LU-ITP-2006/004 PreprinttypesetinJHEPstyle-HYPERVERSION Ramifications of Lineland 6 0 0 2 n u Daniel Grumiller∗ and Rene Meyer∗† J 1 ∗Institute for Theoretical Physics, University of Leipzig 3 Augustusplatz 10-11, D-04109 Leipzig, Germany v 9 4 †Max Planck Institute for Mathematics in the Sciences 0 4 Inselstrasse 22, D-04103 Leipzig, Germany 0 6 0 E-mail: [email protected], [email protected]. / h t - p Abstract: A non-technical overview on gravity in two dimensions is provided. Ap- e h plications discussed in this work comprise 2D type 0A/0B string theory, Black Hole : v evaporation/thermodynamics, toy models for quantum gravity, for numerical Gen- i X eral Relativity in the context of critical collapse and for solid state analogues of r a Black Holes. Mathematical relations to integrable models, non-linear gauge theories, Poisson-sigma models, KdV surfaces and non-commutative geometry are presented. Keywords: Black Holes in String Theory, 2D Gravity, Integrable Models. Contents 1. Introduction 2 2. Gravity as non-linear gauge theory 3 2.1 First order formulation 3 2.2 Generic classical solutions 4 2.3 Constant dilaton vacua 6 2.4 Topological generalizations 7 2.5 Non-topological generalizations 7 3. Strings in 2D 9 3.1 Target space formulation of 2D type 0A/0B string theory 9 3.2 Exact string Black Hole 10 4. Black Holes 13 4.1 Classical analysis 13 4.2 Thermodynamics 13 4.3 Semi-classical analysis 15 4.4 Long time behavior 15 4.5 Killing horizons kill horizon degrees 16 4.6 Critical collapse 17 4.7 Quasinormal modes 19 4.8 Solid state analogues 20 5. Geometry from matter 20 5.1 Scalar matter 21 5.2 Fermionic matter 24 6. Mathematical issues 25 6.1 Remarks on the Einstein-Hilbert action in 2D 26 6.2 Relations to 3D: Chern-Simons and BTZ 26 6.3 Integrable systems, Poisson-sigma models and KdV surfaces 26 6.4 Torsion and non-metricity 27 6.5 Non-commutative gravity 28 – 1 – 1. Introduction The study of gravity in 2D — boring to some, fascinating to others [1] — has the undeniable disadvantage of eliminating a lot of structure that is present in higher dimensions; for instance, the Riemann tensor is determined already by the Ricci scalar, i.e., there is no Weyl curvature and no trace-free Ricci part. On the other hand, it has the undeniable advantage of eliminating a lot of structure that is present in higher dimensions; for instance, non-perturbative results may be obtained with relative ease due to technical simplifications, thus allowing one to understand some important conceptual issues arising in classical and quantum gravity which are uni- versal and hence of relevance also for higher dimensions. The scope of this non-technical overview is broad rather than focussed, since there exist already various excellent reviews and textbooks presenting the techni- cal pre-requisites in detail,1 and because the broadness envisaged here may lead to a cross-fertilization between otherwise only loosely connected communities. Some recent results are presented in more detail. It goes without saying that the topics selected concur with the authors’ preferences; by no means it should be concluded that an issue or a reference omitted here is devoid of interest. The common link between all applications mentioned here is 2D dilaton gravity,2 1 S = d2x√ g XR+U(X) ( X)2 2V(X) , (1.1) 2DG 2 − ∇ − Z (cid:2) (cid:3) the action of which depends functionally on the metric g and on the scalar field X. µν Notethatveryoften,inparticularinthecontextofstringtheory, thefieldredefinition X = e−2φ is employed; the field φ is the dilaton of string theory, hence the name “dilaton