a massive field (a massive graviton), in which case υ_{g} would be given by, in a local inertial frame,
where m_{g} and E are the graviton rest mass and energy, respectively.
The simplest attempt to incorporate a massive graviton into general relativity in a ghostfree manner suffers from the socalled van DamVeltmanZakharov (vDVZ) discontinuity [263, 299]. Because of the 3 additional helicity states available to the massive spin2 graviton, the limit of small graviton mass does not coincide with pure GR, and the predicted perihelion advance, for example, violates experiment. A model theory by Visser [265] attempts to circumvent the vDVZ problem by introducing a nondynamical flatbackground metric. This theory is truly continuous with GR in the limit of vanishing graviton mass; on the other hand, its observational implications have been only partially explored. Braneworld scenarios predict a tower or a continuum of massive gravitons, and may avoid the vDVZ discontinuity, although the full details are still a work in progress [91, 66].
The most obvious way to test this is to compare the arrival times of a gravitational wave and an electromagnetic wave from the same event, e.g., a supernova. For a source at a distance D, the resulting value of the difference 1 − υ_{g}/c is
where Δt = Δt_{a} − (1 + Z)Δt_{e} is the “time difference”, where Δt_{a} and Δt_{e} are the differences in arrival time and emission time of the two signals, respectively, and Z is the redshift of the source. In many cases, Δt_{e} is unknown, so that the best one can do is employ an upper bound on Δt_{e} based on observation or modelling. The result will then be a bound on 1 − υ_{g}/c.
For a massive graviton, if the frequency of the gravitational waves is such that hf ≫ m_{g}c^{2}, where h is Planck’s constant, then \({\upsilon _g}/c \approx 1  {1 \over 2}{(c/{\lambda _g}f)^2}\), where λ_{g} = h/m_{g}c is the graviton Compton wavelength, and the bound on 1 − υ_{g}/c can be converted to a bound on λ_{g}, given by
The foregoing discussion assumes that the source emits both gravitational and electromagnetic radiation in detectable amounts, and that the relative time of emission can be established to sufficient accuracy, or can be shown to be sufficiently small.
However, there is a situation in which a bound on the graviton mass can be set using gravitational radiation alone [285]. That is the case of the inspiralling compact binary. Because the frequency of the gravitational radiation sweeps from low frequency at the initial moment of observation to higher frequency at the final moment, the speed of the gravitons emitted will vary, from lower speeds initially to higher speeds (closer to c) at the end. This will cause a distortion of the observed phasing of the waves and result in a shorter than expected overall time Δt_{a} of passage of a given number of cycles. Furthermore, through the technique of matched filtering, the parameters of the compact binary can be measured accurately (assuming that GR is a good approximation to the orbital evolution, even in the presence of a massive graviton), and thereby the emission time Δt_{e} can be determined accurately. Roughly speaking, the “phase interval” fΔt in Equation (101) can be measured to an accuracy 1/ρ, where ρ is the signaltonoise ratio.
Thus one can estimate the bounds on λ_{g} achievable for various compact inspiral systems, and for various detectors. For stellarmass inspiral (neutron stars or black holes) observed by the LIGO/VIRGO class of groundbased interferometers, D ≈ 200 Mpc, f ≈ 100 Hz, and f Δt ∼ ρ^{−1} ≈ 1/10. The result is λ_{g} > 10^{13} km. For supermassive binary black holes (10^{4} to 10^{7} M_{⊙}) observed by the proposed laser interferometer space antenna (LISA), D ≈ 3 Gpc, f ≈ 10^{−3} Hz, and f Δt ∼ ρ^{−1} ≈ 1/1000. The result is λ_{g} > 10^{17} km.
A full noise analysis using proposed noise curves for the advanced LIGO and for LISA weakens these crude bounds by factors between two and 10 [285, 292, 27, 28]. For example, for the inspiral of two 10^{6} M_{⊙} black holes with aligned spins at a distance of 3 Gpc observed by LISA, a bound of 2 × 10^{16} km could be placed [27]. Other possibilities include using binary pulsar data to bound modifications of gravitational radiation damping by a massive graviton [106], and using LISA observations of the phasing of waves from compact whitedwarf binaries, eccentric galactic binaries, and eccentric inspiral binaries [69, 142].
