Abstract
As first discovered by Choptuik, the black hole threshold in the space of initial data for general relativity shows both surprising structure and surprising simplicity. Universality, powerlaw scaling of the black hole mass, and scale echoing have given rise to the term “critical phenomena”. They are explained by the existence of exact solutions which are attractors within the black hole threshold, that is, attractors of codimension one in phase space, and which are typically selfsimilar. Critical phenomena give a natural route from smooth initial data to arbitrarily large curvatures visible from infinity, and are therefore likely to be relevant for cosmic censorship, quantum gravity, astrophysics, and our general understanding of the dynamics of general relativity.
Introduction
Overview of the subject
Take generic initial data in general relativity, adjust any one parameter p of the initial data to the threshold of black hole formation, and compare the resulting spacetimes as a function of p. In many situations, the following critical phenomena are then observed:

Near the threshold, black holes with arbitrarily small masses can be created, and the black hole mass scales as
$$M \propto {(p  {p_\ast})^\gamma},$$(1)where p parameterises the initial data and black holes form for p > p_{*}.

The critical exponent γ is universal with respect to initial data, that is, independent of the particular 1parameter family, although it depends on the type of collapsing matter.

In the region of large curvature before black hole formation, the spacetime approaches a selfsimilar solution which is also universal with respect to initial data, the critical solution.
Critical phenomena were discovered by Choptuik [48] in numerical simulations of a spherical scalar field. They have been found in numerous other numerical and analytical studies in spherical symmetry, and a few in axisymmetry; in particular critical phenomena have been seen in the collapse of axisymmetric gravitational waves in vacuum [1]. Scaling laws similar to the one for black hole mass have been discovered for black hole charge and conjectured for black hole angular momentum.
It is still unclear how universal critical phenomena in collapse are with respect to matter types and beyond spherical symmetry, in particular for vacuum collapse. Progress is likely to be made over the next few years with better numerical simulations.
Critical phenomena can usefully be described in dynamical systems terms. A critical solution is then characterised as an attracting fixed point within a surface that divides two basins of attraction, a critical surface in phase space. Such a fixed point can be either a stationary spacetime, or one that is scaleinvariant and selfsimilar. The latter is relevant for the (“type II”) critical phenomena sketched above. We shall also see how the dynamical systems approach establishes a connection with critical phenomena in statistical mechanics.
Therefore, we could define the field of critical phenomena in gravitational collapse as the study of the boundaries among the basins of attraction of different end states of selfgravitating systems, such as black hole formation or dispersion. In our view, the main physical motivation for this study is that those critical solutions which are selfsimilar provide a way of achieving arbitrarily large spacetime curvature outside a black hole, and in the limit a naked singularity, by finetuning generic initial data for generic matter to the black hole threshold. Those solutions are therefore likely to be important for quantum gravity and cosmic censorship.
Plan of this review
Faced with more material to review, we have attempted to make this update shorter and more systematic than the original 1999 version [102]. We begin with the abstract theory in Section 2. Since 1999, progress on the theory side has mainly been made on the global spacetime structure of the critical solution and cosmic censorship in the spherical scalar field model. We have included this new material in an enlarged Section 3 on the spherical scalar field, although we hope it will turn out to be sufficiently generic to merit inclusion in Section 2. Nonspherical perturbations of the spherical scalar field are also discussed in Section 3.
In Section 4 we review the rich phenomenology that has been found in many other systems restricted to spherical symmetry. Numerical work in spherical symmetry has proliferated since 1999, but we have tried to keep this section as short as possible. There has been less progress in going beyond spherical symmetry than we anticipated in 1999, even though we continue to believe that important results await there. What is known today is summarised in Section 5.
The reader unfamiliar with the topic is advised to begin with either Sections 2.1, 2.2, and 2.3, which give the key theory of universality, selfsimilarity and scaling, or Sections 3.1 and 3.2, which describe the classic example, the massless scalar field.
This review is limited to numerical and theoretical work on phenomena at the threshold of black hole formation in 3 + 1dimensional general relativity. We report only briefly on work in higher and lower spacetime dimensions and nongravity systems that may be relevant as toy models for general relativity. We exclude other work on selfsimilarity in general relativity and work on critical phenomena in other areas of physics.
Other reviews on the subject are [129], [15], [96], [100], [50], [51], [33], [102], [156]. The 2002 review [105] by Gundlach gives more detailed explanations on some of the basic aspects of the theory. A review of the role of selfsimilarity in the formation of singularities in evolutionary PDEs in general is [71].
Theory
In this section we describe the basic theory underlying critical collapse of the type that forms arbitrarily small black holes (later called type II, see also Section 2.4). We begin with the mathematical origin of its three main characteristics, which were already summarised in the introduction:

universality with respect to initial data;

scaleinvariance of the critical solution;

black hole mass scaling.
Universality
Consider GR as an infinitedimensional continuous dynamical system. Points in the phase space are initial data sets (3metric, extrinsic curvature, and suitable matter variables, which together obey the Einstein constraints). We evolve with the Einstein equations in a suitable gauge (see Section 2.5). Solution curves of the dynamical system are spacetimes obeying the Einsteinmatter equations, sliced by specific Cauchy surfaces of constant time t.
An isolated system in GR can end up in qualitatively different stable end states. Two possibilities are the formation of a single black hole in collapse, or complete dispersion of the massenergy to infinity. For a massless scalar field in spherical symmetry, these are the only possible end states (see Section 3). Any point in phase space can be classified as ending up in one or the other type of end state. The entire phase space therefore splits into two halves, separated by a “critical surface”.
A phase space trajectory that starts on a critical surface by definition never leaves it. A critical surface is therefore a dynamical system in its own right, with one dimension fewer than the full system. If it has an attracting fixed point, such a point is called a critical point. It is an attractor of codimension one in the full system, and the critical surface is its attracting manifold. The fact that the critical solution is an attractor of codimension one is visible in its linear perturbations: It has an infinite number of decaying perturbation modes spanning the tangent plane to the critical surface, and a single growing mode not tangential to it.
As illustrated in Figures 1 and 2, any trajectory beginning near the critical surface, but not necessarily near the critical point, moves almost parallel to the critical surface towards the critical point. Near the critical point the evolution slows down, and eventually moves away from the critical point in the direction of the growing mode. This is the origin of universality. All details of the initial data have been forgotten, except for the distance from the black hole threshold. The closer the initial phase point is to the critical surface, the more the solution curve approaches the critical point, and the longer it will remain close to it. We should stress that this phase picture is extremely simplified. Some of the problems associated with this simplification are discussed in Section 2.5.
Selfsimilarity
Fixed points of dynamical systems often have additional symmetries. In the case of type II critical phenomena, the critical point is a spacetime that is selfsimilar, or scaleinvariant. These symmetries can be discrete or continuous. The critical solution of a spherically symmetric perfect fluid (see Section 4.2), has continuous selfsimilarity (CSS). A CSS spacetime is one that admits a homothetic vector field ξ, defined by [38]:
In coordinates x^{μ} = (τ, x^{i}) adapted to the symmetry, so that
the metric coefficients are of the form
where the coordinate τ is the negative logarithm of a spacetime scale, and the remaining three coordinates x^{i} can be thought of angles around the singular spacetime point τ = ∞ (see Section 3.3).
The critical solution of other systems, in particular the spherical scalar field (see Section 3) and axisymmetric gravitational waves (see Section 5.2), show discrete selfsimilarity (DSS). The simplest way of defining DSS is in adapted coordinates, where
such that g_{μν}(τ, x^{i}) is periodic in τ with period Δ. More formally, DSS can be defined as a discrete conformal isometry [95].
Using the gauge freedom of general relativity, the lapse and shift in the ADM formalism can be chosen (nonuniquely) so that the coordinates become adapted coordinates if and when the solution becomes selfsimilar (see Section 2.5). τ is then both a time coordinate (in the usual sense that surfaces of constant time are Cauchy surfaces), and the logarithm of overall scale at constant x^{i}. The minus sign in Equation (3) and hence Equations (4) and (5), is a convention assuming that smaller scales are in the future. The time parameter used in Figure 2 is of this type.
Mass scaling
Let Z stand for a set of scaleinvariant variables of the problem, such as \({\tilde g_{\mu v}}\) and suitably rescaled matter variables. If the dynamics is scaleinvariant (this is the case exactly for example for the scalar field, and approximately for other systems, see Section 2.6), then Z(x) is an element of the phase space factored by overall scale, and Z(x, τ) a solution. Note that Z(x) is an initial data set for GR only up to scale. The overall scale is supplied by τ.
For simplicity, assume that the critical solution is CSS. It can then be written as Z(x, τ) = Z_{*}(x). Its linear perturbations can depend on τ only exponentially. To linear order, the solution near the critical point must be of the form
The perturbation amplitudes C_{i} depend on the initial data, and hence on p. As Z_{*} is a critical solution, by definition there is exactly one λ_{i} with positive real part (in fact it is purely real), say λ_{0}. As τ → ∞, all other perturbations vanish. In the following we consider this limit, and retain only the one growing perturbation.
From our phase space picture, the evolution ends at the critical solution for p = p_{*}, so we must have C_{0}(p_{*}) = 0. Linearising in p around p_{*}, we obtain
For p ≠ p_{*}, but close to it, the solution has the approximate form (7) over a range of τ. Now we extract Cauchy data at one particular pdependent value of τ within that range, namely τ_{*} defined by
where ϵ is some constant ≪ 1 such that at this τ the linear approximation is still valid. At sufficiently large τ, the linear perturbation has grown so much that the linear approximation breaks down, and for C_{0} > 0 a black hole forms while for C_{0} < 0 the solution disperses. The crucial point is that we need not follow this evolution in detail, nor does the precise value of ϵ matter. It is sufficient to note that the Cauchy data at τ = τ_{*} are
Due to the funnelling effect of the critical solution, the data at τ_{*} is always the same, except for an overall scale, which is given by \({e^{ {\tau _\ast}}}\). For example, the physical spacetime metric, with dimension (length)^{2} is given by \({g_{\mu v}} = {e^{ 2\tau}}{{\tilde g}_{\mu v}}\), and similar scalings hold for the matter variables according to their dimension. In particular, as \({e^{ {\tau _\ast}}}\) is the only scale in the initial data (9), the mass of the final black hole must be proportional to that scale. Therefore
and, comparing with Equation (1), we have found the critical exponent γ = 1/λ_{0}.
When the critical solution is DSS, a periodic or fine structure of small amplitude is superimposed on this basic power law [98, 127]:
where f(z) has period Δ and is universal, and only c depends on the initial data. As the critical solution is periodic in τ with period Δ, the number N of scaling “echos” is approximated by
Note that this holds for both supercritical and subcritical solutions.
Type I
In type I critical phenomena, the same phase space picture as in Section 2.1 applies, but the critical solution is now stationary or timeperiodic instead of selfsimilar or scaleperiodic. It also has a finite mass and can be thought of as a metastable star. (Type I and II were so named after first and second order phase transitions in statistical mechanics, in which the order parameter is discontinuous and continuous, respectively.) Universality in this context implies that the black hole mass near the threshold is independent of the initial data, namely a certain fraction of the mass of the stationary critical solution. The dimensionful quantity that scales is not the black hole mass, but the lifetime t_{p} of the intermediate state where the solution is approximated by the critical solution. This is clearly
Type I critical phenomena occur when a mass scale in the field equations becomes dynamically relevant. (This scale does not necessarily set the mass of the critical solution absolutely: There could be a family of critical solutions selected by the initial conditions.) Conversely, as the type II power law is scaleinvariant, type II phenomena occur in situations where either the field equations do not contain a scale, or this scale is dynamically irrelevant. Many systems, such as the massive scalar field, show both type I and type II critical phenomena, in different regions of the space of initial data [34].
Coordinate choices for the dynamical systems picture
The time evolution of Cauchy data in GR can only be considered as a dynamical system if the ADM evolution equations are complemented by a prescription for the lapse and shift. To realise the phase space picture of Section 2.1, the critical solution must be a fixed point or limit cycle. We have seen how coordinates adapted to the selfsimilarity can be constructed, but is there a prescription of the lapse and shift for arbitrary initial data, such that, given initial data for the critical solution, the resulting time evolution actively drives the metric to a form (5) that explicitly displays the selfsimilarity?
Garfinkle and Gundlach [85] have suggested several combinations of lapse and shift conditions that leave CSS spacetimes invariant and turn the Choptuik DSS spacetime into a limit cycle (see [91, 81] for partial successes). Among these, the combination of maximal slicing with minimal strain shift has been suggested in a different context but for related reasons [188]. Maximal slicing requires the initial data slice to be maximal (\({K_a}^a = 0\)), but other prescriptions, such as freezing the trace of K together with minimal distortion, allow for an arbitrary initial slice with arbitrary spatial coordinates.
All these coordinate conditions are elliptic equations that require boundary conditions, and will turn CSS spacetimes into fixed points (or DSS into limit cycles) only given correct boundary conditions. Roughly speaking, these boundary conditions require a guess of how far the slice is from the accumulation point t = t_{*}, and answers to this problem only exist in spherical symmetry. Appropriate boundary conditions are also needed if the dynamical system is extended to include the lapse and shift as evolved variables, turning the elliptic equations for the lapse and shift into hyperbolic or parabolic equations.
Turning a CSS or stationary spacetime into a fixed point of the dynamical system also requires an appropriate choice of the phase space variables Z(x^{i}). To capture CSS (or DSS) solutions, one needs scaleinvariant variables. Essentially, these can be constructed by dimensional analysis. The coordinates x^{i} and τ are dimensionless, le^{−τ} has dimension length, and g_{μν} has dimension l^{2}. The scaling for the ADM and any matter variables follows.
Even with a prescription for the lapse and shift in place, a given spacetime does not correspond to a unique trajectory in phase space. Rather, for each initial slice through the same spacetime one obtains a different slicing of the entire spacetime. A possibility for avoiding this ambiguity would be to restrict the phase space further, for example by restricting possible data sets to maximal or constant extrinsic curvature slices.
Another open problem is that in order to talk about attractors and repellers on the phase space we need to define a norm on a suitable function space which includes both asymptotically flat data and data for the exact critical solution. The norm itself must favour the central region and ignore what is further out and asymptotically flat if all black holes of the same mass are to be considered as the same end state.
Approximate selfsimilarity and universality classes
The field equations for the massless scalar field coupled to the Einstein equations are scalefree. Realistic matter models introduce length scales, and the field equations then do not allow for exactly selfsimilar solutions. They may however admit solutions which are CSS or DSS asymptotically on small spacetime scales as the dimensionful parameters become irrelevant, including type II critical solutions [49, 34, 52]. This can be explored by a formal expansion in powers of the small parameter Le^{−τ}, where L is a parameter with dimensions length in the evolution equations. The zeroth order of the expansion is the selfsimilar critical solution of the system with L = 0. A similar ansatz can be made for the linear perturbations of the resulting background. The zeroth order of the background expansion determines Δ exactly and independently of L, and the zeroth order term of the linear perturbation expansion determines the critical exponent 1/λ_{0} exactly, so that there is no need in practice to calculate any higher orders in to make predictions for type II critical phenomena where they occur. (With L ≠ = 0, the basin of attraction of the type II critical solution will depend on L, and type I critical phenomena may also occur; see Section 2.4.) A priori, there could also be more than one type II critical solution for L = 0, although this has not been observed.)
This procedure has been carried out for the EinsteinYangMills system [97] and for massless scalar electrodynamics [106]. Both systems have a single length scale 1/e (in geometric units c = G = 1), where e is the gauge coupling constant. All values of e can be said to form one universality class of field equations [112] represented by e = 0. This notion of universality classes is fundamentally the same as in statistical mechanics. Other examples include modifications to the perfect fluid equation of state (EOS) that do not affect the limit of high density [161]. A simple example is that any scalar field potential V(ϕ) becomes dynamically irrelevant compared to the kinetic energy ∇ϕ^{2} in a selfsimilar solution [49], so that all scalar fields with potentials are in the universality class of the free massless scalar field. Surprisingly, even two different models like the SU(2) YangMills and SU(2) Skyrme models in spherical symmetry are members of the same universality class [22].
If there are several scales L_{0}, L_{1}, L_{2} etc. present in the problem, a possible approach is to set the arbitrary scale in Equation (29) equal to one of them, say L_{0}, and define the dimensionless constants l_{i} = L_{i}/L_{0} from the others. The scope of the universality classes depends on where the l_{i} appear in the field equations. If a particular L_{i} appears in the field equations only in positive integer powers, the corresponding l_{i} appears only multiplied by e^{−τ}, and will be irrelevant in the scaling limit. All values of this l_{i} therefore belong to the same universality class. From the example above, adding a quartic selfinteraction λϕ^{4} to the massive scalar field gives rise to the dimensionless number λ/m^{2} but its value is an irrelevant (in the language of renormalisation group theory) parameter.
Contrary to the statement in [106], we conjecture that massive scalar electrodynamics, for any values of e and m, is in the universality class of the massless uncharged scalar field in a region of phase space where type II critical phenomena occur. Examples of dimensionless parameters which do change the universality class are the k of the perfect fluid, the κ of the 2dimensional sigma model or, probably, a conformal coupling of the scalar field [47] (the numerical evidence is weak but a dependence should be expected).
The analogy with critical phase transitions
Some basic aspects of critical phenomena in gravitational collapse, such as finetuning, universality, scaleinvariant physics, and critical exponents for dimensionful quantities, can also be identified in critical phase transitions in statistical mechanics (see [208] for an introductory textbook).
From an abstract point of view, the objective of statistical mechanics is to derive relations between macroscopic observables A of the system and macroscopic external forces f acting on it, by considering ensembles of microscopic states of the system. The expectation values 〈A〉 can be generated as partial derivatives of the partition function
Here the μ are parameters of the Hamiltonian such as the strength of intermolecular forces, and f are macroscopic quantities which are being controlled, such as the temperature or magnetic field.
