Nonsingular spacetimes with a negative cosmological constant: IV. Stationary black hole solutions with matter fields
Abstract.
We use an elliptic system of equations with complex coefficients for a set of complexvalued tensor fields as a tool to construct infinitedimensional families of nonsingular stationary black holes, realvalued Lorentzian solutions of the EinsteinMaxwelldilatonscalar fieldsYangMillsHiggsChernSimons equations with a negative cosmological constant. The families include an infinitedimensional family of solutions with the usual AdS conformal structure at conformal infinity.
Contents
1. Introduction
There is currently considerable interest in the literature in spacetimes with a negative cosmological constant. This is fueled on one hand by studies of the AdSCFT conjecture and of the implications thereof: Indeed, this problem is of immediate physical interest in the context of the weakly coupled supergravity limit of the AdS/CFT correspondence. On the other hand, these solutions are interesting because of a rich dynamical morphology: existence of periodic or quasiperiodic solutions, and of instabilities. All this leads naturally to the question of existence of stationary solutions of the Einstein equations with , with or without sources, and of properties thereof.
This manuscript is the fourth in a series of papers, starting with [13], which are devoted to proving existence of a large class of solutions to the Einstein equations with negative cosmological constant by perturbation of known ones. All of these papers are further related by the fact that the field equations can be transposed to an elliptic system on a conformally compact Riemannian manifold. The system is solved by an implicit function theorem argument under a non degeneracy hypothesis. This can be traced back to earlier work of Graham and Lee [20] on constructing Einstein metrics on dimensional balls with boundary, as generalized to more general infinities by Lee [25] (compare [7] for more general symmetric spaces). Using such methods, in [14, 15] we have constructed infinite dimensional families of nonsingular strictly stationary space times, solutions of the Einstein equations with a negative cosmological constant and with various matter sources. These families include an infinitedimensional family of solutions with the usual AdS conformal structure at conformal infinity. The construction there did not provide any black hole solutions, as strict stationarity is incompatible with existence of horizons. However, black hole solutions are of special interest. In fact, various such solutions have already been constructed numerically: For example, static EinsteinYangMills black holes have been constructed in spacetime dimension five in [29], with fourdimensional solutions constructed in [9], and higher dimensional ones in [28]. In [27] an explicit fivedimensional such solution has been given. Rotating EinsteinMaxwellChernSimons solutions have been presented in [10]. In [19] a family of fivedimensional black holes was constructed satisfying the Einsteincomplex scalar field equations, with a stationary geometry and timeperiodic scalar field; compare [5].
The object of this work is to provide a rigorous existence proof for large families of such solutions. The idea is to use a “Wick rotation” to construct suitable solutions of a system of elliptic equations with complex coefficients for a complex valued “Riemannian metric”. In a nutshell, we show that Lee’s theorem on existence of perturbed PoincaréEinstein Riemannian metrics [25, Theorem A] can be extended to complex valued “metrics”, and to more general equations, and that this can be used to construct stationary Lorentzian black hole solutions with large classes of matter sources. This proceeds as follows:
We wish to construct a Lorentzian metric in any spacedimension , with Killing vector , satisfying the EinsteinMaxwellChernSimonsYangMillsdilatonscalar fields equations, with a stationary geometry but possibly timeperiodic complex fields, or the equations. In adapted coordinates the metric can be written as
(1.1)  
(1.2) 
Let us introduce a complex parameter and consider the complexvalued tensor field
(1.3) 
satisfying (1.2). We will say that a complex valued symmetric tensor field is a complex metric if is symmetric and invertible. Replacing by in (1.3), where , we obtain a complex metric with Riemannian real part:
(1.4) 
Under such a substitution the field equations transform in a controlled way, for example independent vacuum metrics lead to independent, possibly complex valued, tensor fields satisfying the vacuum equations, etc.
Working near a (realvalued) static Einstein metric satisfying a nondegeneracy condition (as defined in the paragraph after Equation (3.5) below) we will

construct complex metrics reminiscent of (1.4) which solve the vacuum Einstein equations for small , and

show that , and are realvalued if .
(Incidentally, we will also show that , and hence , is analytic in , an interesting property of the stationary metrics at hand which has does not seem to have been noticed so far.) After “undoing the Wick rotation” leading from (1.3) to (1.4), we will show that the resulting Lorentzian spacetime has a smooth event horizon at suitable zeros of .