Bounds obtainable from gravitational radiation effects should be compared with the solid bound λ_{g} > 2.8 × 10^{12} km [250] derived from solar system dynamics, which limit the presence of a Yukawa modification of Newtonian gravity of the form
and with the modeldependent bound λ_{g} > 6 × 10^{19} km from consideration of galactic and cluster dynamics [265].
Stronggravity tests
One of the central difficulties of testing GR in the strongfield regime is the possibility of contamination by uncertain or complex physics. In the solar system, weakfield gravitational effects could in most cases be measured cleanly and separately from nongravitational effects. The remarkable cleanliness of the binary pulsar permitted precise measurements of gravitational phenomena in a strongfield context.
Unfortunately, nature is rarely so kind. Still, under suitable conditions, qualitative and even quantitative strongfield tests of GR could be carried out.
One example is in cosmology. From a few seconds after the big bang until the present, the underlying physics of the universe is well understood, in terms of a Standard Model of a nearly spatially flat universe, 13.6 Gyr old, dominated by dark matter and dark energy. Some alternative theories of gravity that are qualitatively different from GR fail to produce cosmologies that meet even the minimum requirements of agreeing qualitatively with bigbang nucleosynthesis (BBN) or the properties of the cosmic microwave background (TEGP 13.2 [281]). Others, such as BransDicke theory, are sufficiently close to GR (for large enough ω_{BD}) that they conform to all cosmological observations, given the underlying uncertainties. The generalized scalartensor theories, however, could have small ω at early times, while evolving through the attractor mechanism to large ω today. One way to test such theories is through bigbang nucleosynthesis, since the abundances of the light elements produced when the temperature of the universe was about 1 MeV are sensitive to the rate of expansion at that epoch, which in turn depends on the strength of interaction between geometry and the scalar field. Because the universe is radiationdominated at that epoch, uncertainties in the amount of cold dark matter or of the cosmological constant are unimportant. The nuclear reaction rates are reasonably well understood from laboratory experiments and theory, and the number of light neutrino families (3) conforms to evidence from particle accelerators. Thus, within modest uncertainties, one can assess the quantitative difference between the BBN predictions of GR and scalartensor gravity under strongfield conditions and compare with observations. For recent analyses, see [234, 84, 58, 60].
Another example is the exploration of the spacetime near black holes and neutron stars via accreting matter. Studies of certain kinds of accretion known as advectiondominated accretion flow (ADAF) in lowluminosity binary Xray sources may yield the signature of the black hole event horizon [185]. The spectrum of frequencies of quasiperiodic oscillations (QPO) from galactic black hole binaries may permit measurement of the spins of the black holes [218]. Aspects of strongfield gravity and framedragging may be revealed in spectral shapes of iron fluorescence lines from the inner regions of accretion disks [225, 224]. Because of uncertainties in the detailed models, the results to date of studies like these are suggestive at best, but the combination of higherresolution observations and better modelling could lead to striking tests of strongfield predictions of GR.
Conclusions
We find that general relativity has held up under extensive experimental scrutiny. The question then arises, why bother to continue to test it? One reason is that gravity is a fundamental interaction of nature, and as such requires the most solid empirical underpinning we can provide. Another is that all attempts to quantize gravity and to unify it with the other forces suggest that the standard general relativity of Einstein is not likely to be the last word. Furthermore, the predictions of general relativity are fixed; the theory contains no adjustable constants so nothing can be changed. Thus every test of the theory is either a potentially deadly test or a possible probe for new physics. Although it is remarkable that this theory, born 90 years ago out of almost pure thought, has managed to survive every test, the possibility of finding a discrepancy will continue to drive experiments for years to come.
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Acknowledgments
This work has been supported since the initial version in part by the National Science Foundation, Grant Numbers PHY 9600049, 0096522, and 0353180, and by the National Aeronautics and Space Administration, Grant Number NAG510186. We also gratefully acknowledge the support of the Centre National de la Recherche Scientifique, and the hospitality of the Institut d’Astrophysique de Paris, where this update was completed. Comments from referees were particularly helpful in improving this update.