Phase transitions in thermodynamics are thresholds in the space of external forces f at which the macroscopic observables A, or one of their derivatives, change discontinuously. We consider two examples: the liquidgas transition in a fluid, and the ferromagnetic phase transition.
The liquidgas phase transition in a fluid occurs at the boiling curve p = p_{b}(T). In crossing this curve, the fluid density changes discontinuously. However, with increasing temperature, the difference between the liquid and gas density on the boiling curve decreases, and at the critical point (p_{*} = p_{b}(T_{*}), T_{*}) it vanishes as a noninteger power:
At the critical point an otherwise clear fluid becomes opaque, due to density fluctuations appearing on all scales up to scales much larger than the underlying atomic scale, and including the wavelength of light. This indicates that the fluid near its critical point is approximately scaleinvariant (for some range of scales between the size of molecules and the size of the container).
In a ferromagnetic material at high temperatures, the magnetisation m of the material (alignment of atomic spins) is determined by the external magnetic field B. At low temperatures, the material shows a spontaneous magnetisation even at zero external field. In the absence of an external field this breaks rotational symmetry: The system makes a random choice of direction. With increasing temperature, the spontaneous magnetisation m decreases and vanishes at the Curie temperature T_{*} as
Again, the correlation length, or length scale of a typical fluctuation, diverges at the critical point, indicating scaleinvariant physics.
Quantities such as m or ρ_{liquid} − ρ_{gas} are called order parameters. In statistical mechanics, one distinguishes between firstorder phase transitions, where the order parameter changes discontinuously, and secondorder, or critical, ones, where it goes to zero continuously. One should think of a critical phase transition as the critical point where a line of firstorder phase transitions ends as the order parameter vanishes. This is already clear in the fluid example. In the ferromagnet example, at first one seems to have only the one parameter T to adjust. But in the presence of a very weak external field, the spontaneous magnetisation aligns itself with the external field B, while its strength is to leading order independent of B. The function m(B, T) therefore changes discontinuously at B = 0. The line B = 0 for T < T_{*} is therefore a line of first order phase transitions between directions (if we consider one spatial dimension only, between m up and m down). This line ends at the critical point (B = 0, T = T_{*}) where the order parameter m vanishes. The critical value B = 0 of B is determined by symmetry; by contrast p_{*} depends on microscopic properties of the material.
We have already stated that a critical phase transition involves scaleinvariant physics. In particular, the atomic scale, and any dimensionful parameters associated with that scale, must become irrelevant at the critical point. This is taken as the starting point for obtaining properties of the system at the critical point.
One first defines a semigroup acting on microstates: the renormalisation group. Its action is to group together a small number of adjacent particles as a single particle of a fictitious new system by using some averaging procedure. This can also be done in a more abstract way in Fourier space. One then defines a dual action of the renormalisation group on the space of Hamiltonians by demanding that the partition function is invariant under the renormalisation group action:
The renormalised Hamiltonian is in general more complicated than the original one, but it can be approximated by the same Hamiltonian with new values of the parameters μ and external forces f. (At this stage it is common to drop the distinction between μ and f, as the new μ′ and f′ depend on both μ and f.) Fixed points of the renormalisation group correspond to Hamiltonians with the parameters at their critical values. The critical values of many of these parameters will be 0 or ∞, meaning that the dimensionful parameters they were originally associated with are irrelevant. Because a fixed point of the renormalisation group can not have a preferred length scale, the only parameters that can have nontrivial values are dimensionless.
The behaviour of thermodynamical quantities at the critical point is in general not trivial to calculate. But the action of the renormalisation group on length scales is given by its definition. The blowup of the correlation length ξ at the critical point is therefore the easiest critical exponent to calculate. The same is true for the black hole mass, which is just a length scale. We can immediately reinterpret the mathematics of Section 2.3 as a calculation of the critical exponent for ξ, by substituting the correlation length ξ for the black hole mass M, T_{*} − T for p − p_{*}, and taking into account that the τevolution in critical collapse is towards smaller scales, while the renormalisation group flow goes towards larger scales: ξ therefore diverges at the critical point, while M vanishes.
In type II critical phenomena in gravitational collapse, we should think of the black hole mass as being controlled by the functions P and Q on phase space defined by Equation (27). Clearly, P is the equivalent of the reduced temperature T − T_{*}. Gundlach [104] has suggested that the angular momentum of the initial data can play the role of B, and the final black hole angular momentum the role of m. Like the magnetic field, angular momentum is a vector, with a critical value that must be zero because all other values break rotational symmetry.
The Scalar Field
Critical phenomena in gravitational collapse were first discovered by Choptuik [47, 48, 49] in the model of a spherically symmetric, massless scalar field ϕ minimally coupled to general relativity. The scalar field matter is both simple, and acts as a toy model in spherical symmetry for the effects of gravitational radiation. Given that it is still the beststudied model in spherical symmetry, we review it here as a case study. For other numerical work on this model, see [109, 80, 111, 77, 182, 212]. Important analytical studies of gravitational collapse in this model have been carried out by Christodoulou [57, 58, 59, 60, 61, 62, 63].
We first review the field equations and Choptuik’s observations at the black hole threshold, mainly as a concrete example for the general ideas discussed above. We then summarise more recent work on the global structure of Choptuik’s critical solution, which throws an interesting light on cosmic censorship. In particular, the exact critical solution contains a curvature singularity that is locally and globally naked, and any critical solution obtained in the limit of perfect finetuning of asymptotically flat initial data is at least locally naked. By perturbing around spherical symmetry, the stability of the Choptuik solution in the full phase space can be investigated, and the scaling of black hole angular momentum can be predicted. By embedding the real scalar field in scalar electrodynamics and perturbing around the Choptuik solution, the scaling of black hole charge can be predicted.
Field equations in spherical symmetry
The Einstein equations are
and the matter equation is
Note that the matter equation of motion is contained within the contracted Bianchi identities. Choptuik chose Schwarzschildlike coordinates,
where dΩ^{2} = dθ^{2} + sin^{2} θ dφ^{2} is the metric on the unit 2sphere. This choice of coordinates is defined by the radius r giving the surface area of 2spheres as 4πr^{2}, and by t being orthogonal to r (polarradial coordinates). One more condition is required to fix the coordinate completely. Choptuik chose α = 1 at r = 0, so that t is the proper time of the central observer.
In the auxiliary variables
the wave equation becomes a firstorder system,
In spherical symmetry there are four algebraically independent components of the Einstein equations. Of these, one is a linear combination of derivatives of the other and can be disregarded. The other three contain only first derivatives of the metric, namely a_{,t}, a_{,r}, and α_{,r}, and are
Because of spherical symmetry, the only dynamics is in the scalar field equations (22, 23). The metric can be found by integrating the ODEs (24) and (25) for a and α at any fixed t, given ϕ and Π. Equation (26) can be ignored in this “fully constrained” evolution scheme.
The black hole threshold
The free data for the system are the two functions Π(r, 0) and Φ(r, 0). Choptuik investigated several 1parameter families of such data by evolving the data for many different values of the parameter. Simple examples of such families are Π(r, 0) = 0 and a Gaussian for Φ(r, 0), with the parameter p taken to be either the amplitude of the Gaussian, with the width and centre fixed, or the width, with position and amplitude fixed, or the position, with width and amplitude fixed. For sufficiently small amplitude (or the peak sufficiently wide), the scalar field will disperse, and for sufficiently large amplitude it will form a black hole.
Generic 1parameter families behave in this way, but this is difficult to prove in generality. Christodoulou showed for the spherically symmetric scalar field system that data sufficiently weak in a welldefined way evolve to a Minkowskilike spacetime [58, 61], and that a class of sufficiently strong data forms a black hole [60].
Choptuik found that in all 1parameter families of initial data he investigated he could make arbitrarily small black holes by finetuning the parameter p close to the black hole threshold. An important fact is that there is nothing visibly special to the black hole threshold. One cannot tell that one given data set will form a black hole and another one infinitesimally close will not, short of evolving both for a sufficiently long time.
As p → p_{*} along the family, the spacetime varies on ever smaller scales. Choptuik developed numerical techniques that recursively refine the numerical grid in spacetime regions where details arise on scales too small to be resolved properly. In the end, he could determine p_{*} up to a relative precision of 10^{−15}, and make black holes as small as 10^{−6} times the ADM mass of the spacetime. The powerlaw scaling (10) was obeyed from those smallest masses up to black hole masses of, for some families, 0.9 of the ADM mass, that is, over six orders of magnitude [49]. There were no families of initial data which did not show the universal critical solution and critical exponent. Choptuik therefore conjectured that γ is the same for all 1parameter families of smooth, asymptotically flat initial data that depend smoothly on the parameter, and that the approximate scaling law holds ever better for arbitrarily small p − p_{*}.
It is an empirical fact that typical 1parameter families cross the threshold only once, so that there is every indication that it is a smooth submanifold, as we assumed in the phase space picture. Taking into account the discussion of mass scaling above, we can formally write the black hole mass as a functional of the initial data z = (ϕ(r, 0), Π(r, 0)) exactly as
where P and Q are smooth functions on phase space and H is the Heaviside function. (Q could be absorbed into P.)
In hindsight, polarradial gauge is welladapted to selfsimilarity. In this gauge, DSS corresponds to
for any integer n, where Z stands for any one of the dimensionless quantities a, α or ϕ (and therefore also for rΠ or rΦ). With
DSS is
The dimensionful constants t_{*} and L depend on the particular 1parameter family of solutions, but the dimensionless critical fields a_{*}, α_{*} and ϕ_{*}, and in particular their dimensionless period Δ, are universal. Empirically, Δ ≃ 3.44 for the scalar field in numerical time evolutions, and Δ = 3.445452402(3) from a numerical construction of the critical solution based on exact selfsimilarity and analyticity [157].
Global structure of the critical solution
In adapted coordinates, the metric of the critical spacetime is of the form \({e^{ 2\tau}}\) times a regular metric. From this general form alone, one can conclude that τ= ∞ is a curvature singularity, where Riemann and Ricci invariants blow up like \({e^{4\tau}}\) (unless the spacetime is flat), and which is at finite proper time from regular points in its past. The Weyl tensor with index position C^{abcd} is conformally invariant, so that components with this index position remain finite as τ → ∞. This type of singularity is called “conformally compactifiable” [198] or “isotropic” [93]. For a classification of all possible global structures of spherically symmetric selfsimilar spacetimes see [107].
The global structure of the scalar field critical solution was determined accurately in [157] by assuming analyticity at the centre of spherical symmetry and at the past light cone of the singularity (the selfsimilarity horizon, or SSH). The critical solution is then analytic up to the future lightcone of the singularity (the Cauchy horizon, or CH). Global adapted coordinates x and τ can be chosen so that the regular centre r = 0, the SSH and the CH are all lines of constant x, and surfaces of constant τ are never tangent to x lines. (A global τ is no longer a global time coordinate.) This is illustrated in Figure 4.
Approaching the CH, the scalar field oscillates an infinite number of times but with the amplitude of the oscillations decaying to zero. The scalar field in regular adapted coordinates (x, τ) is of the form
where F_{reg}(τ), F_{sing}(τ) and H(τ) are periodic with period Δ, and the SSH is at x = 0. These functions have been computed numerically to high accuracy, together with the constants K and ϵ. The scalar field itself is smooth with respect to τ, and as ϵ > 0, it is continuous but not differentiable with respect to x on the CH itself. The same is true for the metric and the curvature. Surprisingly, the ratio m/r of the Hawking mass over the area radius on the CH is of order 10^{−6} but not zero (the value is known to eight significant figures).
As the CH itself is regular with smooth null data except for the singular point at its base, it is not intuitively clear why the continuation is not unique. A partial explanation is given in [157], where all DSS continuations are considered. Within a DSS ansatz, the solution just to the future of the CH has the same form as Equation (31). F_{reg}(τ) is the same on both sides, but F_{sing}(τ) can be chosen freely on the future side of the CH. Within the restriction to DSS this function can be taken to parameterise the information that comes out of the naked singularity.
There is precisely one choice of F_{sing}(τ) on the future side that gives a regular centre to the future of the CH, with the exception of the naked singularity itself, which is then a point. This continuation was calculated numerically, and is almost but not quite Minkowski in the sense that m/r remains small everywhere to the future of the SSH.
All other DSS continuations have a naked, timelike central curvature singularity with negative mass. More exotic continuations including further CHs would be allowed kinematically [42] but are not achieved dynamically if we assume that the continuation is DSS. The spacetime diagram of the generic DSS continued solution is given in Figure 3.
Nearcritical spacetimes and naked singularities
Choptuik’s results have an obvious bearing on the issue of cosmic censorship (see [203] for a general review of cosmic censorship). Roughly speaking, finetuning to the black hole threshold provides a set of data which is codimension one in the space of generic, smooth, asymptotically flat initial data, and whose time evolution contains at least the point singularity of the critical solution. The cosmic censorship hypothesis must therefore be formulated as “generic smooth initial data for reasonable matter do not form naked singularities”. Here we look at the relation between finetuning and naked singularities in more detail.
Christodoulou [63] proves rigorously that naked singularity formation is not generic, but in a rather larger function space, functions of bounded variation, than one would naturally consider. In particular, the instability of the naked singularity found by Christodoulou is not differentiable on the past light cone. This is unnatural in the context of critical collapse, where the naked singularity can arise from generic (up to finetuning) smooth initial data, and the intersection of the past light cone of the singularity with the initial data surface is as smooth as the initial data elsewhere. It is therefore not clear how this theorem relates to the numerical and analytical results strongly indicating that naked singularities are codimension1 generic within the space of smooth initial data.
First, consider the exact critical solution. The lapse α defined by Equation (20) is bounded above and below in the critical solution. Therefore the redshift measured between constant r observers located at any two points on an outgoing radial null geodesic in the critical spacetime to the past of the CH is bounded above and below. Within the exact critical solution, a point with arbitrarily high curvature can therefore be observed from a point with arbitrarily low curvature. Next consider a spacetime where the critical solution in a central region is smoothly matched to an asymptotically flat outer region such that the resulting asymptotically flat spacetime contains a part of the critical solution that includes the singularity and a part of the CH. In this spacetime, a point of arbitrarily large curvature can be seen from ℐ^{+} with finite redshift. This is illustrated in Figure 4.
Now consider the evolution of asymptotically flat initial data that have been finetuned to the black hole threshold. The global structure of such spacetimes has been investigated numerically in [111, 77, 80, 182]. Empirically, these spacetimes can be approximated near the singularity by Equation (6). Almost all perturbations decay as the singularity is approached and the approximation becomes better, until the one growing perturbation (which by the assumption of finetuning starts out small) becomes significant. A maximal value of the curvature is then reached which is still visible from ℐ^{+} and which scales as [84]
Finally, consider the limit of perfect finetuning. The growing mode is then absent, and all other modes decay as the naked singularity is approached. Note that Equation (6) suggests this is true regardless of the direction (future, past, or spacelike, depending on the value of x) in which the singularity is approached. This seems to be in conflict with causality: The issue is sensitively connected to the completeness of the modes Z_{i} and to the stability of the CH, and requires more investigation.
A complication in the supercritical case has been pointed out in [182]. In the literature on supercritical evolutions, what is quoted as the black hole mass is in fact the mass of the first apparent horizon (AH) that appears in the time slicing used by the code (spacelike or null). The black hole mass can and generically will be larger than the AH mass when it is first measured because of matter falling in later, and the region of maximal curvature may well be inside the event horizon, and hidden from observers at ℐ^{+} (see Figure 5 for an illustration.) The true black hole mass can only be measured at ℐ^{+}, where it is defined to be the limit of the Bondi mass m_{B} as the Bondi time u_{B} → ∞. This was implemented in [182]. Only one family of initial data was investigated, but in this family it was found that m_{B} converges to 10^{−4} of the initial Bondi mass in the fine tuning limit. More numerical evidence would be helpful, but the result is plausible. As the underlying physics is perfectly scaleinvariant in the massless scalar field model, the minimum mass must be determined by the family of initial data through the infall of matter into the black hole. Simulations of critical collapse of a perfect fluid in a cosmological context show a similar lower bound [119] due to matter falling back after shock formation, but this may not be true for all initial data [160].
Electric charge
Given the scaling power law for the black hole mass in critical collapse, one would like to know what happens if one takes a generic 1parameter family of initial data with both electric charge and angular momentum (for suitable matter), and finetunes the parameter p to the black hole threshold. A simple model for charged matter is a complex scalar field coupled to electromagnetism with the substitution ∇_{a} → ∇_{a} + ieA_{a}, or scalar electrodynamics. (Note that in geometric units, black hole charge Q has dimension of length, but the charge parameter e has dimension 1/length.)
Gundlach and MartínGarcía [106] have studied scalar massless electrodynamics in spherical symmetry perturbatively. Clearly, the real scalar field critical solution of Choptuik is a solution of this system too. In fact, it remains a critical solution within massless scalar electrodynamics in the sense that it still has only one growing perturbation mode within the enlarged solution space. Some of its perturbations carry electric charge, but as they are all decaying, electric charge is a subdominant effect. The charge of the black hole in the critical limit is dominated by the most slowly decaying of the charged modes. From this analysis, a universal powerlaw scaling of the black hole charge
was predicted. The predicted value δ ≃ 0.88 of the critical exponent (in scalar electrodynamics) was subsequently verified in collapse simulations by Hod and Piran [126] and later again by Petryk [176]. (The mass scales with γ ≃ 0.37 as for the uncharged scalar field.) No other type of criticality can be found in the phase space of this system as dispersion and black holes are the only possible end states, though black holes with Q ≳ M can be formed [176].