The construction guarantees that has zeros when did, and leads indeed to the desired Lorentzian blackhole solution of the Einstein, or Einsteinmatter equations.
Our nondegeneracy condition is satisfied by large classes of metrics, including all fourdimensional Kottler metrics except the spherical ones with a single critical value of the mass parameter[16] (see also [4, Proposition D.2]). It is clear that the method opens further possibilities, which remain to be explored. For example, the technique is used in [17] to construct boson star solutions.
2. Elliptic equations with complex principal symbol
Consider an dimensional (real) manifold . Complexvalued tensor fields over are defined as sections of the usual (real) tensor bundles over tensored with . In other words, all coordinate transformations are real but we allow tensors to have complex components. We emphasise that the “Wick rotation” above is not considered to be a coordinate transformation, but a useful device mapping one set of equations and fields to another, more convenient, one.
As already mentioned, we will say that a twocovariant complex valued tensor is a complex metric if is symmetric and invertible.
Let , , be a collection of complex valued fields, forming a section of a complex bundle over . Let be a complex metric and consider a collection of equations of the form
(2.1) 
with some functions which will be assumed to depend smoothly upon their arguments. This can be rewritten as the following collection of equations for real fields :
(2.2)  
(2.3) 
We will say that (2.1) is elliptic if the system (2.2)(2.3) is elliptic in the usual sense for PDEs involving realvalued functions. The principal symbol of (2.2)(2.3) is blocdiagonal, built out of blocs of the form
(2.4) 
This is an isomorphism for if and only if
(2.5) 
Hence, if is positivedefinite then (2.1) will be elliptic regardless of . More importantly for us, when is small enough all the usual elliptic estimates, as needed for our analysis below, apply to (2.2)(2.3), and hence to (2.1). Likewise, isomorphism properties for a realvalued carry over to nearby complexvalued ’s. As we will be using an implicit function theorem around real valued Riemannian metrics, our perturbation of , as well as will always be sufficiently small for the estimates and the isomorphism properties to remain valid.
3. The setup
We work in spacetime dimension and we normalise the cosmological constant to
(3.1) 
this can always be achieved by a constant rescaling of the metric.
Let denote the covariant derivative associated with the metric , set
(3.2) 
(In (3.2) the derivative is of course understood as a covariant derivative operator acting on tensor densities.) Denoting by the Riemann tensor of , similarly for the Ricci tensor, we set
(3.3) 
Then the linearisation with respect to the metric, at , in dimension , of the map
is the operator
(3.4) 
where the Lichnerowicz Laplacian acts on symmetric twotensor fields as
(3.5) 
We will say that a metric is nondegenerate if has no kernel. This should not be confused with the notion of nondegenerate black holes, also called extreme black holes, which is the requirement of nonzero surface gravity.
Large classes of nondegenerate Einstein metrics are described in [25, 4, 2, 1], see also Remark 4.10 below.
It follows immediately from the openness of the set of invertible operators that if is a realvalued nondegenerate Riemannian metric, then all nearby (in a suitable topology, as determined by the problem at hand) complex valued metrics will also be nondegenerate.
The following is wellknown (cf., e.g., the proof of theorem A at the end of [25], compare [20] for the Poincaré ball):
Proposition 3.1.
Suppose that is nondegenerate and that , for a tensor field satisfying
(3.6) 
as the conformal boundary is approached. Then .
Our solutions will be perturbations of a spacetime with a static metric solving the vacuum Einstein equations with a negative cosmological constant. By definition of staticity, near every point in at which the Killing vector is timelike there exist coordinates in which the metric takes the form (1.1)(1.2) with ,
(3.7) 
The solutions we are about to construct will be defined in the domain of outer communications, where the representation (3.7) is in fact global.
In this work we will consider two cases:

is strictly positive, is diffeomorphic to , where the coordinate along the factor labels the static slices of in . We set
thus time translations in become rotations of the factor of .

We allow to describe a static vacuum blackhole metric with a Killing horizon with nonzero surface gravity and with global structure similar to that of the domain of outer communications in the Schwarzschildanti de Sitter (SAdS) black holes. More precisely, we assume that the Lorentzian manifold takes the form
for some , where is a compact dimensional boundaryless manifold. We require that coincides with the zerolevel set of which, in a suitable extension of , becomes an event horizon with nonzero surface gravity. The coordinate along the factor labels the static slices of . We further assume that after a “Wick rotation”, where is replaced by , the resulting Riemannian metric
(3.8) extends to a smooth metric on
with the action of the flow of the vector field
(3.9) being rotations of the factor.
