Hidden symmetries and decay for the wave equation on the Kerr spacetime L. Andersson and P. Blue REPORT No. 44, 2008/2009, fall ISSN 1103-467X ISRN IML-R- -44-08/09- -SE+fall HIDDEN SYMMETRIES AND DECAY FOR THE WAVE EQUATION ON THE KERR SPACETIME LARSANDERSSON† ANDPIETERBLUE‡ Abstract. Energyanddecayestimatesforthewaveequationontheexterior region of slowly rotating Kerr spacetimes are proved. The method used is a generalizationofthevector-fieldmethod,whichallowstheuseofhigher-order symmetryoperators. Inparticular,ourmethodmakesuseofthesecond-order Carter operator, which is a hidden symmetry in the sense that it does not correspondtoaKillingsymmetryofthespacetime. Themainresultgives,instationaryregions,analmostinverselineardecay rate and the corresponding decay rate at the event horizon and null infinity. Except for the small loss in the decay rate, this generalizes the known decay resultsontheexteriorregionoftheSchwarzschildspacetime. Contents 1. Introduction 1 2. Notation and preliminaries 12 3. The bounded-energy argument 17 4. Decay estimates for the local energy 36 5. Pointwise decay estimates 54 Appendix A. Relevance and global structure of Kerr 64 References 67 1. Introduction Inthispaperweproveboundednessanddecayforsolutionsofthecovariantwave equation α ψ =0 α ∇ ∇ in the exterior region of the Kerr spacetime. In Boyer-Lindquist coordinates, the exterior region is given by (t,r,θ,φ) R (r+, ) S2 with the Lorentz metric ∈ × ∞ × 2Mr 4Mrasin2θ g dxµdxν = 1 dt2 dtdφ µν − − Σ − Σ (cid:18) (cid:19) Πsin2θ Σ + dφ2+ dr2+Σdθ2, (1.1) Σ ∆ where r =M +√M2 a2 and + − ∆=r2 2Mr+a2, Σ=r2+a2cos2θ, Π=(r2+a2)2 a2sin2θ∆. − − For0 a M, theKerrmetricdescribesarotatingblackhole, withmassM and ≤| |≤ angular momentum Ma, and with horizon located at r = r . The Schwarzschild + spacetime is the subcase with a = 0. The exterior region is globally hyperbolic, with the surfaces of constant t, Σ , as Cauchy surfaces. Thus, the wave equation is t well posed in the exterior region, even though the Kerr spacetime can be extended. We consider initial data on the hypersurface Σ . 0 Date:17August2009. 1 2 L.ANDERSSONANDP.BLUE The Kerr black-hole spacetime is expected to be the unique, stationary, asymp- totically flat spacetime containing a nondegenerate Killing horizon [2]. Further, motivated by considerations including the weak cosmic censorship conjecture, the Kerrblackholeisexpectedtobetheasymptoticlimitoftheevolutionofasymptoti- callyflat,vacuumdataingeneralrelativity. Animportantsteptowardsestablishing thevalidityofthisscenarioistoprovethestabilityoftheKerrsolution,i.e. toshow that vacuum spacetimes evolving from data which represent a small perturbation of Kerr initial data asymptotically approach a Kerr solution. Decay for the scalar waveequationontheKerrbackgroundisanimportantmodelproblemforstability, and due to the importance of this application, the wave equation on black hole backgrounds has been actively studied in the last decade. An essential tool in the analysis of both linear and nonlinear Lagrangian field equations is the use of Noetherian currents associated to Killing or conformal sym- metries of the background spacetime. In the relativistic setting, we interpret these currents as momenta. A method, which may be referred to as the vector-field method, based on the systematic use of such currents, has been developed and has played an essential role in the proof of the nonlinear stability of Minkowski space [12], which built on earlier vector-field based estimates for the decay rates of solutions to linear and nonlinear wave equations [28] and to Maxwell’s equa- tions and the spin-2 field equations [11]. Generalizations of these ideas have played a central role in recent work concerning the wave and Maxwell equations on the Schwarzschild spacetime, and the wave equation