gravity”. However, it is emphasized that the natural interpretation of X need not be the one of a dilaton field — it may also play the role of surface area, dual field strength, coordinate of a suitable target space or black hole (BH) entropy, depending on the application. The curvature scalar R and covariant derivative ∇ are associated with the Levi-Civita connection and Minkowskian signature is implied unless stated otherwise. The potentials U, V define the model; several examples will be provided below. A summary is contained in table 1. This proceedings contribution is organized as follows: section 2 is devoted to a reformulation of (1.1) as a non-linear gauge theory, which considerably simplifies the construction of all classical solutions; section 3 discusses applications in 2D string theory; section 4 summarizes applications in BH physics; section 5 demonstrates how to reconstruct geometry from matter in a quantum approach; section 6 contains not only mathematical issues but also some open problems. 1For instance, the status of the field in the late 1980ies is summarized in [2]. 2The 2D Einstein-Hilbert action will not be discussed except in section 6.1. – 2 – Model (cf. (1.1) or (2.2)) U(X) V(X) w(X) (cf. (2.3)) 1. Schwarzschild [5] 1 λ2 2λ2√X −2X − − 2. Jackiw-Teitelboim [6,7] 0 ΛX 1ΛX2 − −2 3. Witten BH/CGHS [8,9] 1 2b2X 2b2X −X − − 4. CT Witten BH [8,9] 0 2b2 2b2X − − 5. SRG (D > 3) D−3 λ2X(D−4)/(D−2) λ2D−2X(D−3)/(D−2) −(D−2)X − − D−3 6. (A)dS ground state [10] a BX a = 2 : B X2−a 2 −X −2 6 −2(2−a) 7. Rindler ground state [11] a BXa BX −X −2 −2 8. BH attractor [12] 0 BX−1 B lnX −2 −2 9. All above: ab-family [13] a BXa+b b = 1 : B Xb+1 −X −2 6 − −2(b+1) 10. Liouville gravity [14] a beαX a = α : b e(a+α)X 6 − a+α 11. Scattering trivial [15] generic 0 const. 12. Reissner-Nordstr¨om [16] 1 λ2 + Q2 2λ2√X 2Q2/√X −2X − X − − 13. Schwarzschild-(A)dS [17] 1 λ2 ℓX 2λ2√X 2ℓX3/2 −2X − − − − 3 14. Katanaev-Volovich [18] α βX2 Λ X eαy(βy2 Λ)dy − − 15. Achucarro-Ortiz [19] 0 Q2 J ΛX Q2lnX + J 1ΛX2 X − 4X3 − R 8X2 − 2 16. KK reduced CS [20,21] 0 1X(c X2) 1(c X2)2 2 − −8 − 17. Symmetric kink [22] generic XΠn (X2 X2) cf. [22] − i=1 − i 18. 2D type 0A/0B [23,24] 1 2b2X + b2q2 2b2X + b2q2 lnX −X − 8π − 8π 19. exact string BH [25,26] (3.11) (3.11) (3.13) Table 1: Selected list of models 2. Gravity as non-linear gauge theory It has been known for a long time how to obtain all classical solutions of (1.1) not only locally, but globally. Two ingredients turned out to be extremely useful: a reformulation of (1.1) as a first order action and the imposition of a convenient (axial or Eddington-Finkelstein type) gauge, rather than using conformal gauge.3 Subsequently we will briefly recall these methods. For a more comprehensive review cf. [4]. 2.1 First order formulation The Jackiw-Teitelboim model (cf. the second model in table (1)) allows a gauge theoretic formulation based upon (A)dS , 2 [P ,P ] = Λε J , [P ,J] = ε bP , (2.1) a b ab a a b with Lorentz generator J, translation generators P and Λ = 0. A corresponding a 6 first order action, S = X FA, has been introduced in [27]. The field strength A 3In string theory almost eRxclusively conformal gauge is used. A notable exception is [3]. – 3 – F = dA + [A,A]/2 contains the SO(1,2) connection A = eaP + ωJ, and the La- a grange multipliers X transform under the coadjoint representation. This example A is exceptional insofar as it allows a formulation in terms of a linear (Yang-Mills type) gauge theory. Similarly, the fourth model in table 1 allows a gauge theoretic formulation [28] based upon the centrally extended Poincar`e