General considerations similar to those in Section 2.6 led Gundlach and MartínGarcía to the general prediction that the two critical exponents are always related, for any matter model, by the inequality
(with the equality holding if the critical solution is charged), so that black hole charge can always be treated perturbatively at the black hole threshold. This has not yet been verified in any other matter model.
Selfinteraction potential
An example of the richer phenomenology in the presence of a scale in the field equations is the spherical massive scalar field with a potential m^{2}ϕ^{2} [34] coupled to gravity: In one region of phase space, with characteristic scales smaller than 1/m, the black hole threshold is dominated by the Choptuik solution and type II critical phenomena occur. In another it is dominated by metastable oscillating boson stars (whose mass is of order 1/m in geometric units) and type I critical phenomena occur. (For the real scalar field, the type I critical solution is an (unstable) oscillating boson star [186] while for the complex scalar field it can be a static (unstable) boson star [120].)
When the scalar field with a potential is coupled to electromagnetism, type II criticality is still controlled by a solution which asymptotically resembles the uncharged Choptuik spacetime, but type I criticality is now controlled by charged boson stars [176]. There are indications that subcritical type I evolutions lead to slow, large amplitude oscillations of stable boson stars [141, 142, 176] and not to dispersion to infinity, as had been conjectured in [120]. Another interesting extension is the study of the dynamics of a real scalar field with a symmetric doublewell potential, in which the system displays type I criticality between the two possible vacua [128].
Nonspherical perturbations: Stability and angular momentum
Critical collapse is really relevant for cosmic censorship only if it is not restricted to spherical symmetry. MartínGarcía and Gundlach [154] have analysed all nonspherical perturbations of the scalar field critical solution by solving a linear eigenvalue problem with an ansatz of regularity at the centre and the SSH. They find that the only growing mode is the known spherical one, while all other spherical modes and all nonspherical modes decay. This strongly suggests that the critical solution is an attractor of codimension one not only in the space of spherically symmetric data but (modulo linearisation stability) of all data in a finite neighbourhood of spherical symmetry.
More recently, Choptuik and collaborators [54] have carried out axisymmetric time evolutions for the massless scalar field using adaptive mesh refinement. They find that in the limit of finetuning generic axisymmetric initial data the spherically symmetric critical solution is approached at first but then deviates from spherical symmetry and eventually develops two centres, each of which approaches the critical solution and bifurcates again in a universal way. This suggests that the critical solution has nonspherical growing perturbation modes, possibly a single l = 2 even parity mode (in axisymmetry, only m = 0 is allowed). There appears to be a conflict between the time evolution results [54] and the perturbative results [154], which needs to be resolved by more work (see Section 5.2).
Perturbing the scalar field around spherical symmetry, angular momentum comes in to second order in perturbation theory. All angular momentum perturbations were found to decay, and a critical exponent μ, ≃ 0.76 for the angular momentum was derived for the massless scalar field in [87]. This prediction has not yet been tested in nonlinear collapse simulations.
More Spherical Symmetry
The pioneering work of Choptuik on the spherical massless scalar field has been followed by a plethora of further investigations. These could be organised under many different criteria. We have chosen the following rough categories:

Systems in which the field equations, when reduced to spherical symmetry, form a single wavelike equation, typically with explicit rdependence in its coefficients. This includes YangMills fields, sigma models, vector and spinor fields, scalar fields in 2 + 1 or in 4 + 1 and more spacetime dimensions, and scalar fields in a semiclassical approximation to quantum gravity.

Perfect fluid matter, either in an asymptotically flat or a cosmological context. The linearised Euler equations are in fact wavelike, but the full nonlinear equations admit shock heating and are therefore not even timereversal symmetric.

Collisionless matter described by the Vlasov equation is a partial differential equation on particle phase space as well as spacetime. Therefore even in spherical symmetry, the matter equation is a partial differential equation in 4 dimensions (rather than two). Intuitively speaking, there are infinitely more matter degrees of freedom than in the scalar field or in nonspherical vacuum gravity.

Spherically symmetric nonlinear wave equations on 3 +1 Minkowski spacetime, and other nonlinear partial differential equations which show a transition between singularity formation and dispersal.
Some of these examples were constructed because they may have intrinsic physical relevance (semiclassical gravity, primordial black holes), others as toy models for 3 + 1dimensional gravity, and others mostly out of a purely mathematical interest. Table 1 gives an overview of these models.
Matter obeying wave equations
2dimensional nonlinear σ model
We have already discussed in Section 3.6 the effects of adding a potential to the evolution of the scalar field. An alternative generalization is a modification of the kinetic term in the Lagrangian, with the general form
for an Ndimensional vector field ^{ϕI}(x) with I = 1,…, N, and where G_{IJ}(ϕ^{K}) is a fixed nonlinear function acting as a metric on the socalled target space of the fields ϕ^{I}. Such a system is called a nonlinear σ model, harmonic map or wave map. The fields ϕ^{I} and metric G_{IJ} are dimensionless, and this allows the introduction of dimensionless parameters in the system, which cannot be asymptotically neglected using the arguments of Section 2.6. (Compare with the potential V(ϕ^{I}), which has dimensions (length)^{−2}, and hence requires dimensionful parameters.)
The case N = 1 gives nothing new, and so Hirschmann and Eardly (HE from now on) studied the N = 2 case with a target manifold with constant curvature [124], proportional to a real dimensionless constant κ. Using a single complex coordinate ϕ the action of the system can be written as
for κ ≥ 0 this system is equivalent to the problem of a real massless scalar field coupled to BransDicke (BD) gravity (with the BD coupling constant given by 8ω_{BD} = −12 + κ^{−1}). Liebling and Choptuik [150] have shown that there is a smooth transition in the BD system from DSS criticality for low κ (the flat target space case κ = 0 is equivalent to a selfgravitating complex massless scalar field, whose critical solution is the original DSS spacetime found by Choptuik) to CSS criticality for larger κ (the case κ = 1 is equivalent to the axiondilaton system, and has been shown to display CSS criticality in [110]).
Generalizing their previous results for κ = 0 [123, 122], HE constructed for each κ a CSS solution based on the ansatz ϕ(τ, x) = e^{iωτ} ϕ(x) for the critical scalar field. Studying its perturbations HE concluded that this solution is critical for κ > 0.0754, but has three unstable modes for κ < 0.0754 and even more for κ > −0.28. Below 0.0754 a DSS solution takes over, as shown in the simulations of Liebling and Choptuik [150], and HE conjectured that the transition is a Hopf bifurcation, such that the DSS cycle smoothly shrinks with growing κ, collapsing onto the CSS solution at the transition and then disappearing with a finite value of the echoing period Δ.
The close relation between the CSS and DSS critical solutions is also manifest in the construction of their global structures. In particular, the results of [122] and [69] for the CSS κ = 0 and κ = 1 solutions respectively show that the Cauchy horizon of the singularity is almost but not quite flat, exactly as was the case with the Choptuik DSS spacetime (see Section 3.3).
Spherical EinsteinSU(2) sigma model
This is an N = 3 sigma model, and it also displays a transition between CSS and DSS criticality, but this is a totally different type of transition, in particular showing a divergence in the echoing period Δ [131].
In a reduction to spherical symmetry, the effective action is
where r is the area radius, and the coupling constant η is dimensionless. It has been shown (numerically [27] and then analytically [28]) that for 0 ≤ η < 1/2, there is an infinite sequence of CSS solutions ϕ_{n} labelled by a nodal number n, and having n growing modes. (The case η = 0, in which the sigma field decouples from gravity, will be revisited below.) The nth solution is always regular in the past light cone of the singularity, but is regular up to the future light cone only for η < η_{n} where η_{0} ≃ 0.0688, η_{1} ≃ 0.1518, and η_{n} < η_{n+1} < 1/2. For larger couplings an apparent horizon develops and the solution cannot be smoothly continued. These results suggest that ϕ_{0} is a stable naked singularity for η < η_{0}, and ϕ_{1} acts as a critical solution between naked singularity formation and dispersal for η < η_{0} and between black hole formation and dispersal for η_{0} < η < η_{1}. The numerical experiments agree with this scenario in the range 0 ≤ η < 0.14. Other CSS solutions of this system are investigated in [25], and the possibility of chaos in [195].
Aichelburg and collaborators [131, 144] have shown that for η ≳ 0.2 there is clear DSS type II criticality at the black hole threshold. The period Δ depends on η, monotonically decreasing towards an asymptotic value for η → ∞. Interesting new behavior occurs in the intermediate range 0.14 < η < 0.2 that lies between clear CSS and clear DSS. With decreasing η the overall DSS includes episodes of approximate CSS [197], of increasing length (measured in the logscale time τ). As η → η_{c} ≃ 0.170 from above the duration of the CSS epochs, and hence the overall DSS period Δ diverges. For 0.14 < η < 0.17 time evolutions of initial data near the black hole threshold no longer show overal DSS, but they still show CSS episodes. Black hole mass scaling is unclear in this regime.
It has been conjectured that this transition from CSS to DSS can be interpreted, in the language of the theory of dynamical systems, as the infinitedimensional analogue of a 3dimensional Shil’nikov bifurcation [145]. Highprecision numerics in [3] further supports this picture: For η > η_{c} a codimension1 CSS solution coexists in phase space with a codimension1 DSS attractor such that the (1dimensional) unstable manifold of the DSS solution lies on the stable manifold of the CSS solution. For η close to η_{c} the two solutions are close and the orbits around the DSS solution become slower because they spend more time in the neighbourhood of the CSS attractor. A linear stability analysis predicts a law \(\Delta \simeq  {2 \over \lambda}\log (\eta  {\eta _c}) + b\) for some constant b, where λ is the Lyapunov exponent of the CSS solution. For η = η_{c} both solutions touch and the DSS cycle dissapears.
EinsteinYangMills
Choptuik, Chmaj and Bizoń [52] have found both type I and type II critical collapse in the spherical EinsteinYangMills system with SU(2) gauge potential, restricting to the purely magnetic case, in which the matter is described by a single real scalar field. The situation is very similar to that of the massive scalar field, and now the critical solutions are the wellknown static n = 1 BartnikMcKinnon solution [10] for type I and a DSS solution (later constructed in [97]) for type II. In both cases the black holes produced in the supercritical regime are Schwarzschild black holes with zero YangMills field strength, but the final states (and the dynamics leading to them) can be distinguished by the value of the YangMills final gauge potential at infinity, which can take two values, corresponding to two distinct vacuum states.
Choptuik, Hirschmann and Marsa [56] have investigated the boundary in phase space between formation of those two types of black holes, using a code that can follow the time evolutions for long after the black hole has formed. This is a new “type III” phase transition whose critical solution is an unstable static black hole with YangMills hair [14, 202], which collapses to a hairless Schwarzschild black hole with either vacuum state of the YangMills field, depending on the sign of its one growing perturbation mode. This “coloured” black hole is actually a member of a 1parameter family parameterized by its apparent horizon radius and outside the horizon it approaches the corresponding BM solution. When the horizon radius approaches zero the three critical solutions meet at a “triple point”. What happens there deserves further investigation.
Millward and Hirschmann [158] have further coupled a Higgs field to the EinsteinYangMills system. New possible end states appear: regular static solutions, and stable hairy black holes (different from the coloured black holes referred to above). Again there are type I or type II critical phenomena depending on the initial conditions.
It is known that the spherical critical solutions within the magnetic ansatz become more unstable when other components of the gauge field are taken into account, and so they will not be critical in the general case.
Vacuum 4 + 1
Bizoń, Chmaj and Schmidt [20] have found a way of constructing asymptotically flat vacuum spacetimes in 4 + 1 dimensions which are spherically symmetric while containing gravitational waves (Birkhoff’s theorem does not hold in more than 3 + 1 dimensions). Recall that in (3 + 1 dimensional) Bianchi IX cosmology the manifold is M^{1} × S^{3} where the S^{3} is equipped with an SU(2) invariant (homogeneous but anisotropic) metric
where the σ^{i} are the SU(2) leftinvariant 1forms, and the L_{i} are functions of time only. Similarly, the spacetimes of Bizoń, Chmaj and Schmidt are of the product form M^{2} × S^{3} where the L_{i} now depend only on r and t. This gives rise to nontrivial dynamics, including a threshold between dispersion and black hole formation. With the additional U(1) symmetry L_{1} = L_{2} (biaxial solutions) there is only one dynamical degree of freedom. At the black hole threshold, type II critical phenomena are seen with Δ ≃ 0.49 and γ ≃ 0.3289.
In evolutions with the general ansatz where all L_{i} are different (triaxial solutions) [21], the U(1) symmetry is recovered dynamically in the approach to the critical surface. However, each biaxial solution, and in particular the critical solution, exists in three copies obtained by permutation of the σ_{i}. Therefore, in the triaxial case, the critical surface contains three critical solutions. The boundaries, within the critical surface, between their basins of attraction contain in turn codimensiontwo DSS attractors. It is conjectured that there is in fact a countable family of DSS solutions with η unstable modes.
Szybka and Chmaj [196] give numerical evidence that these boundaries within the critical surface are fractal (in contrast to the critical surface itself, which is smooth as in all other known systems.)
A similar ansatz can be made in other odd spacetime dimensions, and in 8 + 1 dimensions type II critical behaviour is again observed [19].
Scalar field collapse in 2 + 1
Spacetime in 2 + 1 dimensions is flat everywhere where there is no matter, so that gravity is not acting at a distance in the usual way. There are no gravitational waves, and black holes can only be formed in the presence of a negative cosmological constant (see [39] for a review).
Scalar field collapse in circular symmetry was investigated numerically by Pretorius and Choptuik [178], and Husain and Olivier [134]. In a regime where the cosmological constant is small compared to spacetime curvature they find type II critical phenomena with a universal CSS critical solution, and γ = 1.20 ± 0.05 [178]. The value γ ≃ 0.81 [134] appears to be less accurate.
Looking for the critical solution in closed form, Garfinkle [82] found a countable family of exact spherically symmetric CSS solutions for a massless scalar field with Λ = 0, but his results remain inconclusive. The q = 4 solution appears to match the numerical evolutions inside the past light cone, but its past light cone is also an apparent horizon. The q = 4 solution has three growing modes although the top one would give γ = 8/7 ≃ 1.14 if only the other two could be ruled out [86]. An attempt at this [125] seems unmotivated. At the same time, it is possible to embed the Λ = 0 solutions into a family of Λ < 0 ones [64, 65, 43], which can be constructed along the lines of Section 2.6, so that Garfinkle’s solution could be the leading term in an expansion in e^{−τ} (−Λ)^{−1/2}.
Scalar field collapse in higher dimensions
Critical collapse of a massless scalar field in spherical symmetry in 5 + 1 spacetime dimensions was investigated in [83]. Results are similar to 3 + 1 dimensions, with a DSS critical solution and mass scaling with γ ≃ 0.424. Birukou et al [13, 133] have developed a code for arbitrary spacetime dimension. They confirm known results in 3 + 1 (γ − 0.36) and 5 + 1 (γ ≃ 0.44) dimensions, and investigate 4 + 1 dimensions. Without a cosmological constant they find mass scaling with γ ≃ 0.41 for one family of initial data and γ ≃ 0.52 for another. They see wiggles in the ln M versus ln(p − p*) plot that indicate a DSS critical solution, but have not investigated the critical solution directly. With a negative cosmological constant and the second family, they find γ = 0.49. Bland and Kunstatter [29] have made a more precise determination: γ = 0.4131 ± 0.0001. This was motivated by an attempt to explain this exponent using an holographic duality between the strong coupling regime of 4 + 1 gravity and the weak coupling regime of 3 + 1 QCD [5], which had predicted γ = 0.409552.
Kol [140] relates a solution that is related to the Choptuik solution to a variant of the critical solution in the blackstring black hole transition, and claims to obtain analytic estimates for γ and Δ. This has motivated a numerical determination of γ and Δ for the spherical massless scalar field in noninteger dimension up to 14 [190, 30].
Other systems obeying wave equations
Choptuik, Hirschmann and Liebling [53] have presented perturbative indications that the static solutions found by van Putten [199] in the vacuum BransDicke system are critical solutions. They have also performed full numerical simulations, but only starting from small deviations with respect to those solutions.
Ventrella and Choptuik [200] have performed numerical simulations of collapse of a massless Dirac field in a special state: an incoherent sum of two independent lefthanded zerospin fields having opposite orbital angular momentum. This is prepared so that the total distribution of energymomentum is spherically symmetric. The freedom in the system is then contained in a single complex scalar field obeying a modified linear wave equation in spherical symmetry. There are clear signs of CSS criticality in the metric variables, and the critical complex field exhibits a phase of the form e^{iωτ} for a definite ω (the Hirschmann and Eardly ansatz for the complex scalar field critical solutions), which can be considered as a trivial form of DSS.
Garfinkle, Mann and Vuille [90] have found coexistence of types I and II criticality in the spherical collapse of a massive vector field (the Proca system), the scenario being almost identical to that of a massive scalar field. In the selfsimilar phase the collapse amplifies the longitudinal mode of the Proca field with respect to its transverse modes, which become negligible, and the critical solution is simply the gradient of the Choptuik DSS spacetime.
Sarbach and Lehner [185] find type I critical behaviour in q + 3dimensional spacetimes with U(1) × SO(q + 1) symmetry in EinsteinMaxwell theory at the threshold between dispersion and formation of a black string.