on the Kerr spacetime. For these non-flat background spacetimes, the lack of symmetries presents an important new problem. The 10 dimensional group of isometries of the Minkowski space is broken to a 4 dimensionalgroupfortheSchwarzschildspacetime, generatedby∂ andthespatial t rotations. Further, the Schwarzschild spacetime contains orbiting null geodesics, located at the photon spere, the hypersurface with r = 3M. As high frequency waves can track null geodesics for long times, an analysis of this feature is an essential step in a proof of decay for the wave equation. InthegeneralKerrspacetime,witha=0,whichweconsiderinthispaper,there 6 areonlytwoKillingfields,∂ and∂ . Inaddition,instudyingthewaveequationon t φ the Kerr spacetime, one encounters several new phenomena which are not present in the Schwarzschild case. There is an ergo-region outside the horizon, where the stationary Killing field ∂ fails to be timelike. Thus, the Kerr spacetime admits no t positivedefinite, conservedenergyforthewaveequation. Further, theorbitingnull geodesics in Kerr fill an open region in spacetime. The lack of symmetries of the Kerr spacetime is compensated for by the presence of a fundamentally new feature, which we make essential use of in this paper, a hidden symmetry. By a hidden symmetry we mean an operator which commutes with the wave op- erator, not associated to a Killing vector field, but rather to a second-rank Killing tensor. For the Kerr spacetime, the Killing tensor and the related conserved quan- tity found by Carter [8] provides, via the associated second-order Carter operator, a hidden symmetry. The existence of the two Killing vectors and the Killing ten- sor imply the separability of many important equations on the Kerr spacetime, including the wave equation. One of the fundamentally new ideas introduced in this work is a generalization of the vector-field method which allows the use of not only Killing symmetries but also hidden symmetries in the construction of suitable Noetherian currents for the analysis of Lagrangian field equations. This allows us, in contrast to other, recent work on the wave equation on Kerr, to carry out our proof of uniform boundedness and decay results exclusively in physical space, using only the coordinate functions HIDDEN SYMMETRIES AND DECAY FOR THE WAVE EQUATION ON KERR 3 and differential operators. This technique almost eliminates the need for methods involving separation of variables or Fourier analysis.1 To state our main results, we use the tortoise coordinate r , defined by ∗ dr =(r2 2Mr+a2)(r2+a2), r (3M)=0, dr − ∗ ∗ and the almost null coordinates u given by ± u =t+r , u =t r . + ∗ − − ∗ Our main results are: Theorem 1.1 (Uniformly bounded, positive energy). There are positive constants C and a¯ , and a nonnegative quadratic form on each hypersurface of constant t, 1 1 ETχ[ψ](t), such that, if |a|<a¯1 and ψ :R×(r+,∞)×S2 →R is a solution of the wave equation, α ψ =0, then t α ∇ ∇ ∀ E [ψ](t) C E [ψ](0). Tχ ≤ 1 Tχ Theorem 1.2 (Decayestimates). There are positive constants C , C , and a¯ , and 2 20 2 there is a nonnegative quadratic form on each hypersurface of constant t, ψ 2(t), k k suchthat,if a <a¯2 andψ :R (r+, ) S2 Risasolutionofthewaveequation, | | × ∞ × → α ψ =0, then there are the following decay estimates t>0,(θ,φ) S2: α ∇ ∇ ∀ ∈ (1) Decay in stationary regions: r (3M,4M): ∀ ∈ ψ(t,r,θ,φ) C2max 1,t −1+C20|a| ψ (0). | |≤ { } k k (2) Near decay: r <3M: ∀ ψ(t,r,θ,φ) C2max 1,u+ −1+C20|a| ψ (0). | |≤ { } k k (3) Far decay: r with r >4M and r <t: ∀ 1/2 u u ψ(t,r,θ,φ) C2r−1max 1,u C20|a| +− − ψ (0). | |≤ { −} (cid:18)u+max{1,u−}(cid:19) k k In particular, for t/2<r <t: ψ(t,r,θ,φ) C2r−1max 1,u −1/2+C20|a| ψ (0). | |≤ { −} k k (4) Decay near spatial infinity: r with r >t: ∀ ψ(t,r,θ,φ) C r 1max 1, u 1/2 ψ (0). 