algebra [29]. The gen- eralization to non-linear gauge theories [30] allowed a comprehensive treatment of all models (1.1) with U = 0, which has been further generalized to U = 0 in [31]. The 6 corresponding first order gravity action S = X Ta +XR+ǫ X+X−U(X)+V(X) (2.2) FOG a − Z (cid:2) (cid:0) (cid:1)(cid:3) isequivalent to(1.1)(withthesamepotentialsU,V)uponeliminationoftheauxiliary fields X andthetorsion-dependent partofthespin-connection. Hereisournotation: a ea = eadxµ is the dyad 1-form. Latin indices refer to an anholonomic frame, Greek µ indices to a holonomic one. The 1-form ω represents the spin-connection ωa = b εa ω = εa ω dxµ with the totally antisymmetric Levi-Civita symbol ε (ε = +1). b b µ ab 01 With the flat metric η in light-cone coordinates (η = 1 = η , η = 0 = η ) ab +− −+ ++ −− it reads ε± = 1. The torsion 2-form present in the first term of (2.2) is given by ± ± T± = (d ω) e±. The curvature 2-form Ra can be represented by the 2-form R b ± ∧ defined by Ra = εa R with R = dω. It appears in the second term in (2.2). Since no b b confusion between 0-forms and 2-forms should arise the Ricci scalar is also denoted by R. The volume 2-formis denoted by ǫ = e+ e−. Signs and factorsof the Hodge- ∧ ∗ operation are defined by ǫ = 1. It should be noted that (2.2) is a specific Poisson- ∗ sigma model [31] with a 3D target space, with target space coordinates X,X±, see section 6.3 below. A second order action similar to (1.1) has been introduced in [32]. 2.2 Generic classical solutions It is useful to introduce the following combinations of the potentials U and V: X X I(X) := exp U(y)dy, w(X) := I(y)V(y)dy (2.3) Z Z The integration constants may be absorbed, respectively, by rescalings and shifts of the “mass”, see equation (2.9) below. Under dilaton dependent conformal transfor- mations Xa Xa/Ω, ea eaΩ, ω ω+X eadlnΩ/dX theaction(2.2)ismapped a → → → to a new one of the same type with transformed potentials U˜, V˜. Hence, it is not invariant. It turns out that only the combination w(X) as defined in (2.3) remains invariant, so conformally invariant quantities may depend on w only. Note that I is positive apart from eventual boundaries (typically, I may vanish in the asymptotic ˜ region and/or at singularities). One may transform to a conformal frame with I = 1, solve all equations of motion and then perform the inverse transformation. Thus, it – 4 – is sufficient to solve the classical equations of motion for U˜ = 0, dX +X˜−e˜+ X˜+e˜− = 0 , (2.4) − (d ω˜)X˜± e˜±V˜(X) = 0 , (2.5) ± ∓ (d ω˜) e˜± = 0 , (2.6) ± ∧ which is what we are going to do now. Note that the equation containing dω˜ is redundant, whence it is not displayed. Let us start with an assumption: X˜+ = 0 for a given patch. To get some 6 physical intuition as to what this condition could mean: the quantities Xa, which are the Lagrangemultipliers fortorsion, can be expressed asdirectional derivatives of the dilaton field by virtue of (2.4) (e.g. in the second order formulation a term of the form XaX corresponds to ( X)2). For those who are familiar with the Newman- a ∇ Penrose formalism: for spherically reduced gravity the quantities Xa correspond to the expansion spin coefficients ρ and ρ′ (both are real). If X˜+ vanishes a (Killing) horizon is encountered and one can repeat the calculation below with indices + and swapped everywhere. If bothvanish inan open regionby virtue of (2.4) a constant − dilaton vacuum emerges, which will be addressed separately below. If both vanish on isolated points the Killing horizon bifurcates there and a more elaborate discussion is needed [33]. The patch