Perfect fluid matter
Spherical symmetry
Evans and Coleman [72] performed the first simulations of critical collapse with a perfect fluid with EOS p = kρ (where ρ is the energy density and p the pressure) for k = 1/3 (radiation), and found a CSS critical solution with a massscaling critical exponent γ ≈ 0.36. Koike, Hara and Adachi [138] constructed that critical solution and its linear perturbations from a CSS ansatz as an eigenvalue problem, computing the critical exponent to high precision. Independently, Maison [153] constructed the regular CSS solutions and their linear perturbations for a large number of values of k, showing for the first time that the critical exponents were modeldependent. As Ori and Piran before [171, 172], he claimed that there are no regular CSS solutions for k > 0.89, but Neilsen and Choptuik [161, 162] have found CSS critical solutions for all values of k right up to 1, both in collapse simulations and by making a CSS ansatz. The difficulty comes from a change in character of the sonic point, which becomes a nodal point for k > 0.89, rather than a focal point, making the ODE problem associated with the CSS ansatz much more difficult to solve. Harada [114] has also found that the critical solution becomes unstable to a “kink” (discontinuous at the sonic point of the background solution) mode for k > 0.89, but because it is not smooth it does not seem to have any influence on the numerical simulations of collapse. On the other hand, the limit k → 1 leading to the stiff EOS p = ρ is singular in that during evolution the fluid 4velocity can become spacelike and the density ρ negative. The stiff fluid equations of motion are in fact equivalent to the masslessscalar field, but the critical solutions can differ, dependending on how one deals with the issue of negative density [35]. Summarizing, it is possible to construct the EvansColeman CSS critical (codimension1) solution for all values 0 < k < 1. This solution can be identified in the general classification of CSS perfectfluid solutions as the unique spacetime that is analytic at the center and at the sound cone, is ingoing near the center, and outgoing everywhere else [40, 41, 42]. There is even a Newtonian counterpart of the critical solution: the Hunter (a) solution [115]. ÁlvarezGaumé et al. [7] have calculated γ using perturbation theory for the spherically symmetric perfect fluid with P = kρ for k ≳ 1/4 in d =5, 6, 7 dimensions.
p = kρ is the only EOS compatible with exact CSS (homothetic) solutions for perfect fluid collapse [38] and therefore we might think that other equations of state would not display critical phenomena, at least of type II. Neilsen and Choptuik [161] have given evidence that for the ideal gas EOS \(p = k{\rho _0}\epsilon\) (where ρ_{0} is the rest mass density and ϵ is the internal energy per rest mass unit) the black hole threshold also contains a CSS attractor, and that it coincides with the CSS exact critical solution of the ultrarelativistic case with the same k. This is interpreted a posteriori as a sign that the critical CSS solution is highly ultrarelativistic, ρ = (1 + ϵ)ρ_{0} ≃ ϵρ_{0} ≫ ρ_{0}, and hence rest mass is irrelevant. Novak [167] has also shown in the case k = 1, or even with a more general tabulated EOS, that type II critical phenomena can be found by velocityinduced perturbations of static TOV solutions. A thorough and much more precise analysis by Noble and Choptuik [165, 166] of the possible collapse scenarios of the stiff k = 1 ideal gas has confirmed this surprising result, and again the critical solution (and hence the critical exponent) is that of the ultrarelativistic limit problem. Parametrizing, as usual, the TOV solutions by the central density ρ_{c}, they find that for lowdensity initial stars it is not possible to form a black hole by velocityinduced collapse; for intermediate initial values of ρ_{c}, it is possible to induce type II criticality for large enough velocity perturbations; for large initial central densities they always get type I criticality, as we might have anticipated.
Noble and Choptuik [165] have also investigated the evolution of a perfect fluid interacting with a massless scalar field indirectly through gravity. By tuning of the amplitude of the pulse it is possible to drive a fluid star to collapse. For massive stars type I criticality is found, in which the critical solution oscillates around a member of the unstable TOV branch. For less massive stars a large scalar amplitude is required to induce collapse, and the black hole threshold is always dominated by the scalar field DSS critical solution, with the fluid evolving passively.
Nonspherical perturbations
Nonspherically symmetric perturbations around the spherical critical solution for the perfect fluid can be used to study angular momentum perturbatively. All nonspherical perturbations of the perfect fluid critical solution decay for equations of state p = kρ with k in the range 1/9 < k < 0.49 [101], and so the spherically symmetric critical solution is stable under small deviations from spherical symmetry. Infinitesimal angular momentum is carried by the axial parity perturbations with angular dependence l = 1. From these two facts one can derive the angular momentum scaling law at the black hole threshold [99, 103],
which should be valid in the range 1/9 < k < 0.49. The angular momentum exponent μ(k) is related to the mass exponent γ(k) by
where
is the growth or decay rate of the dominant l = 1 axial perturbation mode. In particular for the value k =1/3, where γ ≃ 0.3558, μ, μ = (5/2)γ ≃ 0.8895.
Ori and Piran [172] have pointed out that there exists a CSS perfect fluid solution for 0 < k < 0.036 generalizing the LarsonPenston solution of Newtonian fluid collapse, and which has a naked singularity for 0 < k < 0.0105. Harada and Maeda [113, 115] have shown that this solution has no growing perturbative modes in spherical symmetry and hence a naked singularity becomes a global attractor of the evolution for the latter range of k. This is also true in the limit k = 0, which can be considered as the Newtonian limit [116, 117]. Their result has been confirmed with very high precision numerics by Snajdr [189]. This seems to violate cosmic censorship, as generic spherical initial data would create a naked singularity. However, the exact result (41) holds for any regular CSS spherical perfect fluid solution, and so all such solutions with k < 1/9 have at least one unstable nonspherical perturbation. Therefore the naked singularity is unstable to infinitesimal perturbations with angular momentum when one lifts the restriction to spherical symmetry.
Cosmological applications
In the early universe, quantum fluctuations of the metric and matter can be important, for example providing the seeds of galaxy formation. Large enough fluctuations will collapse to form primordial black holes. As large quantum fluctuations are exponentially more unlikely than small ones, \(P(\delta) \sim {e^{ {\delta ^2}}}\), where δ is the density contrast of the fluctuation, one would expect the spectrum of primordial black holes to be sharply peaked at the minimal δ that leads to black hole formation, giving rise to critical phenomena [163]. See also [94, 209].
An approximation to primordial black hole formation is a spherically symmetric distribution of a radiation gas (p = ρ/3) with cosmological rather than asymptotically flat boundary conditions. In [163, 164] type II critical phenomena were found, which would imply that the mass of primordial black holes formed are much smaller than the naively expected value of the mass contained within the Hubble horizon at the time of collapse. The boundary conditions and initial data were refined in [119, 160], and a minimum black hole mass of ∼ 10^{−4} of the horizon mass was found, due to matter accreting onto the black hole after strong shock formation. However, when the initial data are constructed more realistically from only the growing cosmological perturbation mode, no minimum mass is found [177, 159].
Collisionless matter
A cloud of collisionless particles can be described by the Vlasov equation, i.e., the Boltzmann equation without collision term. This matter model differs from field theories by having a much larger number of matter degrees of freedom: The matter content is described by a statistical distribution f(x^{μ}, p_{ν}) on the point particle phase space, instead of a finite number of fields ϕ(x^{μ}). When restricted to spherical symmetry, individual particles move tangentially as well as radially, and so individually have angular momentum, but the stressenergy tensor averages out to a spherically symmetric one, with zero total angular momentum. The distribution f is then a function f(r, t, p^{r}, L^{2}) of radius, time, radial momentum and (conserved) angular momentum.
Several numerical simulations of critical collapse of collisionless matter in spherical symmetry have been published to date, and remarkably no type II scaling phenomena has been discovered. Indications of type I scaling have been found, but these do not quite fit the standard picture of critical collapse. Rein et al. [183] find that black hole formation turns on with a mass gap that is a large part of the ADM mass of the initial data, and this gap depends on the initial matter condition. No critical behavior of either type I or type II was observed. Olabarrieta and Choptuik [168] find evidence of a metastable static solution at the black hole threshold, with type I scaling of its life time as in Equation (13). However, the critical exponent depends weakly on the family of initial data, ranging from 5.0 to 5.9, with a quoted uncertainty of 0.2. Furthermore, the matter distribution does not appear to be universal, while the metric seems to be universal up to an overall rescaling, so that there appears to be no universal critical solution. More precise computations by Stevenson and Choptuik [192], using finite volume HRSC methods, have confirmed the existence of static intermediate solutions and nonuniversal scaling with exponents ranging now from 5.27 to 11.65.
MartínGarcía and Gundlach [155] have constructed a family of CSS spherically symmetric solutions for massless particles that is generic by function counting. There are infinitely many solutions with different matter configurations but the same stressenergy tensor and spacetime metric, due to the existence of an exact symmetry: Two massless particles with energymomentum p^{μ} in the solution can be replaced by one particle with 2p^{μ}. A similar result holds for the perturbations. As the growth exponent λ of a perturbation mode can be determined from the metric alone, this means that there are infinitely many perturbation modes with the same λ. If there is one growing perturbative mode, there are infinitely many. Therefore a candidate critical solution (either static or CSS) cannot be isolated or have only one growing mode. This argument rules out the existence of both type I and type II critical phenomena (in their standard form, i.e., including universality) for massless particles in the complete system, but some partial form of criticality could still be found by restricting to sections of phase space in which that symmetry is broken, for example by prescribing a fixed form for the dependence of the distribution function f on angular momentum L, as those numerical simulations have done.
A recent investigation of Andréasson and Rein [8] with massive particles has confirmed again the existence of a mass gap and the existence of metastable static solutions at the black hole threshold, though there is no estimation of the scaling of their lifetimes. More interestingly, they show that the subcritical regime can lead to either dispersion or an oscillating steady state depending on the binding energy of the system. They also conclude, based on perturbative arguments, that there cannot be an isolated universal critical solution.
More numerical work is still required, but current evidence suggests that there are no type II critical phenomena, and that there is a continuum of critical solutions in type I critical phenomena and hence only limited universality.
Criticality in singularity formation without gravitational collapse
It is well known that the YangMills field does not form singularities from smooth initial conditions in 3 + 1 dimensions [70], but Bizoń and Tabor [26] have shown singularity formation in 4 + 1 (the critical dimension for this system from the point of view of energy scaling arguments) and 5 + 1 (the first supercritical dimension). In 5 + 1 there is the countable family W_{n} of CSS solutions with n unstable modes, such that W_{1} acts as a critical solution separating singularity (W_{0}) formation from dispersal to infinity. In 4 + 1 there are no selfsimilar solutions and the formation of singularities seems to proceed through adiabatic shrinking of a static solution.
Completely parallel results can be found for wave maps, for which the critical dimension is 2 + 1. For the wave map from 3 + 1 Minkowski to the 3sphere, Bizoń [16] has shown that there is a countable family of regular (before the CH) CSS solutions labeled by a nodal number ≥ 0, such that each solution has n unstable modes. Simulations of collapse in spherical symmetry [23, 151] and in 3 dimensions [149] show that n = 0 is a global attractor and the n = 1 solution is the critical solution (see also [67, 68] for computations of the largest perturbationeigenvalues of W_{0} and W_{1}). Again, for the wave map from 2 + 1 Minkowski to the 2sphere generic singularity formation proceeds through adiabatic shrinking of a static solution [24].
These results have led to the suggestion in [26] that criticality (in the sense of the existence of a codimension1 solution separating evolution towards qualitatively different end states) could be a generic and robust feature of evolutionary PDE systems in supercritical dimensions, and not an effect particular of gravity.
Garfinkle and Isenberg [88] examine the threshold between the round end state and pinching off in Ricci flow for a familiy of spherically symmetric geometries on S^{3}. They have found intermediate approach to a special “javelin” geometry, but have not investigated whether this is universal. See also [89] and [135].
A scaling of the shape of the event horizon at the moment of merger in binary black hole mergers is noted in [44], but this is really a kinematic effect.
Analytic studies and toy models
Exact solutions of EinsteinKleinGordon
A number of authors have explored the possibility of finding critical phenomena with CSS (rather than DSS) massless scalar critical solutions. The Roberts 1parameter family of 3 + 1 solutions [184] has been analyzed along this line in [173, 32, 206, 137]. This family contains black holes whose masses (with a suitable matching to an asymptotically flat solution) scale as (p − 1)^{1/2} for p ≳ 1, but such a special family of solutions has no direct relevance for collapse from generic data. Its generalization to other dimensions has been considered in [75]. A fully analytic construction of all (spherical and nonspherical) linear perturbations of the Roberts solution by Frolov [73, 74] has shown that there is a continuum of unstable spherical modes filling a sector of the complex plane with Re λ ≥ 1, so that it cannot be a critical solution. Interestingly, all nonspherical perturbations decay.
Frolov [76] has also suggested that the critical (p = 1) Roberts solution, which has an outgoing null singularity, plus its most rapidly growing (spherical) perturbation mode would evolve into the Choptuik solution, which would inherit the oscillation in τ with a period 4.44 of that mode.
A similar transition within a single 1parameter family of solutions has been pointed out in [170] for the Wyman solution [207].
Hayward [121, 66] and Clement and Fabbri [64, 65] have also proposed critical solutions with a null singularity, and have attempted to construct black hole solutions from their linear perturbations. This is probably irrelevant to critical collapse, as the critical spacetime does not have an outgoing null singularity. Rather, the singularity is naked but first appears in a point. The future light cone of that point is not a null singularity but a CH with finite curvature.
Other authors have attempted analytic approximations to the Choptuik solution. Pullin [181] has suggested describing critical collapse approximately as a perturbation of the Schwarzschild spacetime. Price and Pullin [180] have approximated the Choptuik solution by two flat space solutions of the scalar wave equation that are matched at a “transition edge” at constant selfsimilarity coordinate x. The nonlinearity of the gravitational field comes in through the matching procedure, and its details are claimed to provide an estimate of the echoing period Δ.
Toy models
Birmingham and Sen [12] considered the formation of a black hole from the collision of two point particles of equal mass in 2 + 1 gravity. Peleg and Steif [175] have investigated the collapse of a dust ring. In both cases the mass of the black holes is a known function of the parameters of the initial condition, giving a “critical exponent” 1/2, but no underlying selfsimilar solution is involved.
Mahajan et al. [152] expand the initial data for Einstein clusters in powers of the radius and, assuming that there are no shell crossings, find a mass scaling exponent of 3/2 for two such expansion coefficients. Universality is not demonstrated, and so the connection with a CSS solution discussed by Harada and Mahajan [118] is unclear.
Frolov [78], and Frolov, Larsen and Christensen [79] consider a stationary 2 + 1dimensional NambuGoto membrane held fixed at infinity in a stationary 3 + 1dimensional black hole background spacetime. The induced 2 + 1 metric on the membrane can have wormhole, black hole, or Minkowski topology. The critical solution between Minkowski and black hole topology has 2 + 1 CSS. The mass of the apparent horizon of induced black hole metrics scales with γ= 2/3, superimposed with a wiggle of period \(3\pi/\sqrt 7\) in ln p. The mass scaling is universal with respect to different background black hole metrics, as they can be approximated by Rindler space in the mass scaling limit.
Horowitz and Hubeny [130], and Birmingham [11] have attempted to calculate the critical exponent in toy models from the adSCFT correspondence. ÁlvarezGaumhé et al. have attempted to use the adSCFT correspondence for calculating γ from the QCD side for the spherical massless scalar field in 5 dimensions [5], and the spherical perfect fluid with P = kρ for k = 1/(d − 1) in d = 5, 6, 7 dimensions [7].
ÁlvarezGaumé et al. [6] have calculated critical exponents for the formation of an apparent horizon in the collision of two gravitational shock waves.
Burko [37] considers the transition between existence and nonexistence of a null branch of the singularity inside a spherically symmetric charged black hole with massless scalar field matter thrown in.
Wang [205] has constructed homothetic cylindrically symmetric solutions of 3 + 1 EinsteinKleinGordon and studied their cylindrically symmetric perturbations. It is not clear how these are related to a critical surface in phase space.
Quantum effects
Type II critical phenomena provide a relatively natural way of producing arbitrarily high curvatures, where quantum gravity effects should become important, from generic initial data. Approaching the Planck scale from above, one would expect to be able to write down a critical solution that is the classical critical solution asymptotically at large scales, as an expansion in inverse powers of the Planck length (see Section 2.6).
Black hole evolution in semiclassical gravity has been investigated in 1 + 1 dimensional models which serve as toy models for spherical symmetry (see [92] for a review). The black hole threshold in such models has been investigated in [45, 9, 194, 137, 210, 174, 31]. In some of these models, the critical exponent is 1/2 for kinematical reasons.
In [36] a 3 + 1dimensional but perturbative approach is taken. The quantum effects then give an additional unstable mode with λ = 2. If this is larger than the positive Lyapunov exponent λ_{0}, it will become the dominant perturbation for sufficiently good finetuning, and therefore sufficiently good finetuning will reveal a mass gap. The mass gap is found also in numerical evolutions of a spherical scalar field in 3 + 1 dimensions with the semiclassical equations obtained in the framework of “singularity resolution” in loop quantum gravity [132, 211].
Beyond Spherical Symmetry
Numerical studies of critical collapse should go beyond spherical symmetry (and in the first instance to axisymmetry) for three reasons:

Weak gravitational waves in vacuum general relativity can focus and collapse. The black hole threshold in this process shows what in critical phenomena in gravitational collapse is intrinsic to gravity rather than the matter model.

Black holes are characterised by charge and angular momentum as well as mass. Angular momentum is the more interesting of the two because it is again independent of matter, but cannot be studied in spherical symmetry.

Angular momentum resists collapse, but angular momentum in the initial data is needed to make a black hole with angular momentum. Therefore it is an interesting question to ask what happens to the dimensionless ratio J/M^{2} at the black hole threshold.
In the following we review what has been done so far.
Perturbative approach to angular momentum
We have already mentioned that when angular momentum is small, a critical exponent for J can be derived in perturbation theory. This has been done for the perfect fluid (see Section 4.2) in firstorder perturbation theory and for the massless scalar field (see Section 3.7), where secondorder perturbation theory in the scalar field is necessary to obtain an angular momentum perturbation in the stressenergy tensor [87]. However, neither of the predicted angular momentum scaling laws has been verified in numerical evolutions.