An example of H2 above is given by the dimensional Schwarzschildanti de Sitter metrics with nonvanishing surface gravity, where is the dimensional sphere and . More generally, the dimensional Birmingham metrics [8], where is an dimensional Einstein manifold, with nonextreme horizons are of this form.
4. The construction
To avoid a discussion of the technicalities associated with the matter fields, we will start by describing in some detail the construction of the vacuum solutions. Note, however, that the argument is essentially the same in both cases, once the isomorphisms needed to handle matter fields have been established. The key difference is in the boundary conditions: the vacuum stationary solutions are determined by their asymptotic data at the conformal boundary, and might have a nonstandard conformal infinity when these data are not the usual AdS ones.^{1}^{1}1Note that some nontrivial asymptotic data are compatible with the usual locally conformally flat structure of the conformal boundary. An example is provided by the DemiańskiCarter “Kerr antide Sitter” solutions, see [22, Appendix B]. On the other hand, our nonvacuum solutions are determined by both the asymptotic data for matter fields and for the metric, which allows existence of nontrivial solutions with the manifestly standard AdS conformal structure at timelike infinity.
4.1. Vacuum solutions
We denote by a coordinate near which vanishes at , and by the space of oneforms on of differentiability class.
We have the following:
Theorem 4.1.
Let , , , and consider a static Lorentzian realvalued Einstein metric of the form (1.1)(1.2) as described in Section 3, such that the associated Riemannian metric is compactifiable and nondegenerate, with smooth conformal infinity. We further assume that the hypotheses H1 or H2 of Section 3 hold. For all with small enough and every smooth realvalued there exists a unique, modulo diffeomorphisms which are the identity at the boundary, nearby stationary Lorentzian realvalued vacuum metric of the form (1.1)(1.2) such that, in local coordinates near the conformal boundary ,
(4.1) 
The Lorentzian solutions with (case H1) are globally stationary, in the sense that they have a globally timelike Killing vector. We show in Section 5 below that the Lorentzian solutions with vanishing (case H2) describe smooth black holes.
Remark 4.2.
Remark 4.3.
The decay rates in (4.1) have to be compared with the leading order behavior both for and in local coordinates near the conformal boundary. A precise version of (4.1) in terms of weighted function spaces reads (our notation for function spaces follows [25])
(4.2)  
(4.3) 
and the norms of the differences above are small in those spaces. If the boundary data are smooth, then the solution has a complete polyhomogeneous expansion at the conformal boundary.
Proof: We start by solving on the “harmonicallyreduced Riemannian Einstein equations”,
(4.4) 
for a complexvalued tensorfield , with the asymptotic conditions
(4.5) 
Here we have extended from to by imposing invariance under rotations of the factor.
The existence of a solution, for all with small enough, follows by rewriting the equations as in (2.2)(2.3) (with ), and applying the implicit function theorem. This can be done because of our hypothesis of nondegeneracy of ; see [4, 13] for the analytical details. In particular (4.2)(4.3) hold.
The implicit function theorem guarantees that the solutions sufficiently close to with the asymptotics (4.5) are uniquely determined by . We denote by this solution.
The usual argument, spelledout in detail e.g. in [15, Section 4], applies to show that , so that:
Lemma 4.4.
The complex metrics solve the Riemannian Einstein equations. ∎
We continue by showing that:
Lemma 4.5.
The complex metrics are invariant under rotations of the factor of in the case H1, or of the factor in the case H2.
Proof.
Let us denote by the operator obtained by linearising (4.4) at ; compare (3.4). The Lie derivative of (4.4) with respect to gives
(4.6) 
where is the Liederivative with respect to the vector field generating rotations of the factor of in the case H1, or of the factor in the case H2; we have also used the fact that . It follows from (4.2)(4.3) and polyhomogeneity of the solutions that
(4.7)  
(4.8) 
This, together with Proposition 3.1, implies , as desired.
Denoting by the usual angular coordinate on the factor (H2 case), or the parameter along (H1 case), we can thus write the metrics in coordinates adapted to the flow of in the form
(4.9) 
Lemma 4.6.