2 − − | |≤ { − −} k k Theorem1.1istheconclusionofsection3andisgivenintheorem3.13. Theorem 1.2 follows from the conclusions of theorems 5.1, 5.2, and 5.4. Except for the loss in the exponent of C a, the estimates stated in theorem 1.2 are the same as the 20| | results proven using vector-field techniques in the Schwarzschild spacetime.2 The decay along null infinity is the same (modulo the loss in the exponent) as can be obtained in Minkowski space from initial data on t = 0 which decays like r 3/2. − This is roughly the decay rate we require for the initial data in Kerr. Note that the Kerr spacetime has a discrete, time reversal symmetry (t,φ) ( t, φ), so that 7→ − − the results of theorems 1.1 and 1.2 can also be reversed to t<0. The norm ψ (0) in theorem 1.2 is bounded if ψ is smooth on the hypersurface k k Σ and if ψ and its first nine derivatives (with respect to the Boyer-Lindquist 0 1WedonotuseseparabilityortheFouriertransform(whichareessentiallyequivalenttoeach other) in the t coordinate. In the proof of lemma 3.12, we need to use separability in the φ coordinatetocontroltheaxiallysymmetriccomponentofthesolution,butareotherwiseableto avoid using separability (or the Fourier transform) in φ. We never use separability in the r and θ coordinates, since this is only possible once separation in the t and φ coordinates has already beenperformed. 2Recently, better decayestimates havebeen provenintheSchwarzschild case, giving, for any positiveδ,decayratesoft−3/2+δ andu−+3/2+δ instationaryandnearregionsrespectively[32]. 4 L.ANDERSSONANDP.BLUE coordinates) decay like r 3/2+δ for some positive δ as r . By “smooth” we − → ∞ meanC withrespecttolocalcoordinates. Asr r ,thisisnotthesameasbeing ∞ + → smoothwithrespecttotheBoyer-Lindquistcoordinates,sincetheydegeneratehere. The quadratic form E [ψ](0) in theorem 1.1 is bounded when ψ 2(0) in theorem Tχ k k 1.2 is bounded. We briefly comment on other related work. Estimates for the decay rate of so- lutions to the wave equation have been proven in the subcase of the Schwarzschild spacetime, where a = 0. Birkhoff’s theorem states that the Schwarzschild space- time is the unique spherically symmetric, vacuum spacetime solution of Einstein’s equation. For the coupled Einstein and scalar wave system, a decay rate and non- linear stability of the Schwarzschild solution have been proven in the spherically symmetric setting [13]. As mentioned earlier, for the wave equation without a symmetry assumption but on a fixed background spacetime, the case of the linear wave equation on the Schwarzschild spacetime is significantly simpler than the corresponding case in the Kerr spacetime, since the ∂ Killing vector is timelike in the entire exterior region t andgeneratesaconserved,positiveenergy,thereisthefullSO(3)groupofrotation symmetries available to generate higher energies, and the orbiting null geodesics are restricted to r = 3M. The first two of these properties imply that solutions remain bounded. Following the introduction of a Morawetz vector field and of the equivalent of an almost conformal vector field to the Schwarzschild spacetime [30], decay estimates for the wave equation were proven [7], proven with a weaker decay ratebutlessregularityloss[6],andprovenseparatelywithastrongerestimatenear the event horizon [14]. These were extended to Strichartz estimates for the wave equation [33] and decay estimates for Maxwell’s equation [4]. The Morawetz vector field which made these estimates possible was centred about the orbiting geodesics atr =3M. Thisconstructionofaclassicalvectorfieldfailswhena=0,sincethere 6 are orbiting geodesics filling an open set in spacetime. Recently,newMorawetzvectorfieldswithcoefficientsthatdependonbothspace- time position and on Fourier operators have been introduced to construct a uni- formly bounded, positive