For a perfect fluid with EOS p = kρ with 0 < k < 1/9, precisely one mode that carries angular momentum is unstable, and this mode and the known spherical mode are the only two unstable modes of the spherical critical solution. (Note by comparison that from dimensional analysis one would not expect an uncharged critical solution to have a growing perturbation mode carrying charge.) The presence of two growing modes of the critical solution is expected to give rise to interesting phenomena [104]. Near the critical solution, the two growing modes compete. J and M of the final black hole are expected to depend on the distance to the black hole threshold and the angular momentum of the initial data through universal functions of one variable that are similar to “universal scaling functions” in statistical mechanics (see also the end of Section 2.7). While they have not yet been computed, these functions can in principle be determined from time evolutions of a single 2parameter family of initial data, and then determine J and M for all initial data near the black hole threshold and with small angular momentum. They would extend the simple powerlaw scalings of J and M into a region of initial data space with larger angular momentum.
Axisymmetric vacuum gravity
Abrahams and Evans [1] have numerically investigated black hole formation in axisymmetric vacuum gravity. They write the metric as
where the lapse a, shift components β^{r} and β^{θ}, and 3metric coefficients ϕ and η are functions of r, t, and θ. Axisymmetry limits gravitational waves to one polarisation out of two, so that there are as many physical degrees of freedom as in a single wave equation. On the initial slice, η and \(K_\theta ^r\) are given as free data, and \(\phi, K_r^r\), and \(K_\varphi ^\varphi\) are determined by solving the Hamiltonian constraint and the two independent components of the momentum constraint. Afterwards,\(\eta, K_\theta ^r,K_r^r\), and \(K_\varphi ^\varphi\) are evolved, and only ϕ is obtained by solving the Hamiltonian constraint. α is obtained by solving the maximal slicing condition \({K_i}^i = 0\) for α, and β^{r} and β^{θ} are obtained from the time derivatives of the quasiisotropic spatial gauge conditions g_{θθ} = r^{2}g_{rr} and g_{r}θ = 0.
In order to keep their numerical grid as small as possible, Abrahams and Evans chose their initial data to be mostly ingoing. The two free functions \(K_\theta ^r\) in the initial data were chosen to have the same functional form they would have in a linearised gravitational wave with pure l = 2, m = 0 angular dependence. This ansatz reduced the freedom in the initial data to one free function of advanced time. A specific peaked function was chosen, and only the overall amplitude was varied.
Limited numerical resolution allowed Abrahams and Evans to find black holes with masses only down to 0.2 of the ADM mass. Even this far from criticality, they found powerlaw scaling of the black hole mass, with a critical exponent γ ≃ 0.36. The black hole mass was determined from the apparent horizon surface area, and the frequencies of the lowest quasinormal modes of the black hole. There was tentative evidence for scale echoing in the time evolution, with Δ ≃ 0.6, with about three echos seen. Here η has the echoing property η(e^{Δ}r, e^{Δ}t) = η(r, t), and the same echoing property is expected to hold also for α, ϕ, β^{r}, and r^{−1}β^{θ}. In a subsequent paper [2], some evidence for universality of the critical solution, echoing period and critical exponent was given in the evolution of a second family of initial data, one in which η= 0 at the initial time. In this family, black hole masses down to 0.06 of the ADM mass were achieved.
It is striking that at 14 years later, these results have not yet been independently verified. An attempt with a 3dimensional numerical relativity code (but axisymmetric initial data), using free evolution in the BSSN formulation with maximal slicing and zero shift, to repeat the results of Abrahams and Evans was not successful [4]. The reason could be a combination of rather lower resolution than that of Abrahams and Evans, growing constraint violations, and an inappropriate choice of gauge. (It is now known that BSSN with maximal slicing and zero shift is an illposed system [108].)
Scalar field
In 2003, Choptuik, Hirschmann, Liebling and Pretorius reported on numerical evolutions at the black hole threshold of an axisymmetric massless scalar field [54]. In axisymmetry with scalar field matter there is no angular momentum and only one polarisation of gravitational waves. The slicing condition is maximal slicing and the spatial gauge, in cylindrical coordinates, is g_{zz} = g_{ρρ}, g_{zρ} = 0, similar to the gauge used by Abrahams and Evans. The Hamiltonian constraint is solved at every time step, and the time derivatives of the spatial gauge conditions are substituted into the momentum constraints to obtain secondorder elliptic equations for the two shift components. Thus the evolution is partially constrained. Adaptive meshrefinement was used in the numerical time evolution.
The initial data were either timesymmetric or approximately ingoing, with the scalar field either symmetric or antisymmetric in z. In the symmetric case, even strongly nonspherical data were attracted to the known spherical critical solution for the massless scalar field. Scaling with the known γ was observed in the Ricci scalar. However, with sufficiently good finetuning to the black hole threshold, the approximately spherical region that approaches the critical solution suffers an l = 2 (and by ansatz m = 0) instability and splits into two new spherical regions which again approach the critical solution. The spatial separation of the two new centres is related to the smallest length scale that developed prior to the branching. There is evidence that with increasing finetuning each of these centres splits again. The antisymmetric initial data cannot approach a single spherical critical solution, but the solution splits initially into two approximately spherical regions where the critical solution is approached (up to an overall sign in the scalar field). The separation of these initial two centres is determined by the initial data, but there is evidence that they in turn split.
All this is consistent with the assumption that the spherical critical solution has, besides the known one spherical unstable mode, precisely one further l = 2 unstable mode. (Without the restriction to axisymmetry, if such a mode exists, it would be 5fold degenerate with m = −2, …, 2.) This contradicts the calculation of the perturbation spectrum in [154]. Choptuik and coworkers do not state with certainty that the mode they see in numerical evolutions is a continuum mode, although they have no indication that it is a numerical artifact. The growth rate of the putative mode is measured to be λ ≃ 0.1 −0.4, which should be compared with the growth rate λ ≃ 2.7 of the spherical mode and the relatively small decay rate of λ ≃ −0.02 claimed in [154] for the least damped mode, which is also an l = 2 mode. We observe that the range of τ in Figures 6 and 7 of [54] is about 10, and over this range the plot of the amplitude of the l = 2 perturbation against the logscale coordinate τ seems equally consistent with linear growth in τ as with exponential growth. (In the notation of [54], τ denotes proper time and τ_{*} the accumulation point of echos, so that the logscale coordinate τ used in this review corresponds to − ln(τ − τ_{*}) in the notation of [54].)
An interesting extension was made in [55] by considering a complex scalar field giving rise to an axisymmetric spacetime with angular momentum. The stressenergy tensor of a complex scalar field Ψ is
An axisymmetric spacetime with azimuthal Killing vector ξ admits a conserved vector field T_{ab}ξ^{b}, so that the total angular momentum
is independent of the Cauchy surface Σ. With the ansatz
in adapted coordinates where ξ= ∂/∂_{φ} and with m being an integer, the stressenergy tensor becomes axisymmetric and hence compatible with the Einstein equations for an axisymmetric spacetime.
For any Σ tangent to ξ, in particular a hypersurface t = constant, the angular momentum density measured by a normal observer becomes
where Φ = Ae^{iδ} with δ and A being real and Π = naΦ_{,a}. By comparison the energy density measured by a normal observer is
where D_{a} is the derivative operator projected into Σ. This means that the ratio of energy density to angular momentum density can be adjusted arbitrarily in the initial data, including zero angular momentum for a Φ that is real (up to a constant phase). On the other hand, even in the absence of angular momentum a purely real Φ obeys a wave equation with an explicit m^{2}/ρ^{2} centrifugal term. For the same reason, regular solutions must have Φ ∼ ρ^{m} on the axis, and there are no spherically symmetric solutions. Intuitively speaking, the centrifugal force resisting collapse appears unrelated to the angular momentum component of the stressenergy tensor in a way that differs from what one would expect in rotating fluid collapse or rotating (nonaxisymmetric) vacuum collapse.
For all initial data in numerical evolutions, a critical solution is approached that is discretely selfsimilar with logscale period Δ ≃ 0.42. (By ansatz this solution is axisymmetric but spherical symmetry is ruled out and so the critical solution cannot be the Choptuik solution.) The same critical solution is approached in particular for initial data with \(\Pi = 0\) and hence no angular momentum, and initial data where \(\Pi = \bar \Phi\) and hence with large angular momentum. A scaling exponent of γ ≃ 0.11 is observed in the Ricci scalar in subcritical evolutions. The critical solution is purely real (up to an initial datadependent constant phase) and hence has no angular momentum. Only m = 1 was investigated, but it is plausible that a different critical solution exists for each integer m.
Far from the black hole threshold J ∼ M^{2} in the final black hole, but nearer the black hole threshold, J ∼ M^{6}, where J and M are measured on the apparent horizon when it first forms. J/M^{2} → 0 is compatible with a nonrotating critical solution.
In the absence of angular momentum, the wave equations for the real and imaginary part of Φ decouple. Assuming that the background critical solution is purely real with δ = 0 and A ∼ 1, and angular momentum is provided by a perturbation with δ∼ e^{λτ}, one would expect ρ ∼ ΔA^{2} ∼ e^{2τ} and j ∼ A∇δ∼ e^{(1+λ)τ}. Integrating over a region of size e^{−3τ} when the black hole forms, we find \(M \sim {e^{ {\tau _\ast}}}\) and \(J \sim {e^{( 2 + \lambda){\tau _\ast}}}\). Then J ∼ M^{6} would imply λ = −4.
Olabarrieta et al. [169] study a similar system in spherical symmetry, by arranging 2l + 1 scalar fields given by ϕ_{lm} = ψ(r, t)Y_{lm}(θ, φ) for m = −l, …l with the same ψ(r, t) for all values of m so that the total stressenergy tensor and the spacetime are spherically symmetric. Note that not the ϕ_{lm} are added but their stressenergies, and each value of l describes a different matter content.ψ then sees a centrifugal barrier in its evolution equation, but there is no angular momentum in the spacetime. DSS critical behaviour is found, and the logarithmic echoing period Δ and mass scaling exponent γ both decrease approximately exponentially with l. It is observed empirically that the radius r_{0} of maximum compactness during evolution (a measure of the scale of the initial data) and the accumulation time of echos T* (measured from the initial data) obey T* ≃ r_{0}/(4.35Δ) for all l and families of initial data.
Lai [141] has studied type I critical phenomena for boson (massive complex scalar field) stars in axisymmetry, the first study of type I in axisymmetry. He finds that the subcritical end state is a boson star with a large amplitude fundamental mode oscillation.
Neutron star collision in axisymmetry
A first investigation of type I critical collapse in an astrophysically motivated scenario was carried out in [136, 204]. The matter is a perfect fluid with “Gamma law” EOS P = (γ − 1)ρϵ, where Γ ≃ 2 is a constant, P is the pressure, ρ the rest mass density, and ϵ the internal energy per rest mass (so that ρ(1 + ϵ) is the total energy density). The initial data are constructed with P = kρ^{Γ} for a constant k, which corresponds to the “cold” (constant entropy) limit of the Gamma law EOS. The initial data correspond to two identical stars which have fallen from infinity. (The evolution starts at finite distance, with an initial velocity calculated in the first postNewtonian approximation). The entire solution is axisymmetric with an additional reflection symmetry that maps one star to the other.
The parameter of the initial data that is varied is the mass of the two stars. Supercritical data form a single black hole, while subcritical data form a single star. The diagnostics given are plots against time of the lapse and the Ricci scalar at the symmetry centre of the spacetime, and of outgoing l = 2 gravitational waves. Collapse of the lapse and blowup of the Ricci scalar are taken as indications of black hole formation. The limited numerical evidence is compatible with type I critical phenomena, with the putative critical solution showing oscillations with the same period in the lapse and Ricci scalar. For the critical solution to be exactly timeperiodic, it would have to be spherical in order to not lose energy through gravitational waves, and there is some evidence that indeed it does not radiate gravitational waves.
Other 1parameter families of initial data were obtained by fixing the mass of the stars in the first family very near its critical value and then varying one of the following: the initial separation, the initial speed, and Γ. For all approximately the same scaling law for the survival time of the critical solution was found. However, as all these initial data are very close together, this only confirms the validity of the general perturbation theory explanation of critical phenomena, but does not provide evidence of universality.
A key question that has not been answered is how shock heating affects the standard critical collapse scenario. A priori the existence of a universal spherical critical solution is unlikely in the presence of shock heating as there is no dynamical mechanism to arrive at a universal spherical temperature profile, in contrast to nondissipative matter models or vacuum gravity where all equations are wavelike.
Black hole collisions
Interesting numerical evidence for critical phenomena in the grazing collision of two black holes has been found by Pretorius [179]. He evolved initial data for two equal mass nonrotating black holes, with a reflection symmetry through the orbital plane, parameterised by their relative boost at constant impact parameter. The threshold in initial data space is between data which merge immmediately and those which do not (although they will merge later for initial data which are bound). The critical solution is a circular orbit that loses 1 − 1.5% of the total energy per orbit through radiation. On both sides of the threshold, the number of orbits scales as
for γ ≃ 0.31 −0.38. The simulations are currently limited by numerical accuracy to n < 5 and ΔE/E < 0.08, but Pretorius conjectures that the total energy loss and hence the number of orbits is limited only by the irreducible mass of the initial data, a much larger number. In particular, he speculates that for highly boosted initial data such that the total energy of the initial data is dominated by kinetic energy, almost all the energy can be converted into gravitational radiation.
The standard dynamical systems picture of critical collapse, with the critical solution an attractor in the threshold hypersurface, appears to be consistent with these observation. Pretorius compares his data with the unstable circular geodesics in the spacetime of the hypothetical rotating black hole that would result if merger occurred promptly. These give orbital periods, and their linear perturbations give a critical exponent, in rough agreement with the numerical values for the full black hole collision.
Pretorius does not comment on the nature of the critical solution, but because of the mass loss through gravitational radiation it cannot be strictly stationary. The mass loss would be compatible with selfsimilarity, with a helical homothetic vector field, but exact selfsimilarity would be compatible with the presence of black holes only in the infinite boost limit. Therefore the critical solution is likely to be more complicated, and can perhaps be written as an expansion with an exactly stationary or homothetic spacetime as the leading term.
Pretorius also speculates that these phenomena generalise to generic initial data with unequal masses and black hole spins which are not aligned. This seems uncertain, given the claim by Levin [146] that the threshold of immediate merger is fractal if the spins are not aligned, and that the system is therefore chaotic. However, Levin’s analysis is based on a 2nd order postNewtonian approximation to general relativity, in which there is no radiation reaction, while the rapid energy loss observed here may suppress chaos. Nevertheless, the phase space is much bigger when the orbit is not confined to an orbital plane, and so the critical solution observed here may not be an attractor in the full critical surface.
Sperhake et al. [191] push the numerics towards higher energies emitted and conjecture that the merger of nonspinning black holes can in principle yield a black hole arbitrarily close to extremal Kerr. See also [187].
References
 [1]
Abrahams, A.M., and Evans, C.R., “Critical behavior and scaling in vacuum axisymmetric gravitational collapse”, Phys. Rev. Lett., 70, 2980–2983, (1993). [DOI]. (Cited on pages 7 and 37.)
 [2]
Abrahams, A.M., and Evans, C.R., “Universality in axisymmetric vacuum collapse”, Phys. Rev. D, 49, 3998–4003, (1994). [DOI]. (Cited on page 38.)
 [3]
Aichelburg, P.C., Bizoń, P., and Tabor, Z., “Bifurcation and fine structure phenomena in critical collapse of a selfgravitating σfield”, Class. Quantum Grav., 23, S299–S306, (2006). [DOI], [grqc/0512136]. (Cited on page 27.)
 [4]
Alcubierre, M., Allen, G., Brügmann, B., Lanfermann, G., Seidel, E., Suen, W.M., and Tobias, M., “Gravitational collapse of gravitational waves in 3D numerical relativity”, Phys. Rev. D, 61, 041501, 1–5, (2000). [DOI], [grqc/9904013]. (Cited on page 38.)
 [5]
ÁlvarezGaumé, L., Gómez, C., and VázquezMozo, M.A., “Scaling Phenomena in Gravity from QCD”, Phys. Lett. B, 649, 478–482, (2007). [DOI], [hepth/0611312]. (Cited on pages 29 and 36.)
 [6]
ÁlvarezGaumé, L., Gómez, C., Vera, A.S., Tavanfar, A., and VázquezMozo, M.A., “Critical formation of trapped surfaces in the collision of gravitational shock waves”, J. High Energy Phys.(02), 009, (2009). [DOI], [arXiv:0811.3969 [hepth]]. (Cited on page 36.)
 [7]
ÁlvarezGaumé, L., Gómez, C., Vera, A.S., Tavanfar, A., and VázquezMozo, M.A., “Critical gravitational collapse: Towards a holographic understanding of the Regge region”, Nucl. Phys. B, 806, 327–385, (2009). [DOI], [arXiv:0804.1464 [hepth]]. (Cited on pages 30, 31, and 36.)
 [8]
Andreasson, H., and Rein, G., “A numerical investigation of the stability of steady states and critical phenomena for the spherically symmetric EinsteinVlasov system”, Class. Quantum Grav., 23, 3659–3677, (2006). [DOI], [grqc/0601112]. (Cited on page 34.)
 [9]
Ayal, S., and Piran, T., “Spherical collapse of a massless scalar field with semiclassical corrections”, Phys. Rev. D, 56, 4768–4774, (1997). [DOI], [grqc/9704027]. (Cited on page 36.)
 [10]
Bartnik, R., and McKinnon, J., “Particlelike Solutions of the EinsteinYangMills Equations”, Phys. Rev. Lett., 61, 141–144, (1988). [DOI]. (Cited on pages 28 and 31.)
 [11]
Birmingham, D., “Choptuik scaling and quasinormal modes in the antide Sitter space/conformalfield theory correspondence”, Phys. Rev. D, 64, 064024, 1–5, (2001). [DOI], [hepth/0101194]. (Cited on page 36.)