In coordinates as in (4.9), the functions , and are even functions of .
Proof.
Let denote the map which, in the coordinates of (4.9) changes to its negative, leaving the remaining coordinates unchanged. Then is a smooth isometry of . The metric satisfies the same equation, with same asymptotic data, as , and is close to for sufficiently small, so that uniqueness gives
(4.10) 
which implies the claim.
Lemma 4.7.
The metrics are holomorphic functions of .
Proof.
It is standard to show that the metrics are continuously differentiable functions of . Differentiating (4.4) with respect to gives
where is the usual complexderivative operator with respect to the complex conjugate of . The vanishing of the asymptotic data for gives .
Now, if , we can repeat the above construction in the space of realvalued Riemannian metrics. Uniqueness implies then that the corresponding metrics are realvalued. Hence all the coefficients in the convergent Taylor expansions
(4.11) 
are real. Lemma 4.6 implies that , the ’s, and the ’s are real for real . It follows that for real the realvalued Lorentzian metrics
(4.12) 
satisfy all our claims.
4.2. Matter fields
We now seek solutions to the EinsteinYangMillsHiggsMaxwelldilatonscalar fieldsChernSimons equations defined by the action
(4.13) 
Here is the Ricci scalar of the metric , and are smooth functions, is the gaugeinvariant norm of a possibly nonAbelian YangMills field, is allowed to be a section of a bundle associated to the possibly nonAbelian gaugegroup, with depending if desired upon the YangMills gauge potential. Finally, in even space dimension , is the ChernSimons action which, in the Abelian case, takes the form:
(4.14) 
for a constant . In the general (nonAbelian) case is given by [11, Equation (3.5)]
(4.15) 
when and
(4.16) 
We obtain:
Theorem 4.8.
Let , , , and consider a static Lorentzian realvalued Einstein metric of the form (1.1)(1.2) as described in Section 3, such that the associated Riemannian metric is compactifiable and nondegenerate, with smooth conformal infinity, has no harmonic oneforms which are in , with which is not an eigenvalue of the operator . We further assume that the hypotheses H1 or H2 of Section 3 hold and that
(4.17) 
For all with small enough, every smooth realvalued and and

with , and (where ) which are sufficiently small smooth fields on , or

, and which is a sufficiently small smooth field on ,
there exists a unique, modulo diffeomorphisms which are the identity at the boundary, nearby stationary Lorentzian solution of the EinsteinMaxwelldilatonscalar fieldsChernSimons equations, or of the YangMillsHiggsChernSimonsdilaton equations with a trivial principal bundle, so that, in local coordinates near , we have
(4.18) 
with all convergences in norm. The hypothesis of nonexistence of harmonic oneforms is not needed if , in which case the Maxwell field or the YangMills field are identically zero.
Remark 4.9.
The remarks in [15, Section 7] concerning the energy and the asymptotics of the solutions remain valid wordforword in the current setting.
Remark 4.10.
For the convenience of the reader we repeat here the comments from [15] concerning the kernel conditions in the theorem.
First, it is shown in [15, Appendix C] that the condition of nonexistence of harmonic forms is satisfied near antide Sitter spacetime in any case.
Next, it has been shown by Lee [24, Theorem A] that there are no eigenvalues of when the Yamabe invariant of the conformal infinity is positive, in particular near antide Sitter and Schwarzschild antide Sitter spacetime. Furthermore, and quite generally, is never an eigenvalue by the maximum principle. Finally, again quite generally, the spectrum of for asymptotically hyperbolic manifolds is together with possibly a finite set of eigenvalues, with finite multiplicity, between and [21] (compare [26]), so our noneigenvalue condition is true except for at most a finite number of values of for all asymptotic geometries. ∎
Proof.
This follows directly from the arguments of [15]: The indicial exponents of the relevant equations remain unchanged, as terms containing , which are here multiplied by , are of lower order in . Note that the solutions obtained below using the implicit function theorem might a priori depend upon the “periodic time coordinate” , but this is irrelevant for the calculation of the indicial exponents.
We start by sketching the argument in the case of a single realvalued scalar field , which satisfies the equation
Its indicial exponents are , unchanged from those in [15], as so that terms arising from derivatives are of lower order. We assume so that the solutions show the desired asymptotics. By [15, Theorem D.1], using the assumption that is not an eigenvalue of , it follows that the linearisation is an isomorphism from to for small .