energy and to show that solutions to the Kerr wave equa- tion remain uniformly bounded [17, 39]. These might reasonably be called Fourier- analytic,pseudodifferential,microlocal,orphase-spacetechniques,sincetheFourier operators represent coordinates in momentum space in contrast to spacetime co- ordinates in physical or configuration space. These results include a form of weak decay, sincetheMorawetzestimateimpliesintegrabilityofthelocalenergy. In[16], it has been announced that these Fourier-analytic techniques can be extended to decay results very similar to ours in theorem 1.2. Like our work, these require that a is very small relative to M. We compare this work with our own in more detail | | in section 1.3. Fourier-analytic vector fields were used previously to prove Mourre estimates, which are similar to Morawetz estimates, in the proof of scattering for the Klein-Gordon equation [25] and the Dirac equation [26]. Finally, we recall that decay for the wave equation has previously been obtained [21]fromanexplicitrepresentationofsolutions[20]usingthecompleteseparability oftheKerrwaveequation. Thesedecayresultsareoftheformlim ψ(t,r,θ,φ) t 0, where ψ(t,r,θ,φ) = ψLz(t,r,θ)eiLzφ or where ψ is made up o→f∞a|finite numb|e→r of azimuthal modes of this form. Decay rates have been obtained from this sepa- rability method for solutions to the Dirac equation [19] and spherically symmetric solutions to the wave equation when a=0 [29]. The decay without rate results for thewaveequationbuiltontheearlierresultthattherearenoexponentiallygrowing modes [43] and applies for all a [0,M]. | |∈ 1.1. Hiddensymmetriesandthevector-fieldmethod. Intheclassicalvector- field method, one seeks to control the solution of field equations by making use of HIDDEN SYMMETRIES AND DECAY FOR THE WAVE EQUATION ON KERR 5 energyfluxesanddeformationtermscalculatedforsuitablychosenvectorfields. The vector fields used are often (approximate) conformal symmetries of the background spacetime. Higher-orderestimatesareachievedbyLiedifferentiatingthefieldalong furthervectorfields. Animportantadvantageofthevector-fieldmethodisthatone works entirely in terms of quantities in physical space. Aparticularlyclearexampleofthevector-fieldmethodwastheuseoftheLorentz group in the Minkowski spacetime, R1+3, to construct norms which could be used intheKlainerman-Sobolevinequalityandtousethistoprovethewell-posednessof nonlinear wave equations [28]. The term “vector-field method” seems to have been introduced relatively recently, especially to describe generalisations of the work in R1+3 to situations where one lacks the full Lorentz group of symmetries. Applying this terminology retroactively, we would now describe the early uses of the radial Morawetz vector-field [35] and the conformal vector-field [24] as applications of the vector-field method. This may have previously been refered to as the method of multipliersortheEuler-Lagrangemethod,althoughthesetermscanalsobeapplied to more general techniques. A central result in mathematical relativity, and perhaps the most important ap- plication of the vector-field method, was the proof of the nonlinear stability of the Minkowski spacetime [12]. There had also been earlier work on the stability of the Minkowski spacetime, but this required hyperboloidal initial data [22]. The mon- umental proof of nonlinear stability built upon previous vector-field estimates for linearandnonlinearwaveequations[28],andfortheMaxwellandspin-2fieldequa- tions [11], which are better models for Einstein’s equations. This partly motivates our work on the linear wave equation in the Kerr spacetime using generalisations of the vector-field method. Since the original proof of nonlinear stability for the Minkowski spacetime, a simpler proof has been developed, but this also makes use of the