 [12]
Birmingham, D., and Sen, S., “Gott Time Machines, BTZ Black Hole Formation, and Choptuik Scaling”, Phys. Rev. Lett., 84, 1074–1077, (2000). [DOI], [hepth/9908150]. (Cited on page 35.)
 [13]
Birukou, M., Husain, V., Kunstatter, G., Vaz, E., and Olivier, M., “Spherically symmetric scalar field collapse in any dimension”, Phys. Rev. D, 65, 104036, 1–7, (2002). [DOI], [grqc/0201026]. (Cited on pages 29 and 31.)
 [14]
Bizoó, P., “Colored black holes”, Phys. Rev. Lett., 64, 2844–2847, (1990). [DOI]. (Cited on pages 28 and 31.)
 [15]
Bizoń, P., “How to Make a Tiny Black Hole?”, Acta Cosm., 22, 81, (1996). [grqc/9606060]. (Cited on page 8.)
 [16]
Bizoń, P., “Equivariant SelfSimilar Wave Maps from Minkowski Spacetime into 3Sphere”, Commun. Math. Phys., 215, 45–56, (2000). [DOI], [ADS], [mathph/9910026]. (Cited on page 34.)
 [17]
Bizoń, P., and Chmaj, T., “Formation and critical collapse of Skyrmions”, Phys. Rev. D, 58, 041501, 1–4, (1998). [DOI], [grqc/9801012]. (Cited on page 31.)
 [18]
Bizoń, P., and Chmaj, T., “Remark on formation of colored black holes via finetuning”, Phys. Rev. D, 61, 067501, 1–2, (2000). [DOI], [grqc/9906070]. (Cited on page 31.)
 [19]
Bizoń, P., Chmaj, T., Rostworowski, A., Schmidt, B.G., and Tabor, Z., “On vacuum gravitational collapse in nine dimensions”, Phys. Rev. D, 72, 121502, 1–4, (2005). [DOI], [grqc/0511064]. (Cited on page 29.)
 [20]
Bizoń, P., Chmaj, T., and Schmidt, B.G., “Critical Behavior in Vacuum Gravitational Collapse in 4+1 Dimensions”, Phys. Rev. Lett., 95, 071102, 1–4, (2005). [DOI], [grqc/0506074]. (Cited on page 28.)
 [21]
Bizoń, P., Chmaj, T., and Schmidt, B.G., “CodimensionTwo Critical Behavior in Vacuum Gravitational Collapse”, Phys. Rev. Lett., 97, 131101, 1–4, (2006). [DOI], [grqc/0608102]. (Cited on page 28.)
 [22]
Bizoń, P., Chmaj, T., and Tabor, Z., “Equivalence of critical collapse of nonAbelian fields”, Phys. Rev. D, 59, 104003, 1–3, (1999). [DOI], [grqc/9901039]. (Cited on pages 15 and 31.)
 [23]
Bizoń, P., Chmaj, T., and Tabor, Z., “Dispersion and collapse of wave maps”, Nonlinearity, 13, 1411–1423, (2000). [DOI], [mathph/9912009]. (Cited on page 34.)
 [24]
Bizoń, P., Chmaj, T., and Tabor, Z., “Formation of singularities for equivariant (2+1)dimensional wave maps into the 2sphere”, Nonlinearity, 14, 1041–1053, (2001). [DOI], [mathph/0011005]. (Cited on page 34.)
 [25]
Bizoń, P., Szybka, S.J., and Wasserman, A., “Periodic selfsimilar wave maps coupled to gravity”, Phys. Rev. D, 69, 064014, 1–6, (2004). [DOI], [grqc/0310038]. (Cited on page 27.)
 [26]
Bizoń, P., and Tabor, Z., “On blowup for YangMills fields”, Phys. Rev. D, 64, 121701, 1–4, (2001). [DOI], [mathph/0105016]. (Cited on page 34.)
 [27]
Bizoń, P., and Wasserman, A., “Selfsimilar spherically symmetric wave maps coupled to gravity”, Phys. Rev. D, 62, 084031, 1–7, (2000). [DOI], [grqc/0006034]. (Cited on page 27.)
 [28]
Bizoń, P., and Wasserman, A., “On the existence of selfsimilar spherically symmetric wave maps coupled to gravity”, Class. Quantum Grav., 19, 3309–3321, (2002). [DOI], [grqc/0201046]. (Cited on page 27.)
 [29]
Bland, J., and Kunstatter, G., “The 5D Choptuik critical exponent and holography”, Phys. Rev. D, 75, 101501, 1–4, (2007). [DOI], [hepth/0702226]. (Cited on page 29.)
 [30]
Bland, J., Preston, B., Becker, M., Kunstatter, G., and Husain, V., “Dimension dependence of the critical exponent in spherically symmetric gravitational collapse”, Class. Quantum Grav., 22, 5355–5364, (2005). [DOI], [grqc/0507088]. (Cited on page 29.)
 [31]
Bose, S., Parker, L., and Peleg, Y., “Predictability and semiclassical approximation at the onset of black hole formation”, Phys. Rev. D, 54, 7490–7505, (1996). [DOI], [hepth/9606152]. (Cited on page 36.)
 [32]
Brady, P.R., “Analytic example of critical behaviour in scalar field collapse”, Class. Quantum Grav., 11, 1255–1260, (1994). [DOI], [grqc/9402023]. arXiv preprint title: Does scalar field collapse produce ‘zero mass’ black holes? (Cited on page 35.)
 [33]
Brady, P.R., and Cai, M.J., “Critical phenomena in gravitational collapse”, in Piran, T., ed., The Eighth Marcel Grossmann Meeting on Recent Developments in Theoretical and Experimental General Relativity, Gravitation and Relativistic Field Theories, Proceedings of the meeting held at the Hebrew University of Jerusalem, June 22–27, 1997, pp. 689–704, (World Scientific, Singapore, 1999). [grqc/9812071]. (Cited on page 8.)
 [34]
Brady, P.R., Chambers, C.M., and Goncalves, S.M.C.V., “Phases of massive scalar field collapse”, Phys. Rev. D, 56, R6057–R6061, (1997). [DOI], [ADS], [grqc/9709014]. (Cited on pages 13, 14, 24, and 31.)
 [35]
Brady, P.R., Choptuik, M.W., Gundlach, C., and Neilsen, D.W., “Blackhole threshold solutions in stiff fluid collapse”, Class. Quantum Grav., 19, 6359, (2002). [DOI], [grqc/0207096]. (Cited on page 30.)
 [36]
Brady, P.R., and Ottewill, A.C., “Quantum corrections to critical phenomena in gravitational collapse”, Phys. Rev. D, 58, 024006, 1–6, (1998). [DOI], [grqc/9804058]. (Cited on page 36.)
 [37]
Burko, L.M., “BlackHole Singularities: A New Critical Phenomenon”, Phys. Rev. Lett., 90, 121101, 1–4, (2003). [DOI], [grqc/0209084]. (Cited on page 36.)
 [38]
Cahill, M.E., and Taub, A.H., “Spherically symmetric similarity solutions of the Einstein field equations for a perfect fluid”, Commun. Math. Phys., 21, 1–40, (1971). [DOI]. Related online version (cited on 18 May 2005): http://projecteuclid.org/getRecord?id=euclid.cmp/1103857257. (Cited on pages 9 and 32.)
 [39]
Carlip, S., “The (2+1)dimensional black hole”, Class. Quantum Grav., 12, 2853–2879, (1995). [DOI], [grqc/9506079]. (Cited on page 29.)
 [40]
Carr, B.J., and Coley, A.A., “Complete classification of spherically symmetric selfsimilar perfect fluid solutions”, Phys. Rev. D, 62, 044023, 1–25, (2000). [DOI], [grqc/9901050]. (Cited on page 30.)
 [41]
Carr, B.J., Coley, A.A., Goliath, M., Nilsson, U.S., and Uggla, C., “Critical phenomena and a new class of selfsimilar spherically symmetric perfectfluid solutions”, Phys. Rev. D, 61, 081502, 1–5, (2000). [DOI], [grqc/9901031]. (Cited on page 30.)
 [42]
Carr, B.J., and Gundlach, C., “Spacetime structure of selfsimilar spherically symmetric perfect fluid solutions”, Phys. Rev. D, 67, 024035, 1–13, (2003). [DOI], [grqc/0209092]. (Cited on pages 21 and 30.)
 [43]
Cavaglià, M., Clément, G., and Fabbri, A., “Approximately selfsimilar critical collapse in 2+1 dimensions”, Phys. Rev. D, 70, 044010, 1–5, (2004). [DOI], [grqc/0404033]. (Cited on page 29.)
 [44]
Caveny, S.A., and Matzner, R.A., “Adaptive event horizon tracking and critical phenomena in binary black hole coalescence”, Phys. Rev. D, 68, 104003, 1–13, (2003). [DOI], [grqc/0303109]. (Cited on page 35.)
 [45]
Chiba, T., and Siino, M., “Disappearance of black hole criticality in semiclassical general relativity”, Mod. Phys. Lett. A, 12, 709–718, (1997). [DOI], [ADS]. (Cited on page 36.)
 [46]
Choptuik, M.W., personal communication. (Cited on page 31.)
 [47]
Choptuik, M.W., “‘Critical’ behavior in massless scalar field collapse”, in d’Inverno, R.A., ed., Approaches to Numerical Relativity, Proceedings of the International Workshop on Numerical Relativity, Southampton, December 1991, p. 202, (Cambridge University Press, Cambridge; New York, 1992). [ADS]. (Cited on pages 15, 18, and 31.)
 [48]
Choptuik, M.W., “Universality and scaling in gravitational collapse of a massless scalar field”, Phys. Rev. Lett., 70, 9–12, (1993). [DOI], [ADS]. (Cited on pages 7, 18, and 31.)
 [49]
Choptuik, M.W., “Critical behavior in scalar field collapse”, in Hobill, D., Burd, A., and Coley, A., eds., Deterministic Chaos in General Relativity, Proceedings of a NATO Advanced Research Workshop on Deterministic Chaos in General Relativity, held July 25–30, 1993, in Kananaskis, Alberta, Canada, p. 155, (Plenum Press, New York, 1994). (Cited on pages 14, 15, 18, 19, and 31.)
 [50]
Choptuik, M.W., “The (Unstable) Threshold of Black Hole Formation”, in Dadhich, N., and Narlikar, J.V., eds., Gravitation and Relativity: At the Turn of the Millenium, Proceedings of the 15th International Conference on General Relativity and Gravitation (GR15), held at IUCAA, Pune, India, December 16–21, 1997, pp. 67–85, (IUCAA, Pune, 1998). [grqc/9803075]. (Cited on page 8.)
 [51]
Choptuik, M.W., “Critical behavior in gravitational collapse”, Prog. Theor. Phys. Suppl., 136, 353–365, (1999). [DOI]. (Cited on page 8.)
 [52]
Choptuik, M.W., Chmaj, T., and Bizoń, P., “Critical Behavior in Gravitational Collapse of a YangMills Field”, Phys. Rev. Lett., 77, 424–427, (1996). [DOI], [grqc/9603051]. (Cited on pages 14, 28, and 31.)
 [53]
Choptuik, M.W., Hirschmann, E.W., and Liebling, S.L., “Instability of an ‘approximate black hole’”, Phys. Rev. D, 55, 6014–6018, (1997). [DOI], [grqc/9701011]. (Cited on pages 30 and 31.)
 [54]
Choptuik, M.W., Hirschmann, E.W., Liebling, S.L., and Pretorius, F., “Critical collapse of the massless scalar field in axisymmetry”, Phys. Rev. D, 68, 044007, 1–9, (2003). [DOI], [grqc/0305003]. (Cited on pages 25, 38, and 39.)
 [55]
Choptuik, M.W., Hirschmann, E.W., Liebling, S.L., and Pretorius, F., “Critical Collapse ofa Complex Scalar Field with Angular Momentum”, Phys. Rev. Lett., 93, 131101, 1–4, (2004). [DOI], [grqc/0405101]. (Cited on page 39.)
 [56]
Choptuik, M.W., Hirschmann, E.W., and Marsa, R.L., “New critical behavior in EinsteinYangMills collapse”, Phys. Rev. D, 60, 124011, 1–9, (1999). [DOI], [grqc/9903081]. (Cited on pages 28 and 31.)
 [57]
Christodoulou, D., “Violation of cosmic censorship in the gravitational collapse of a dust cloud”, Commun. Math. Phys., 93, 171–195, (1984). [DOI]. Related online version (cited on 18 May 2005): http://projecteuclid.org/getRecord?id=euclid.cmp/1103941053. (Cited on page 18.)
 [58]
Christodoulou, D., “The problem of a selfgravitating scalar field”, Commun. Math. Phys., 105, 337–361, (1986). [DOI]. Related online version (cited on 18 May 2005): http://projecteuclid.org/getRecord?id=euclid.cmp/1104115427. (Cited on pages 18 and 19.)
 [59]
Christodoulou, D., “A mathematical theory of gravitational collapse”, Commun. Math. Phys., 109, 613–647, (1987). [DOI], [ADS]. (Cited on page 18.)
 [60]
Christodoulou, D., “The formation of black holes and singularities in spherically symmetric gravitational collapse”, Commun. Pure Appl. Math., 44, 339–373, (1991). [DOI]. (Cited on pages 18 and 19.)
 [61]
Christodoulou, D., “Bounded Variation Solutions of the Spherically Symmetric EinsteinScalar Field Equations”, Commun. Pure Appl. Math., 46, 1131–1220, (1993). [DOI]. (Cited on pages 18 and 19.)
 [62]
Christodoulou, D., “Examples of Naked Singularity Formation in the Gravitational Collapse of a Scalar Field”, Ann. Math. (2), 140, 607–653, (1994). [DOI]. (Cited on page 18.)
 [63]
Christodoulou, D., “The Instability of Naked Singularities in the Gravitational Collapse of a Scalar Field”, Ann. Math. (2), 149, 183–217, (1999). [DOI]. (Cited on pages 18 and 21.)
 [64]
Clément, G., and Fabbri, A., “Analytical treatment of critical collapse in (2+1)dimensional AdS spacetime: a toy model”, Class. Quantum Grav., 18, 3665–3680, (2001). [DOI], [grqc/0101073]. (Cited on pages 29 and 35.)
 [65]
Clément, G., and Fabbri, A., “Critical collapse in (2+1)dimensional AdS spacetime: quasiCSS solutions and linear perturbations”, Nucl. Phys. B, 630, 269–292, (2002). [DOI], [grqc/0109002]. (Cited on pages 29 and 35.)
 [66]
Clément, G., and Hayward, S.A., “Comment on ‘An extreme critical spacetime: echoing and blackhole perturbations’”, Class. Quantum Grav., 18, 4715–4716, (2001). [DOI], [grqc/0108024]. (Cited on page 35.)
 [67]
Donninger, R., and Aichelburg, P.C., “A note on the eigenvalues for equivariant maps of the SU(2) sigmamodel”, arXiv eprint, (2006). [mathph/0601019]. (Cited on page 34.)
 [68]
Donninger, R., and Aichelburg, P.C., “On the mode stability of a selfsimilar wave map”, J. Math. Phys., 49, 043515, (2008). [DOI], [mathph/0702025]. (Cited on page 34.)
 [69]
Eardley, D.M., Hirschmann, E.W., and Horne, J.H., “S duality at the black hole threshold in gravitational collapse”, Phys. Rev. D, 52, R5397–R5401, (1995). [DOI], [grqc/9505041]. (Cited on pages 27 and 31.)
 [70]
Eardley, D.M., and Moncrief, V., “The Global Existence of YangMillsHiggs Fields in 4Dimensional Minkowski Space. I. Local Existence and Smoothness Properties”, Commun. Math. Phys., 83, 171–191, (1982). [DOI]. (Cited on page 34.)
 [71]
Eggers, J., and Fontelos, M.A., “The role of selfsimilarity in singularities of partial differential equations”, Nonlinearity, 22, R1–R44, (2009). [DOI], [arXiv:0812.1339 [mathph]]. (Cited on page 8.)
 [72]
Evans, C.R., and Coleman, J.S., “Critical Phenomena and SelfSimilarity in the Gravitational Collapse of Radiation Fluid”, Phys. Rev. Lett., 72, 1782–1785, (1994). [DOI], [grqc/9402041]. (Cited on pages 30 and 31.)
 [73]
Frolov, A.V., “Perturbations and critical behavior in the selfsimilar gravitational collapse of a massless scalar field”, Phys. Rev. D, 56, 6433–6438, (1997). [DOI], [grqc/9704040]. (Cited on page 35.)
 [74]
Frolov, A.V., “Critical collapse beyond spherical symmetry: General perturbations of the Roberts solution”, Phys. Rev. D, 59, 104011, 1–7, (1999). [DOI], [grqc/9811001]. (Cited on page 35.)
 [75]
Frolov, A.V., “Selfsimilar collapse of scalar field in higher dimensions”, Class. Quantum Grav., 16, 407–417, (1999). [DOI], [grqc/9806112]. (Cited on page 35.)
 [76]
Frolov, A.V., “Continuous selfsimilarity breaking in critical collapse”, Phys. Rev. D, 61, 084006, 1–14, (2000). [DOI], [grqc/9908046]. (Cited on page 35.)
 [77]
Frolov, A.V., and Pen, U.L., “The naked singularity in the global structure of critical collapse spacetimes”, Phys. Rev. D, 68, 124024, 1–6, (2003). [DOI], [grqc/0307081]. (Cited on pages 18 and 22.)
 [78]
Frolov, V.P., “Merger transitions in braneblackhole systems: Criticality, scaling and selfsimilarity”, Phys. Rev. D, 74, 044006, 1–9, (2006). [DOI], [grqc/0604114]. (Cited on page 35.)