implied by X˜+ = 0 is a “basic Eddington-Finkelstein patch”, 6 i.e., a patch with a conformal diagram which, roughly speaking, extends over half of the bifurcate Killing horizon and exhibits a coordinate singularity on the other half. In such a patchone mayredefine e˜+ = X˜+Z with anew 1-formZ. Then (2.4) implies e˜− = dX/X˜+ + X˜−Z and the volume form reads ǫ˜ = e˜+ e˜− = Z dX. The + ∧ ∧ component of (2.5) yields for the connection ω˜ = dX˜+/X˜+ +ZV˜(X). One of the − torsion conditions (2.6) then leads to dZ = 0, i.e., Z is closed. Locally (in fact, in the whole patch) it is also exact: Z = du. It is emphasized that, besides the integration of (2.8) below, this is the only integration needed! After these elementary steps one obtains already the conformally transformed line element in Eddington-Finkelstein (EF) gauge ds˜2 = 2e˜+e˜− = 2du dX +2X˜+X˜−du2, (2.7) which nicely demonstrates the power of the first order formalism. In the final step the combination X˜+X˜− has to be expressed as a function of X. This is possible by noting that the linear combination X˜+ [(2.5) with index] + X˜− [(2.5) with + × − × index] together with (2.4) establishes a conservation equation, d(X˜+X˜−)+V˜(X)dX = d(X˜+X˜− +w(X)) = 0. (2.8) Thus, there isalways a conserved quantity (dM = 0), which in theoriginalconformal frame reads M = X+X−I(X) w(X), (2.9) − − – 5 – where the definitions (2.3) have been inserted. It should be noted that the two free integration constants inherent to the definitions (2.3) may be absorbed by rescalings and shifts of M, respectively. The classical solutions are labelled by M, which may be interpreted as mass (see section 4.2). Finally, one has to transform back to the original conformal frame (with conformal factor Ω = I(X)). The line element (2.7) by virtue of (2.9) may be written as ds2 = 2I(X)du dX 2I(X)(w(X)+M)du2. (2.10) − Evidently there is always a Killing vector K ∂ = ∂/∂u with associated Killing norm · K2 = 2I(w +M). Since I = 0 Killing horizons are encountered at X = X where h − 6 X is a solution of h w(X )+M = 0. (2.11) h It is recalled that (2.10) is valid in a basic EF patch, e.g., an outgoing one. By redoing the derivation above, but starting from the assumption X− = 0 one may 6 obtain an ingoing EF patch, and by gluing together these patches appropriately one may construct the Carter-Penrose diagram, cf. [4,33,34]. As pointed out in the introduction the full geometric information resides in the Ricci scalar. The one related to the generic solution (2.10) reads 2 d R = U(X)(M +w(X))+I(X)V(X) . (2.12) I(X)dX (cid:16) (cid:17) There are two important special cases: for U = 0 the Ricci scalar simplifies to R = 2V′(X), while for w(X) 1/I(X) it scales proportional to the mass, R = ∝ 2MU′(X)/I(X). Thelattercasecomprisesso-calledMinkowskian groundstatemod- els (for examples cf. the first, third, fifth and last line in table 1). Note that for many models in table 1 the potential U(X) has a singularity at X = 0 and consequently a curvature singularity arises. 2.3 Constant dilaton vacua For sake of completeness it should be mentioned that in addition to the family of generic solutions (2.10), labelled by the mass M, isolated solutions may exist, so- called constant dilaton vacua (cf. e.g. [22]), which have to obey4 X = X = const. CDV with V(X ) = 0. The corresponding geometry has constant curvature, i.e., only CDV Minkowski, Rindler or (A)dS are possible space-times for constant dilaton vacua.5 2 The Ricci scalar is determined by R = 2V′(X ) = const. (2.13) CDV CDV 4Incidentally, for the generic case (2.10) the value of the dilaton on an extremal Killing horizon is also subject to these two constraints. 