vector-field method [31]. As mentioned above, in the Kerr spacetime, the lack of symmetries, as well as the complicated nature of the orbiting null geodesics, makes it impossible to derive the required estimates using only classical vector fields. In this section, we outline ageneralizationofthevector-fieldmethodwhichallowsustotakeadvantageofthe presence of hidden symmetries in the Kerr spacetime. In particular, we consider energiesbasedonoperatorsofordergreaterthanone,ratherthanjustvectorfields. Let (cid:3)g = α α. In the discussion here, we consider the scalar wave equation ∇ ∇ (cid:3)gψ = 0, but much of the discussion applies equally to general field equations derivedfromaquadraticaction. Wedefineasymmetryoperatortobeadifferential operator S such that if (cid:3)gψ = 0, then also (cid:3)gSψ = 0. The set of symmetry operatorsisclosedunderscalarmultiplication,addition,andcomposition,andeach symmetryoperatorhasawell-definedorderasadifferentialoperator. Thus,theset of symmetry operators forms a graded algebra. Given a set of generators of the set of symmetries, we can consider the subset consisting of generators of order n. We denote this subset of the generators of the symmetry operators by Sn and denote the elements of Sn with an underlined index, e.g. Sa Sn. ∈ IfXisaconformalKillingfield,thentheoperator generatedbyLiedifferenti- X L ationwithrespecttoXisclearlyasymmetryoperator. Wetakeahiddensymmetry tobeasymmetryoperatorwhichisnotinthealgebrageneratedbytheKillingvec- torfields. SincetheMinkowskispacetimesaturatestheDelong-Takeuchi-Thompson inequality, there are no hidden symmetries [10]. In the Schwarzschild spacetime, there are no hidden symmetries [9]. In the Kerr spacetime, it is well-known that thereisCarter’sKilling2tensorandthatthisgeneratesahiddensymmetry[8,42]. The energy-momentum tensor for the wave equation is 1 T[ψ] = ψ ψ g ( ψ γψ). (1.2) αβ α β αβ γ ∇ ∇ − 2 ∇ ∇ 6 L.ANDERSSONANDP.BLUE The momentum associated with a vector field X and the energy associated with a vector field X and evaluated on a hypersurface Σ are P [ψ] =T[ψ] Xβ, X α αβ E [ψ](Σ)= P [ψ] dηα, X X α ZΣ where dη is integration with respect to the surface volume induced by g on Σ. In the following, unless there is room for confusion, we will drop reference to ψ in the notation for momentum and energy. When the spacetime is foliated by surfaces of constant time, we will denote these surfaces by Σ and typically denote the energy t on such a surface by E (t)=E (Σ ). X X t The energy momentum tensor (1.2) satisfies the dominant energy condition, and hence for X timelike, the energy induced on a hypersurface with a timelike normal (i.e. a spacelike hypersurface) is positive definite. The energy conservation law takes the form E (Σ ) E (Σ )= ( Pα) g d4x, X 2 − X 1 ∇α X −| | ZΩ p where Ω is the region enclosed between Σ and Σ . This is often referred to as the 1 2 deformationformula. Energyestimatesareoftenperformedbycontrollingthebulk (also called deformation) terms Pα. However, for the Morawetz estimate (e.g. ∇α X inequality (3.6)), one makes use of the sign of the bulk term itself. Byestimatinghigher-orderenergiesonemay,viaSobolevestimates,getpointwise control of the fields. Higher-order energies may be defined by using symmetries. If for 0 i n, there is a collection of order-n differential operators, Si, then we can ≤ ≤ define the higher-order energy (of order n) for a vector field X to be n E [ψ](Σ)= E [Sψ](Σ). X,n X Xi=0SX∈Si Since the energy momentum tensor is quadratic in ψ, we can define a bilinear form of the energy momentum by 1 T[ψ ,ψ ] = (T[ψ +ψ ] T[ψ ψ ] ). 