 [79]
Frolov, V.P., Larsen, A.L., and Christensen, M., “Domain wall interacting with a black hole: A new example of critical phenomena”, Phys. Rev. D, 59, 125008, 1–8, (1999). [DOI], [hepth/9811148]. (Cited on page 35.)
 [80]
Garfinkle, D., “Choptuik scaling in null coordinates”, Phys. Rev. D, 51, 5558–5561, (1995). [DOI], [ADS], [grqc/9412008]. (Cited on pages 18 and 22.)
 [81]
Garfinkle, D., “Choptuik scaling and the scale invariance of Einstein’s equation”, Phys. Rev. D, 56, 3169–3173, (1997). [DOI], [grqc/9612015]. (Cited on page 14.)
 [82]
Garfinkle, D., “Exact solution for (2+1)dimensional critical collapse”, Phys. Rev. D, 63, 044007, 1–5, (2001). [DOI], [grqc/0008023]. (Cited on page 29.)
 [83]
Garfinkle, D., Cutler, C., and Duncan, G.C., “Choptuik scaling in six dimensions”, Phys. Rev. D, 60, 104007, 1–5, (1999). [DOI], [ADS], [grqc/9908044]. (Cited on pages 29 and 31.)
 [84]
Garfinkle, D., and Duncan, G.C., “Scaling of curvature in subcritical gravitational collapse”, Phys. Rev. D, 58, 064024, 1–4, (1998). [DOI], [grqc/9802061]. (Cited on page 22.)
 [85]
Garfinkle, D., and Gundlach, C., “Symmetryseeking spacetime coordinates”, Class. Quantum Grav., 16, 4111–4123, (1999). [DOI], [grqc/9908016]. (Cited on page 14.)
 [86]
Garfinkle, D., and Gundlach, C., “Perturbations of an exact solution for 2+1dimensional critical collapse”, Phys. Rev. D, 66, 044015, 1–4, (2002). [DOI], [grqc/0205107]. (Cited on page 29.)
 [87]
Garfinkle, D., Gundlach, C., and MartínGarcía, J.M., “Angular momentum near the black hole threshold in scalar field collapse”, Phys. Rev. D, 59, 104012, 1–5, (1999). [DOI], [grqc/9811004]. (Cited on pages 25 and 37.)
 [88]
Garfinkle, D., and Isenberg, J., “Numerical studies of the behavior of Ricci flow”, in Chang, S.C., Chow, B., Chu, S.C., and Lin, C.S., eds., Geometric Evolution Equations, National Center for Theoretical Sciences Workshop on Geometric Evolution Equations, National Tsing Hua University, Hsinchu, Taiwan, July 15–August 14, 2002, Contemporary Mathematics, vol. 367, p. 103, (American Mathematical Society, Providence, RI, 2004). [math/0306129]. (Cited on page 34.)
 [89]
Garfinkle, D., and Isenberg, J., “The modelling of degenerate neck pinch singularities in Ricci flow by Bryant solitons”, J. Math. Phys., 49, 073505, (2007). [DOI], [arXiv:0709.0514[math.DG]]. (Cited on page 34.)
 [90]
Garfinkle, D., Mann, R., and Vuille, C., “Critical collapse of a massive vector field”, Phys. Rev. D, 68, 064015, 1–6, (2003). [DOI], [grqc/0305014]. (Cited on pages 30 and 31.)
 [91]
Garfinkle, D., and Meyer, K., “Scale invariance and critical gravitational collapse”, Phys. Rev. D, 59, 064003, 1–5, (1999). [DOI], [grqc/9806052]. (Cited on page 14.)
 [92]
Giddings, S.B., “Quantum mechanics of black holes”, arXiv eprint, (1994). [hepth/9412138]. Lectures presented at the 1994 Trieste Summer School in High Energy Physics and Cosmology. (Cited on page 36.)
 [93]
Goode, S.W., Coley, A.A., and Wainwright, J., “The isotropic singularity in cosmology”, Class. Quantum Grav., 9, 445–455, (1992). [DOI]. (Cited on page 20.)
 [94]
Green, A.M., and Liddle, A.R., “Critical collapse and the primordial black hole initial mass function”, Phys. Rev. D, 60, 063509, 1–8, (1999). [DOI], [astroph/9901268v2]. (Cited on page 33.)
 [95]
Gundlach, C., “The Choptuik Spacetime as an Eigenvalue Problem”, Phys. Rev. Lett., 75, 3214–3217, (1995). [DOI], [grqc/9507054]. (Cited on pages 12 and 31.)
 [96]
Gundlach, C., “Critical phenomena in gravitational collapse”, in Chruiściel, P.T., ed., Mathematics of Gravitation, Part I: Lorentzian Geometry and Einstein Equations, Proceedings of the Workshop on Mathematical Aspects of Theories of Gravitation, held in Warsaw, Poland, February 29–March 30, 1996, Banach Center Publications, vol. 41, pp. 143–152, (Polish Academy of Sciences, Institute of Mathematics, Warsaw, 1997). [grqc/9606023]. (Cited on page 8.)
 [97]
Gundlach, C., “Echoing and scaling in EinsteinYangMills critical collapse”, Phys. Rev. D, 55, 6002–6013, (1997). [DOI], [grqc/9610069]. (Cited on pages 15, 28, and 31.)
 [98]
Gundlach, C., “Understanding critical collapse of a scalar field”, Phys. Rev. D, 55, 695–713, (1997). [DOI], [grqc/9604019]. (Cited on pages 13 and 31.)
 [99]
Gundlach, C., “Angular momentum at the black hole threshold”, Phys. Rev. D, 57, 7080–7083, (1998). [DOI], [grqc/9711079]. (Cited on page 32.)
 [100]
Gundlach, C., “Critical phenomena in gravitational collapse”, Adv. Theor. Math. Phys., 2, 1–49, (1998). [grqc/9712084]. (Cited on page 8.)
 [101]
Gundlach, C., “Nonspherical perturbations of critical collapse and cosmic censorship”, Phys. Rev. D, 57, 7075–7079, (1998). [DOI], [grqc/9710066]. (Cited on pages 31 and 32.)
 [102]
Gundlach, C., “Critical Phenomena in Gravitational Collapse”, Living Rev. Relativity, 2, lrr19994, (1999). URL (cited on 1 February 2003): http://www.livingreviews.org/lrr19994. (Cited on pages 7 and 8.)
 [103]
Gundlach, C., “Critical gravitational collapse of a perfect fluid: Nonspherical perturbations”, Phys. Rev. D, 65, 084021, 1–22, (2002). [DOI], [grqc/9906124]. (Cited on pages 31 and 32.)
 [104]
Gundlach, C., “Critical gravitational collapse with angular momentum: from critical exponents to universal scaling functions”, Phys. Rev. D, 65, 064019, (2002). [DOI], [grqc/0110049]. (Cited on pages 17 and 37.)
 [105]
Gundlach, C., “Critical phenomena in gravitational collapse”, Phys. Rep., 376, 339–405, (2003). [DOI], [grqc/0210101]. (Cited on page 8.)
 [106]
Gundlach, C., and MartínGarcía, J.M., “Charge scaling and universality in critical collapse”, Phys. Rev. D, 54, 7353–7360, (1996). [DOI], [grqc/9606072]. (Cited on pages 15, 24, and 31.)
 [107]
Gundlach, C., and MartínGarcía, J.M., “Kinematics of discretely selfsimilar spherically symmetric spacetimes”, Phys. Rev. D, 68, 064019, 1–11, (2003). [DOI], [grqc/0306001]. (Cited on page 20.)
 [108]
Gundlach, C., and MartínGarcía, J.M., “Wellposedness of formulations of the Einstein equations with dynamical lapse and shift conditions”, Phys. Rev. D, 74, 024016, 1–19, (2006). [DOI], [grqc/0604035]. (Cited on page 38.)
 [109]
Gundlach, C., Price, R.H., and Pullin, J., “Latetime behavior of stellar collapse and explosions. II. Nonlinear evolution”, Phys. Rev. D, 49, 890–899, (1994). [DOI], [ADS], [grqc/9307010]. (Cited on page 18.)
 [110]
Hamadé, R.S., Horne, J.H., and Stewart, J.M., “Continuous selfsimilarity and Sduality”, Class. Quantum Grav., 13, 2241–2253, (1996). [DOI], [ADS], [grqc/9511024]. (Cited on pages 27 and 31.)
 [111]
Hamadé, R.S., and Stewart, J.M., “The spherically symmetric collapse of a massless scalar field”, Class. Quantum Grav., 13, 497–512, (1996). [DOI], [ADS], [grqc/9506044]. (Cited on pages 18 and 22.)
 [112]
Hara, T., Koike, T., and Adachi, S., “Renormalization group and critical behavior in gravitational collapse”, arXiv eprint, (1996). [grqc/9607010]. (Cited on pages 15 and 31.)
 [113]
Harada, T., “Final fate of the spherically symmetric collapse of a perfect fluid”, Phys. Rev. D, 58, 104015, 1–10, (1998). [DOI], [grqc/9807038]. (Cited on page 32.)
 [114]
Harada, T., “Stability criterion for selfsimilar solutions with perfect fluids in general relativity”, Class. Quantum Grav., 18, 4549–4567, (2001). [DOI], [grqc/0109042]. (Cited on page 30.)
 [115]
Harada, T., and Maeda, H., “Convergence to a selfsimilar solution in general relativistic gravitational collapse”, Phys. Rev. D, 63, 084022, 1–14, (2001). [DOI], [grqc/0101064]. (Cited on pages 30 and 32.)
 [116]
Harada, T., and Maeda, H., “Critical phenomena in Newtonian gravity”, Phys. Rev. D, 64, 124024, 1–7, (2001). [DOI], [grqc/0109095]. (Cited on page 33.)
 [117]
Harada, T., Maeda, H., and Semelin, B., “Criticality and convergence in Newtonian collapse”, Phys. Rev. D, 67, 084003, 1–10, (2003). [DOI], [grqc/0210027]. (Cited on page 33.)
 [118]
Harada, T., and Mahajan, A., “Analytical solutions for blackhole critical behaviour”, Gen. Relativ. Gravit., 39, 1847–1854, (2007). [DOI], [arXiv:0707.3000 [grqc]]. (Cited on page 35.)
 [119]
Hawke, I., and Stewart, J.M., “The dynamics of primordial blackhole formation”, Class. Quantum Grav., 19, 3687–3707, (2002). [DOI]. (Cited on pages 22 and 33.)
 [120]
Hawley, S.H., and Choptuik, M.W., “Boson stars driven to the brink of black hole formation”, Phys. Rev. D, 62, 104024, 1–19, (2000). [DOI], [grqc/0007039]. (Cited on pages 24 and 31.)
 [121]
Hayward, S.A., “An extreme critical spacetime: echoing and blackhole perturbations”, Class. Quantum Grav., 17, 4021–4030, (2000). [DOI], [grqc/0004038]. (Cited on page 35.)
 [122]
Hirschmann, E.W., and Eardley, D.M., “Critical exponents and stability at the black hole threshold for a complex scalar field”, Phys. Rev. D, 52, 5850–5856, (1995). [DOI], [grqc/9506078]. (Cited on page 27.)
 [123]
Hirschmann, E.W., and Eardley, D.M., “Universal scaling and echoing in gravitational collapse of a complex scalar field”, Phys. Rev. D, 51, 4198–4207, (1995). [DOI], [grqc/9412066]. (Cited on page 27.)
 [124]
Hirschmann, E.W., and Eardley, D.M., “Criticality and bifurcation in the gravitational collapse of a selfcoupled scalar field”, Phys. Rev. D, 56, 4696–4705, (1997). [DOI], [grqc/9511052]. (Cited on pages 26 and 31.)
 [125]
Hirschmann, E.W., Wang, A., and Wu, Y., “Collapse of a Scalar Field in 2+1 Gravity”, Class. Quantum Grav., 21, 1791–1824, (2004). [DOI], [grqc/0207121]. (Cited on page 29.)
 [126]
Hod, S., and Piran, T., “Critical behavior and universality in gravitational collapse of a charged scalar field”, Phys. Rev. D, 55, 3485–3496, (1997). [DOI], [ADS], [grqc/9606093]. (Cited on pages 24 and 31.)
 [127]
Hod, S., and Piran, T., “Finestructure of Choptuik’s massscaling relation”, Phys. Rev. D, 55, 440–442, (1997). [DOI], [grqc/9606087]. (Cited on page 13.)
 [128]
Honda, E.P., and Choptuik, M.W., “Fine structure of oscillons in the spherically symmetric ϕ^{4} KleinGordon model”, Phys. Rev. D, 65, 084037, 1–12, (2002). [DOI], [hepph/0110065]. (Cited on page 24.)
 [129]
Horne, J.H., “Critical behavior in black hole collapse”, Matters of Gravity (7), 14–15, (1996). [grqc/9602001]. (Cited on page 8.)
 [130]
Horowitz, G.T., and Hubeny, V.E., “Quasinormal modes of AdS black holes and the approach to thermal equilibrium”, Phys. Rev. D, 62, 024027, 1–11, (2000). [DOI], [hepth/9909056]. (Cited on page 36.)
 [131]
Husa, S., Lechner, C., Pürrer, M., Thornburg, J., and Aichelburg, P.C., “Type II critical collapse of a selfgravitating nonlinear σ model”, Phys. Rev. D, 62, 104007, 1–11, (2000). [DOI], [ADS], [grqc/0002067]. (Cited on page 27.)
 [132]
Husain, V., “Critical Behaviour in Quantum Gravitational Collapse”, Adv. Sci. Lett., 2, 214–220, (2009). [DOI], [arXiv:0808.0949 [grqc]]. (Cited on page 36.)
 [133]
Husain, V., Kunstatter, G., Preston, B., and Birukou, M., “Antide Sitter gravitational collapse”, Class. Quantum Grav., 20, L23–L29, (2003). [DOI], [grqc/0210011]. (Cited on page 29.)
 [134]
Husain, V., and Olivier, M., “Scalar field collapse in threedimensional AdS spacetime”, Class. Quantum Grav., 18, L1–L9, (2001). [DOI], [grqc/0008060]. (Cited on page 29.)
 [135]
Husain, V., and Seahra, S.S., “Ricci flows, wormholes and critical phenomena”, Class. Quantum Grav., 25, 222002, (2008). [DOI], [arXiv:0808.0880 [grqc]]. (Cited on page 34.)
 [136]
Jin, K.J., and Suen, W.M., “Critical Phenomena in HeadOn Collisions of Neutron Stars”, Phys. Rev. Lett., 98, 131101, 1–4, (2007). [DOI], [grqc/0603094]. (Cited on page 40.)
 [137]
Kiem, Y., “Phase Transition in Spherically Symmetric Gravitational Collapse of a Massless Scalar Field”, arXiv eprint, (1994). [hepth/9407100]. (Cited on pages 35 and 36.)
 [138]
Koike, T., Hara, T., and Adachi, S., “Critical Behavior in Gravitational Collapse of Radiation Fluid: A Renormalization Group (Linear Perturbation) Analysis”, Phys. Rev. Lett., 74, 5170–5173, (1995). [DOI], [grqc/9503007]. (Cited on page 30.)
 [139]
Koike, T., Hara, T., and Adachi, S., “Critical behavior in gravitational collapse of a perfect fluid”, Phys. Rev. D, 59, 104008, 1–9, (1999). (Cited on page 31.)
 [140]
Kol, B., “Choptuik Scaling and The Merger Transition”, J. High Energy Phys., 2006(10), 017, 1–18, (2006). [DOI], [hepth/0502033]. (Cited on page 29.)
 [141]
Lai, C.W., A Numerical Study of Boson Stars, Ph.D. Thesis, (University of British Columbia, Vancouver, 2004). [grqc/0410040]. Related online version (cited on 15 June 2007): http://laplace.physics.ubc.ca/People/matt/Doc/Theses/. (Cited on pages 24 and 40.)
 [142]
Lai, C.W., and Choptuik, M.W., “Final Fate ofSubcritical Evolutions of Boson Stars”, arXiv eprint, (2007). [arXiv:0709.0324]. (Cited on page 24.)
 [143]
Lavrelashvili, G., and Maison, D., “A remark on the instability of the BartnikMcKinnon solutions”, Phys. Lett. B, 343, 214–217, (1995). [DOI]. (Cited on page 31.)
 [144]
Lechner, C., Staticity, selfsimilarity and critical phenomena in a selfgravitating nonlinear sigma model, Ph.D. Thesis, (University of Vienna, Vienna, 2001). [grqc/0507009]. (Cited on page 27.)
 [145]
Lechner, C., Thornburg, J., Husa, S., and Aichelburg, P.C., “A new transition between discrete and continuous selfsimilarity in critical gravitational collapse”, Phys. Rev. D, 65, 081501, 1–4, (2002). [DOI], [grqc/0112008]. (Cited on page 27.)
 [146]
Levin, J., “Gravity Waves, Chaos, and Spinning Compact Binaries”, Phys. Rev. Lett., 84, 3515–3518, (2000). [DOI], [grqc/9910040]. (Cited on page 41.)
 [147]
Liebling, S.L., “Multiply unstable black hole critical solutions”, Phys. Rev. D, 58, 084015, 1–8, (1998). [DOI], [grqc/9805043]. (Cited on page 31.)
 [148]
Liebling, S.L., “Critical phenomena inside global monopoles”, Phys. Rev. D, 60, 061502, 1–5, (1999). [DOI], [grqc/9904077]. (Cited on page 31.)
 [149]
Liebling, S.L., “Singularity threshold of the nonlinear sigma model using 3D adaptive mesh refinement”, Phys. Rev. D, 66, 041703, 1–5, (2002). [DOI], [grqc/0202093]. (Cited on page 34.)