5In quintessence cosmology in 4D such solutions serve as late time dS4 attractor [35]. In 2D dilaton supergravity solutions preserving both supersymmetries are necessarily constant dilaton vacua [36]. – 6 – Examples are provided by the last eighth entries in table 1. For instance, 2D type 0A strings with an equal number q of electric and magnetic D0 branes (cf. the penultimate entry in table 1) allow for an AdS vacuum with X = q2/(16π) and 2 CDV R = 4b2 [37]. CDV − 2.4 Topological generalizations In 2D there are neither gravitons nor photons, i.e. no propagating physical modes exist [38]. This feature makes the inclusion of Yang-Mills fields in 2D dilaton gravity or an extension to supergravity straightforward. Indeed, both generalizations can be treated again in the first order formulation as a Poisson-sigma model, cf. e.g. [39]. In addition to M (see (2.9)) more locally conserved quantities (Casimir functions) may emerge and the integrability concept is extended. As a simple example we include an abelian Maxwell field, i.e., instead of (2.2) we take S = X Ta +XR+BF +ǫ X+X−U(X,B)+V(X,B) , (2.14) MDG a − Z (cid:2) (cid:0) (cid:1)(cid:3) where B is an additional scalar field and F = dA is the field strength 2-form. Varia- tion with respect to A immediately establishes a constant of motion, B = Q, where Q is some real constant, the U(1) charge. Variation with respect to B may produce a relation that allows to express B as a function of the dilaton and the dual field strength F. For example, suppose that V(X,B) = V(X) + 1B2. Then, variation ∗ 2 with respect to B gives B = F. Inserting this back into the action yields a stan- −∗ dard Maxwell term. The solution of the remaining equations of motion reduces to the case without Maxwell field. One just has to replace B by its on-shell value Q in the potentials U, V. Concerning supergravity we just mention a couple of references for further ori- entation [36,40,41]. 2.5 Non-topological generalizations To get a non-topological theory one can add scalar or fermionic matter. The action for a real, self-interacting and non-minimally coupled scalar field , T 1 S = F(X)d d +ǫf(X, ) , (2.15) T 2 T ∧∗ T T Z h i in our convention requires F < 0 for the kinetic term to have the correct sign; e.g. F = κ or F = κX. − − While scalar matter couples to the metric and the dilaton, fermions6 couple directly to the Zweibein (A←→dB = AdB (dA)B), − i S = F(X) ( ea) (χγ ←→d χ)+ǫH(X)g(χχ) , (2.16) χ a 2 ∗ ∧ Z h i 6We use the same definition for the Dirac matrices as in [42]. – 7 – but not — and this is a peculiar feature of 2D — to the spin connection. The self-interaction is at most quartic (a constant term may be absorbed in V(X)), g(χχ) = mχχ+λ(χχ)2. (2.17) The quartic term (henceforth: Thirring term [43]) can also be recast into a classically equivalent form by introducing an auxiliary vector potential, λ λ ǫ(χχ)2 = [A A+2A ( e )χγaχ] , (2.18) a 2 ∧∗ ∧ ∗ Z Z which lacks a kinetic term and thus does not propagate by itself. We speak of minimal coupling if the coupling functions F(X),f(X, ),H(X) do T not depend on the dilaton X, and of nonminimal coupling otherwise. As an illustration we present the spherically reduced Einstein-massless-Klein- Gordon model (EMKG). It emerges from dimensional reduction of 4D Einstein- Hilbert (EH) gravity (cf. the first model in table 1) with a minimally coupled scalar field, with the choices f(X,τ) = 0 and 1 w(X) = 2λ2√X , F(X) = κX , I(X) = , (2.19) − − √X where λ is an irrelevant scale parameter and κ encodes the (also irrelevant) Newton