Then by the same argument as in the proof of Lemma 4.5 above: Applying to the equations and using gives
(4.19) 
Here is a linear combination of the energymomentum tensor of the scalar field and its trace times the metric, and is at least quadratic in its arguments and their derivatives, so that is a linear firstorder differential operator in . Furthermore, each term in is linear in or its derivatives and contains an derivative of one of the arguments. behaves asymptotically as in [15], i.e. and, using , we have . As the coefficients of and , and of their first derivatives, on the righthand side of (4.19) are small in the relevant spaces (e.g. the coefficients of the terms on the righthand side of the second equation are small in ), we can use Proposition 3.1 and the isomorphism properties of described above to conclude that and .
Next, we show that are even functions of : satisfy the same equations as , with identical asymptotic data (since is independent of , the only relevant terms are the asymptotic data for which are unchanged under ) and by uniqueness we have , similarly for , , .
Therefore , , , are real for real by the same argument as for , , above.
Rather similar considerations apply for the whole system of EinsteinMaxwelldilatonYangMillsHiggsChernSimonsscalar fields equations: The matter equations arising from the action (4.13) are
(4.20) 
where^{2}^{2}2One can check by a direct timeandspace decomposition of the equations that the “Wick rotation” , , is consistent with the ChernSimons terms in the equations by defining as , where is totally antisymmetric with values in , with the cut in the definition of , , lying e.g. on the positive imaginary axis, so that both near and near .
(4.21) 
After replacing by , the asymptotic data for the Riemannian solution, say , take the form . They are clearly invariant under
(4.22)  and . 
The only other asymptotic data that are possibly affected by (4.22) are those associated with the inverse metric components . These change sign under each of and . It follows
which again implies that are even functions of . As before, analyticity holds and we conclude that all these fields are real for real .
Note that the implicit function theorem in the Riemannian regime produces essentially complex electric fields for real and nonzero ’s, which will however be mapped to real ones when one returns to the Lorentzian setting. ∎
4.3. theories
Our method allows the construction of blackhole solutions to specific theories: As described in e.g. [18, Section 2.3] these can be reduced to the Einsteinscalar field equations with a specific potential by a conformal transformation, if the function fulfills certain conditions. These conditions are satisfied simultaneously with our assumptions on (in Theorem 4.8) if
(4.23) 
(this is shown in detail in [15, Section 5.5]). An example of a function which fulfills these conditions is
(4.24) 
where is the Ricci scalar in the theory (i.e. before the conformal transformation) and , , , and
are constants.
4.4. Timeperiodic scalar fields
Similarly to [15, Section 6.1] we can use the method there to construct solutions with a timeperiodic complex scalar field where is a constant and is allowed to be complex. We assume that
(4.25)  and 
for some differentiable functions and , and replace the term in the action by , where is the complex conjugate of .
The Lorentzian equation for a complex scalar field takes the form
(4.26) 
This leads to the following associated Riemannian equation
(4.27)  
where the crucial difference to a naive replacement (and therefore ) is that the argument of and is instead of . The equations (4.26) and (4.27), together with the respective Lorentzian and Riemannian equations for the other variables, are equivalent: The bijection
(4.28) 
maps Lorentzian solutions to Riemannian ones.
As such, the first equation (4.27) does not make sense for periodic ’s, but the second does. Note, however, that and are singular at an axis of rotation of , if there is one. This forces us to restrict ourselves to strictly stationary configurations, without black holes, when . As a consequence, in this section we merely reproduce the results already proved in [15] for rotating complex fields, albeit by a somewhat simpler argument.
Applying to the second equation in (4.27) gives
and therefore
(4.29) 
where each term in is at least linear in its arguments or their derivatives and contains an derivative of , , , or .
We can now argue as before to obtain if the asymptotic data are invariant under .
The equation for is then
(4.30) 
All terms in this equation are well defined and, by the results of [15], we obtain a solution of the complete system of equations. As is independent of , no difficulties associated with the periodicity of the coordinate arise. After transforming back via the inverse of (4.28) we obtain a timeperiodic solution to the original equations.