 [150]
Liebling, S.L., and Choptuik, M.W., “Black Hole Criticality in the BransDicke Model”, Phys. Rev. Lett., 77, 1424–1427, (1996). [DOI], [grqc/9606057]. (Cited on pages 27 and 31.)
 [151]
Liebling, S.L., Hirschmann, E.W., and Isenberg, J.A., “Critical phenomena in nonlinear sigma models”, J. Math. Phys., 41(8), 5691–5700, (2000). [DOI], [mathph/9911020]. (Cited on page 34.)
 [152]
Mahajan, A., Harada, T., Joshi, P., and Nakao, K., “Critical Collapse of Einstein Cluster”, Prog. Theor. Phys., 118, 865–878, (2007). [DOI], [arXiv:0710.4315 [grqc]]. (Cited on page 35.)
 [153]
Maison, D., “Nonuniversality of critical behaviour in spherically symmetric gravitational collapse”, Phys. Lett. B, 366, 82–84, (1996). [DOI], [grqc/9504008]. (Cited on pages 30 and 31.)
 [154]
MartínGarcía, J.M., and Gundlach, C., “All nonspherical perturbations of the Choptuik spacetime decay”, Phys. Rev. D, 59, 064031, 1–19, (1999). [DOI], [grqc/9809059]. (Cited on pages 24, 25, 31, and 39.)
 [155]
MartínGarcía, J.M., and Gundlach, C., “Selfsimilar spherically symmetric solutions of the massless EinsteinVlasov system”, Phys. Rev. D, 65, 084026, 1–18, (2002). [DOI], [grqc/0112009]. (Cited on pages 31 and 34.)
 [156]
MartínGarcía, J.M., and Gundlach, C., “Critical Phenomena in Gravitational Collapse: The Role of Angular Momentum”, in FernándezJambrina, L., and GonzálezRomero, L.M., eds., Current Trends in Relativistic Astrophysics: Theoretical, Numerical, Observational, Proceedings of the 24th Spanish Relativity Meeting on Relativistic Astrophysics, Madrid, 2001, Lecture Notes in Physics, vol. 617, pp. 68–86, (Springer, Berlin; New York, 2003). (Cited on page 8.)
 [157]
MartínGarcía, J.M., and Gundlach, C., “Global structure of Choptuik’s critical solution in scalar field collapse”, Phys. Rev. D, 68, 024011, 1–25, (2003). [DOI], [grqc/0304070]. (Cited on pages 20 and 21.)
 [158]
Millward, R.S., and Hirschmann, E.W., “Critical behavior of gravitating sphalerons”, Phys. Rev. D, 68, 024017, 1–12, (2003). [DOI], [grqc/0212015]. (Cited on pages 28 and 31.)
 [159]
Musco, I., Miller, J.C., and Polnarev, A.G., “Primordial black hole formation in the radiative era: investigation of the critical nature of the collapse”, Class. Quantum Grav., 26, 235001, (2009). [DOI], [arXiv:0811.1452 [grqc]]. (Cited on page 33.)
 [160]
Musco, I., Miller, J.C., and Rezzolla, L., “Computations of primordial blackhole formation”, Class. Quantum Grav., 22, 1405–1424, (2005). [DOI], [grqc/0412063]. (Cited on pages 22 and 33.)
 [161]
Neilsen, D.W., and Choptuik, M.W., “Critical phenomena in perfect fluids”, Class. Quantum Grav., 17, 761–782, (2000). [DOI], [grqc/9812053]. (Cited on pages 15, 30, 31, and 32.)
 [162]
Neilsen, D.W., and Choptuik, M.W., “Ultrarelativistic fluid dynamics”, Class. Quantum Grav., 17, 733–759, (2000). [DOI], [grqc/9904052]. (Cited on page 30.)
 [163]
Niemeyer, J.C., and Jedamzik, K., “NearCritical Gravitational Collapse and the Initial Mass Function of Primordial Black Holes”, Phys. Rev. Lett., 80, 5481–5484, (1998). [DOI], [grqc/9709072]. (Cited on page 33.)
 [164]
Niemeyer, J.C., and Jedamzik, K., “Dynamics of primordial black hole formation”, Phys. Rev. D, 59, 124013, 1–8, (1999). [DOI], [astroph/9901292]. (Cited on page 33.)
 [165]
Noble, S.C., A Numerical Study of Relativistic Fluid Collapse, Ph.D. Thesis, (University of Texas at Austin, Austin, 2003). [grqc/0310116]. (Cited on page 32.)
 [166]
Noble, S.C., and Choptuik, M.W., “Type II critical phenomena of neutron star collapse”, Phys. Rev. D, 78, 064059, (2008). [DOI], [arXiv:0709.3527]. (Cited on page 32.)
 [167]
Novak, J., “Velocityinduced collapses of stable neutron stars”, Astron. Astrophys., 376, 606–613, (2001). [DOI], [grqc/0107045]. (Cited on page 32.)
 [168]
Olabarrieta, I., and Choptuik, M.W., “Critical phenomena at the threshold of black hole formation for collisionless matter in spherical symmetry”, Phys. Rev. D, 65, 024007, 1–10, (2001). [DOI], [grqc/0107076]. (Cited on pages 31 and 33.)
 [169]
Olabarrieta, I., Ventrella, J.F., Choptuik, M.W., and Unruh, W.G., “Critical behavior in the gravitational collapse of a scalar field with angular momentum in spherical symmetry”, Phys. Rev. D, 76, 124014, (2007). [DOI], [arXiv:0708.0513 [grqc]]. (Cited on page 40.)
 [170]
OliveiraNeto, G., and Takakura, F.I., “Wyman’s solution, selfsimilarity, and critical behaviour”, J. Math. Phys., 46, 062503, 1–6, (2005). [DOI], [grqc/0309099]. (Cited on page 35.)
 [171]
Ori, A., and Piran, T., “Naked singularities in selfsimilar spherical gravitational collapse”, Phys. Rev. Lett., 59, 2137–2140, (1987). [DOI]. (Cited on page 30.)
 [172]
Ori, A., and Piran, T., “Naked singularities and other features of selfsimilar generalrelativistic gravitational collapse”, Phys. Rev. D, 42, 1068–1090, (1990). [DOI]. (Cited on pages 30 and 32.)
 [173]
Oshiro, Y., Nakamura, K., and Tomimatsu, A., “Critical behavior of black hole formation in a scalar wave Collapse”, Prog. Theor. Phys., 91, 1265–1270, (1994). [DOI], [grqc/9402017]. (Cited on page 35.)
 [174]
Peleg, Y., Bose, S., and Parker, L., “Choptuik scaling and quantum effects in 2D dilaton gravity”, Phys. Rev. D, 55, 4525–4528, (1997). [DOI], [grqc/9608040]. (Cited on page 36.)
 [175]
Peleg, Y., and Steif, A.R., “Phase transition for gravitationally collapsing dust shells in 2+1 dimensions”, Phys. Rev. D, 51, R3992–R3996, (1995). [DOI], [grqc/9412023]. (Cited on page 35.)
 [176]
Petryk, R., MaxwellKleinGordon Fields in Black Hole Spacetimes, Ph.D. Thesis, (University of British Columbia, Vancouver, 2005). Related online version (cited on 15 June 2007): http://laplace.physics.ubc.ca/People/matt/Doc/Theses/. (Cited on page 24.)
 [177]
Polnarev, A.G., and Musco, I., “Curvature profiles as initial conditions for primordial black hole formation”, Class. Quantum Grav., 24, 1405–1432, (2007). [DOI], [grqc/0605122]. (Cited on page 33.)
 [178]
Pretorius, F., and Choptuik, M.W., “Gravitational collapse in 2+1 dimensional AdS spacetime”, Phys. Rev. D, 62, 124012, 1–15, (2000). [DOI], [grqc/0007008]. (Cited on page 29.)
 [179]
Pretorius, F., and Khurana, D., “Black hole mergers and unstable circular orbits”, Class. Quantum Grav., 24, S83–S108, (2007). [DOI], [grqc/0702084]. (Cited on page 41.)
 [180]
Price, R.H., and Pullin, J., “Analytic approximations to the spacetime of a critical gravitational collapse”, Phys. Rev. D, 54, 3792–3799, (1996). [DOI], [grqc/9601009]. (Cited on page 35.)
 [181]
Pullin, J., “Is there a connection between nohair behavior and universality in gravitational collapse?”, Phys. Lett. A, 204, 7–10, (1995). [DOI], [grqc/9409044]. (Cited on page 35.)
 [182]
Pürrer, M., Husa, S., and Aichelburg, P.C., “News from critical collapse: Bondi mass, tails and quasinormal modes”, Phys. Rev. D, 71, 10400, 1–13, (2005). [DOI], [grqc/0411078]. (Cited on pages 18 and 22.)
 [183]
Rein, G., Rendall, A.D., and Schaeffer, J., “Critical collapse of collisionless matter: A numerical investigation”, Phys. Rev. D, 58, 044007, 1–8, (1998). [DOI], [grqc/9804040]. (Cited on pages 31 and 33.)
 [184]
Roberts, M.D., “Scalar field counterexamples to the cosmic censorship hypothesis”, Gen. Relativ. Gravit., 21, 907–939, (1989). [DOI]. (Cited on page 35.)
 [185]
Sarbach, O., and Lehner, L., “Critical bubbles and implications for critical black strings”, Phys. Rev. D, 71, 026002, 1–11, (2005). [DOI], [hepth/0407265]. (Cited on page 30.)
 [186]
Seidel, E., and Suen, W.M., “Oscillating soliton stars”, Phys. Rev. Lett., 66, 1659–1662, (1991). [DOI]. (Cited on pages 24 and 31.)
 [187]
Shibata, M., Okawa, H., and Yamamoto, T., “Highvelocity collision of two black holes”, Phys. Rev. D, 78, 101501(R), (2008). [DOI], [arXiv:0810.4735 [grqc]]. (Cited on page 42.)
 [188]
Smarr, L.L., and York Jr, J.W., “Kinematical conditions in the construction of spacetime”, Phys. Rev. D, 17, 2529–2551, (1978). [DOI]. (Cited on page 14.)
 [189]
Snajdr, M., “Critical collapse of an ultrarelativistic fluid in the Γ → 1 limit”, Class. Quantum Grav., 23, 3333–3352, (2006). [DOI], [grqc/0508062]. (Cited on page 33.)
 [190]
Sorkin, E., and Oren, Y., “Choptuik’s scaling in higher dimensions”, Phys. Rev. D, 71, 124005, (2005). [DOI], [hepth/0502034]. (Cited on page 29.)
 [191]
Sperhake, U., Cardoso, V., Pretorius, F., Berti, E., Hinderer, T., and Yunes, N., “Cross Section, Final Spin, and ZoomWhirl Behavior in HighEnergy BlackHole Collisions”, Phys. Rev. Lett., 103, 131102, (2009). [DOI], [arXiv:0907.1252 [grqc]]. (Cited on page 42.)
 [192]
Stevenson, R., The Spherically Symmetric Collapse of Collisionless Matter: Exploring Critical Phenomena through Finite Volume Methods, Masters Thesis, (University of British Columbia, Vancouver, 2005). Related online version (cited on 15 June 2007): http://laplace.physics.ubc.ca/Theses/. (Cited on page 33.)
 [193]
Straumann, N., and Zhou, Z.H., “Instability of a colored black hole solution”, Phys. Lett. B, 243, 33–35, (1990). [DOI]. (Cited on page 31.)
 [194]
Strominger, A., and Thorlacius, L., “Universality and scaling at the onset of quantum black hole formation”, Phys. Rev. Lett., 72, 1584–1587, (1994). [DOI], [hepth/9312017]. (Cited on page 36.)
 [195]
Szybka, S.J., “Chaotic selfsimilar wave maps coupled to gravity”, Phys. Rev. D, 69, 084014, 1–7, (2004). [DOI], [grqc/0310050]. (Cited on page 27.)
 [196]
Szybka, S.J., and Chmaj, T., “Fractal Threshold Behavior in Vacuum Gravitational Collapse”, Phys. Rev. Lett., 100, 101102, (2008). [DOI], [arXiv:0711.4612 [grqc]]. (Cited on page 29.)
 [197]
Thornburg, J., Lechner, C., Pürrer, M., Aichelburg, P.C., and Husa, S., “Type II Critical Collapse of a SelfGravitating Nonlinear σ Model”, in Gurzadyan, V.G., Jantzen, R.T., and Ruffini, R., eds., The Ninth Marcel Grossmann Meeting on recent developments in theoretical and experimental general relativity, gravitation and relativistic field theories, Proceedings of the MGIX MM meeting held at the University of Rome ‘La Sapienza’, July 2–8, 2000, pp. 1645–1646, (World Scientific, Singapore; River Edge, NJ, 2001). [grqc/0012043]. (Cited on page 27.)
 [198]
Tod, P., personal communication. (Cited on page 20.)
 [199]
van Putten, M.H.P.M., “Approximate black holes for numerical relativity”, Phys. Rev. D, 54, R5931–R5934, (1996). [DOI], [grqc/9607074]. (Cited on pages 30 and 31.)
 [200]
Ventrella, J.F., and Choptuik, M.W., “Critical phenomena in the EinsteinmasslessDirac system”, Phys. Rev. D, 68, 044020, 1–10, (2003). [DOI], [grqc/0304007]. (Cited on pages 30 and 31.)
 [201]
Volkov, M.S., Brodbeck, O., Lavrelashvili, G., and Straumann, N., “The number of sphaleron instabilities of the BartnikMcKinnon solitons and nonAbelian black holes”, Phys. Lett. B, 349, 438–442, (1995). [DOI], [hepth/9502045]. (Cited on page 31.)
 [202]
Volkov, M.S., and Gal’tsov, D.V., “NonAbelian EinsteinYangMills Black Holes”, J. Exp. Theor. Phys. Lett., 50, 346–350, (1989). (Cited on pages 28 and 31.)
 [203]
Wald, R.M., “Gravitational Collapse and Cosmic Censorship”, arXiv eprint, (1997). [grqc/9710068]. (Cited on page 21.)
 [204]
Wan, M.B., Jin, K.J., and Suen, W.M., “Dynamical analysis of the structure of neutron star critical collapses”, Poster version presented at the ‘2nd Course of the International School on Astrophysical Relativity’, Erice, Italy, June 27–July 5, 2008, conference paper, (2008). [arXiv:0807.1710 [grqc]]. (Cited on page 40.)
 [205]
Wang, A., “Critical collapse of cylindrically symmetric scalar field in fourdimensional Einstein’s theory of gravity”, Phys. Rev. D, 68, 064006, 1–12, (2003). [DOI], [grqc/0307071]. (Cited on page 36.)
 [206]
Wang, A., and de Oliveira, H.P., “Critical phenomena of collapsing massless scalar wave packets”, Phys. Rev. D, 56, 753–761, (1997). [grqc/9608063]. (Cited on page 35.)
 [207]
Wyman, M., “Static spherically symmetric scalar fields in general relativity”, Phys. Rev. D, 24, 839–841, (1981). [DOI]. (Cited on page 35.)
 [208]
Yeomans, J.M., Statistical Mechanics of Phase Transitions, Oxford Science Publications, (Clarendon Press; Oxford University Press, Oxford; New York, 1992). [Google Books]. (Cited on page 15.)
 [209]
Yokoyama, J., “Cosmological constraints on primordial black holes produced in the nearcritical gravitational collapse”, Phys. Rev. D, 58, 107502, (1998). [grqc/9804041]. (Cited on page 33.)
 [210]
Zhou, J.G., MüllerKirsten, H.J.W., and Yang, M.Z., “New look at the critical behaviour near the threshold of black hole formation in the RussoSusskindThorlacius model”, Phys. Rev. D, 51, R314–R318, (1995). (Cited on page 36.)
 [211]
Ziprick, J., and Kunstatter, G., “Dynamical singularity resolution in spherically symmetric black hole formation”, Phys. Rev. D, 80, 024032, (2009). [DOI], [arXiv:0902.3224 [grqc]]. (Cited on page 36.)
 [212]
Ziprick, J., and Kunstatter, G., “Spherically symmetric black hole formation in PainlevéGullstrand coordinates”, Phys. Rev. D, 79, 101503(R), (2009). [DOI], [arXiv:0812.0993 [grqc]]. (Cited on page 18.)
Acknowledgements
We would like to thank David Garfinkle for a critical reading of the manuscript and the Louisiana State University for hospitality while this work was begun. JMM was supported by the I3P framework of CSIC and the European Social Fund, and by the Spanish MEC Project FIS200505736C0302.
For the 2010 revision, JMM was supported by the French A.N.R. (Agence Nationale de la Recherche) through Grant BLAN071_201699 entitled “LISA Science”, and CG was supported by the A.N.R. Grant 062134423 entitled “Mathematical Methods in General Relativity”. CG would also like to thank the IAP and LUTH for hospitality.
Author information
Affiliations
Corresponding author
Rights and permissions
Open Access This article is distributed under the terms of the Creative Commons Attribution 4.0 International License (https://creativecommons.org/licenses/by/4.0), which permits use, duplication, adaptation, distribution, and reproduction in any medium or format, as long as you give appropriate credit to the original author(s) and the source, provide a link to the Creative Commons license, and indicate if changes were made.
About this article
Cite this article
Gundlach, C., MartínGarcía, J.M. Critical Phenomena in Gravitational Collapse. Living Rev. Relativ. 10, 5 (2007). https://doi.org/10.12942/lrr20075
Accepted:
Published:
Keywords
 Black Hole
 Scalar Field
 Spherical Symmetry
 Critical Phenomenon
 Apparent Horizon
Latest
Critical Phenomena in Gravitational Collapse Published:
 11 December 2007
 Accepted:
 06 December 2007
DOI: https://doi.org/10.12942/lrr20075
Original
Critical Phenomena in Gravitational Collapse Published:
 22 December 1999
DOI: https://doi.org/10.12942/lrr19994