coupling. Minimally coupled Dirac fermions in four dimensions yield upon dimen- sionalreductiontwo2-spinorscoupledtoeachotherthroughintertwinor terms, which isnotcoveredby (2.16)(see[44]fordetailsonsphericalreductionoffieldsofarbitrary spin and the spherical reduced standard model). With matter the equation of motion (2.5) and the conservation law (2.8) obtain contributions W± = δ(S +S )/δe∓ and X−W+ +X+W−, respectively, destroying T χ integrability because Z is not closed anymore: dZ = W+ Z/X+. In special cases ∧ exact solutions can be obtained: 1. For (anti-)chiral fermions and (anti-)selfdual scalars with W+ = 0 (W− = 0) the geometric solution (2.7) is still valid [4] and the second equation of motion (2.5) implies W− = W−du. Such solutions have been studied e.g. in [45,46]. u They arise also in the Aichelburg-Sexl limit [47] of boosted BHs [48]. 2. A one parameter family of static solutions of the EMKG has been discovered in [49]. Studies of static solutions in generic dilaton gravity may be found in [50,51]. A static solution for the line-element with time-dependent scalar field (linear in time) has been discussed for the first time in [52]. It has been studied recently in more detail in [53]. 3. A (continuously) self-similar solution of the EMKG has been discoverd in [54]. – 8 – 4. Specific models allow for exact solutions even in the presence of more general matter sources; for instance, the conformally transformed CGHS model (fourth in table 1), Rindler ground state models (seventh in table 1) and scattering trivial models (eleventh in table 1). 3. Strings in 2D Stringspropagatingina2Dtargetspacearecomparativelysimpletodescribebecause theonlypropagatingdegree offreedomisthetachyon (andifthelatter isswitched off the theory becomes topological). Hence several powerful methods exist to describe the theory efficiently, e.g. as matrix models. In particular, strings in non-trivial backgrounds may be studied in great detail. Here are some references for further orientation: For the matrix model description of 2Dtype 0A/0Bstring theory cf.[23, 55] (for an extensive review on Liouville theory and its relation to matrix models and strings in 2D cf. [14]; some earlier reviews are refs. [56]; the matrix model for the 2D Euclidean string BH has been constructed in [57]; a study of Liouville theory from the 2D dilaton gravity point of view may be found in [58]). The low energy effective action for 2D type 0A/0Bstring theory in the presence of RR fluxes has been studied from various aspects e.g. in [23,24,37,59]. 3.1 Target space formulation of 2D type 0A/0B string theory Forsakeofdefiniteness focuswillbeon2Dtype0Awithanequalnumber q ofelectric and magnetic D0 branes, but other cases may be studied as well. For vanishing tachyon the corresponding target space action is given by (setting κ2 = 1) 1 b2q2 S = d2x√ g e−2φ R 4 ( φ)2 +4b2 , (3.1) 0A 2 − − ∇ − 4π Z (cid:20) (cid:21) (cid:0) (cid:1) Obviously, this is a special case of the generic model (1.1), with U,V given by the penultimate model in table 1, to which all subsequent considerations — in particular thermodynamical issues — apply. Note that the dilaton fields X and φ are related by X = exp( 2φ). The constant b2 = 2/α′ defines the physical scale. In the absence − of D0 branes, q = 0, the model simplifies to the Witten BH, cf. the third line in table 1. The action defining the tachyon sector up to second order in is given by T (cf. (2.15)) 1 S = d2x√ g[F(X)gµν(∂ )(∂ )+f(X, )] , (3.2) T µ ν 2 − T T T Z with q2 F(X) = X , f( ,X) = b2 2 X . (3.3) T T − 2π (cid:18) (cid:19) The total action is S +S . 0A T – 9 –