5. Geometry of the solutions
We wish to show that the solutions constructed above with topology correspond to smooth black holes on the Lorentzian side. (In fact, the Lorentzian metric will be onesidedanalytic uptohorizon [6] near the horizon, but this is irrelevant for the problem here.) For this, we recall some standard facts about isometries. Let us denote by
(5.1) 
the codimensiontwo submanifold of which is the zeroset of the Killing vector generating rotations of . Then is a totallygeodesic submanifold of . In coordinates normal for the metric , on each of the planes –orthogonal to the vector field takes the standard Euclidean form
This shows that in these coordinates a rotation by an angle , which is the map , is an isometry of which leaves invariant . Let us choose local coordinates on , and extend them to be constant along –geodesics –normal to . We will denote by the coordinates . One obtains
In particular all oddorder derivatives of the metric functions and vanish on .
An analogous argument applies to using .
Let us assume for definiteness that , thus is purely real. It is then standard to derive the following form of the metric in coordinates , where (compare [12, Section 3] for detailed calculations in a closely related setting):
(5.2)  
(5.3)  
(5.4) 
and where all the nonexplicit functions are smooth functions of .
Passing to the Lorentzian regime, and replacing by a coordinate
(5.5) 
one checks that the Lorentzian metric smoothly extends to a Killing horizon at after a final change of coordinates . Indeed, the Lorentzian metric is then given by
(5.6) 
after substituting in (5.2) and applying the coordinate transformation (5.5).
Indeed, (5.6) shows that that the Killing vector is null on the hypersurface , and that this hypersurface is null, hence a Killing horizon. This is a nonrotating horizon, in the sense that the Killing vector which is timelike at infinity is also tangent to the Killing horizon. (This explains why our solutions, which can have no further symmetries than stationarity, are compatible with the HollandsIshibashiWald [23] rigidity theorem, which provides at least one more symmetry for rotating horizons.) We also see from (5.6) that the surface gravity of the Killing horizon , calculated for the vector field , equals one. Rescaling to the scale of the original nearby seed Birmingham solution, the surface gravity of our solutions will coincide with that of the seed metric in those cases with matter sources where the gravitational free data at infinity have been chosen to coincide with the original ones; otherwise a nearby surface gravity will result when the asymptotic behaviour of the metric imposes a natural rescaling of the horizon Killing vector field.
Acknowledgements The research of PTC was supported in part by the Austrian Science Fund (FWF), Projects P23719N16 and P29517N27, and by the Polish National Center of Science (NCN) under grant 2016/21/B/ST1/00940. PK was supported by a uni:docs grant of the University of Vienna. We are grateful to the Erwin Schrödinger Institute for hospitality and support during part of work on this paper.
References
 [1] M.T. Anderson, Einstein metrics with prescribed conformal infinity on manifolds, Geom. Funct. Anal. 18 (2001), 305–366, arXiv:math.DG/0105243. MR 2421542
 [2] by same author, Boundary regularity, uniqueness and nonuniqueness for AH Einstein metrics on manifolds, Adv. in Math. 179 (2003), 205–249, arXiv:math.DG/0104171. MR 2010802
 [3] M.T. Anderson, P.T. Chruściel, and E. Delay, Nontrivial, static, geodesically complete vacuum spacetimes with a negative cosmological constant, Jour. High Energy Phys. 10 (2002), 063, 22 pp., arXiv:grqc/0211006. MR 1951922
 [4] by same author, Nontrivial, static, geodesically complete spacetimes with a negative cosmological constant. II. , AdS/CFT correspondence: Einstein metrics and their conformal boundaries, IRMA Lect. Math. Theor. Phys., vol. 8, Eur. Math. Soc., Zürich, 2005, arXiv:grqc/0401081, pp. 165–204. MR MR2160871
 [5] D. Astefanesei and E. Radu, Boson stars with negative cosmological constant, Nucl. Phys. B 665 (2003), 594–622, arXiv:grqc/0309131. MR 2000918
 [6] R. Beig and P.T. Chruściel, On analyticity of stationary vacuum metrics at nondegenerate Killing horizons, in preparation.
 [7] O. Biquard, Métriques d’Einstein asymptotiquement symétriques (Asymptotically symmetric Einstein metrics), Astérisque 265, Paris: Société Mathématique de France, 109 pp. , 2000.
 [8] D. Birmingham, Topological black holes in antide Sitter space, Class. Quantum Grav. 16 (1999), 1197–1205, arXiv:hepth/9808032. MR MR1696149 (2000c:83062)
 [9] J. Bjoraker and Y. Hosotani, Monopoles, dyons and black holes in the fourdimensional EinsteinYangMills theory, Phys. Rev. D62 (2000), 043513, arXiv:hepth/0002098.
 [10] J.L. BlázquezSalcedo, J. Kunz, F. NavarroLérida, and E. Radu, Static EinsteinMaxwell Magnetic Solitons and Black Holes in an Odd Dimensional AdS Spacetime, Entropy 18 (2016), 438, arXiv:1612.03747 [grqc].
 [11] S.S. Chern and J. Simons, Characteristic forms and geometric invariants, Ann. of Math. (2) 99 (1974), 48–69. MR 0353327
 [12] P.T. Chruściel, On analyticity of static vacuum metrics at nondegenerate horizons, Acta Phys. Pol. B36 (2005), 17–26, arXiv:grqc/0402087.
 [13] P.T. Chruściel and E. Delay, Nonsingular, vacuum, stationary spacetimes with a negative cosmological constant, Ann. Henri Poincaré 8 (2007), 219–239. MR MR2314449
 [14] P.T. Chruściel and E. Delay, Nonsingular spacetimes with a negative cosmological constant: II. Static solutions of the EinsteinMaxwell equations, (2016), arXiv:1612.00281 [math.DG].
 [15] P.T. Chruściel, E. Delay, and P. Klinger, Nonsingular spacetimes with a negative cosmological constant: III. Stationary solutions with matter fields, (2017).
 [16] by same author, On nondegeneracy of Riemannian Schwarzschildanti de Sitter metrics, (2017), arXiv:1710.07597 [grqc].
 [17] P.T. Chruściel, E. Delay, P. Klinger, A. Kriegl, P.W. Michor, and A. Rainer, Nonsingular spacetimes with a negative cosmological constant: V. Boson stars, (2017).
 [18] Antonio De Felice and Shinji Tsujikawa, f(r) theories, Living Reviews in Relativity 13 (2010), no. 1, 3.
 [19] O.J.C. Dias, G.T. Horowitz, and J.E. Santos, Black holes with only one Killing field, Jour. High Energy Phys. (2011), 115, 43, arXiv:1105.4167 [hepth]. MR 2875937
 [20] C.R. Graham and J.M. Lee, Einstein metrics with prescribed conformal infinity on the ball, Adv. Math. 87 (1991), 186–225.
 [21] C. Guillarmou, Meromorphic properties of the resolvent on asymptotically hyperbolic manifolds, Duke Math. Jour. 129 (2005), 1–37. MR 2153454
 [22] M. Henneaux and C. Teitelboim, Asymptotically anti–de Sitter spaces, Commun. Math. Phys. 98 (1985), 391–424. MR 86f:83030
 [23] S. Hollands, A. Ishibashi, and R.M. Wald, A higher dimensional stationary rotating black hole must be axisymmetric, Commun. Math. Phys. 271 (2007), 699–722, arXiv:grqc/0605106.
 [24] J.M. Lee, The spectrum of an asymptotically hyperbolic Einstein manifold, Comm. Anal. Geom. 3 (1995), 253–271.
 [25] by same author, Fredholm operators and Einstein metrics on conformally compact manifolds, Mem. Amer. Math. Soc. 183 (2006), vi+83, arXiv:math.DG/0105046. MR MR2252687
 [26] R.R. Mazzeo and R.B. Melrose, Meromorphic extension of the resolvent on complete spaces with asymptotically constant negative curvature, Jour. Funct. Anal. 75 (1987), 260–310. MR 916753
 [27] N. Okuyama and K. Maeda, Fivedimensional black hole and particle solution with nonAbelian gauge field, Phys. Rev. D67 (2003), 104012, arXiv:grqc/0212022 [grqc].
 [28] Eugen Radu and D. H. Tchrakian, Gravitating YangMills fields in all dimensions, 418th WEHeraeusSeminar: Models of Gravity in Higher Dimensions: From theory to Experimental search Bremen, Germany, August 2529, 2008, 2009, arXiv:0907.1452 [grqc].
 [29] E. Winstanley, Existence of stable hairy black holes in SU(2) Einstein YangMills theory with a negative cosmological constant, Class.Quant.Grav. 16 (1999), 1963–1978.