1 1 αβ 1 2 αβ 1 2 αβ 4 − − It is convenient to define an index version of the bilinear energy momentum, with respect to a set of symmetry operators S by a { } T[ψ] =T[S ψ,S ψ] . abαβ a b αβ Given a double-indexed collection of vector fields, Xab , we define the associated { } generalized momentum and energy to be P [ψ] =T[ψ] Xabβ, Xab α abαβ E [ψ](Σ)= P [ψ] dηα. Xab Xab α ZΣ Inpracticeitisconvenienttoconsidermomentawithlower-orderterms,designed to improve certain deformation terms in Pα. For a scalar function, q ([33], but ∇α X previously appearing in [15]), or a double-indexed collection of functions, qab, the associated momenta are defined to be 1 P [ψ] =q( ψ)ψ (∂ q)ψ2, q α α α ∇ − 2 1 P [ψ] =qab( S ψ)S ψ (∂ qab)(S ψ)(S ψ). qab α ∇α a b − 2 α a b For a pair consisting of a vector field and a scalar function, (X,q), the associated momentumisdefinedtobethesumofthemomentaassociatedwiththevectorfield and the scalar. For a pair of collections, (Xab,qab), again the momentum is defined HIDDEN SYMMETRIES AND DECAY FOR THE WAVE EQUATION ON KERR 7 tobethesumofthemomenta. Inallcases,theenergyonahypersurfaceisgivenby the flux, defined with respect to the momentum vector, through the hypersurface. Itisimportanttopointout,asweshowinlemma2.1,thatthedeformationterms forthegeneralizedmomentaarecomputationallynotmuchmoredifficulttohandle than the classical ones. As for the classical momenta and energies, in defining the generalizedvectorfields,momenta,andenergiesasoutlinedabove,oneisinterested in getting positive definiteness of the energies or bulk terms. Here, an additional subtlety arises. Namely, in the Morawetz estimate presented in equation (3.6), one achieves positive definiteness only modulo boundary terms. We generate these boundarytermswhenweintegratebypartstousetheformalself-adjointnessofthe second-order symmetry operators. These boundary terms can then be controlled by the energy. The presence of these boundary terms is a completely new feature compared to the classical energies and momenta. 1.2. Symmetries and null geodesics of Kerr. For any geodesic, the quantity g γ˙αγ˙β is a constant of the motion. Given a Killing field ξ, the quantity p = αβ ξ g γ˙αξβ is an additional conserved quantity. In Kerr, we have the Killing fields ∂ αβ t and ∂ with the associated constants of motion p and p . For a timelike or null φ t φ geodesic, these correspond to the energy and the angular momentum of a particle or photon with world line γ and are denoted E and L . z More generally, if the spacetime admits a Killing k-tensor, i.e. a symmetric tensor K which solves the Killing equation K = 0, then K = Kα1···αkγ˙αα11····α··kγ˙αk is a conserved quantity. In th∇e(αp1artαi2c·u··lαakr+1c)ase of Killing 2- tensors, which is the only case we are interested in here, there is associated to the Killing tensor a symmetry operator K = αKαβ β, such that [K,(cid:3)g]=0 [8, 42]. ∇ ∇ Since the commutator is zero, this operator is clearly a symmetry in the slightly weaker sense defined in the previous section. In Kerr, Carter’s Killing 2-tensor, provides a fourth constant of the motion Q= Q γ˙αγ˙β. For a null geodesic, we have αβ cos2θ Q=p2+ p2 +a2sin2θp2. θ sin2θ φ t A similar expression exists for timelike or spacelike geodesics. Any linear combi- nation of E2, EL , and L2 can be added to Q to give an alternate choice for the z z fourth constant of the motion. The form we have chosen is uncommon, but useful for our purposes because it is nonnegative. As was demonstrated by Carter, the presence of the extra conserved quantity allows one to separate the equations of geodesic motion. Of most interest to us is the equation for the r coordinate of null geodesics, 2 dr Σ2 = (r;M,a;E,L ,Q), (1.3) z dλ −R (cid:18) (cid:19) where (r;M,a;E,L ,Q)= (r2+a2)2E2 4aMrEL +(∆ a2)L2+∆Q. (1.4) R z − − z − z Onefindsthatorbitingnullgeodesics,i.e. oneswhichdonotfallintotheblackhole orescapetoinfinity,musthaveorbitswithconstantr. Thervaluesallowingorbiting null geodesics are solutions to the equations = 0, ∂ /∂r = 0. The solutions to R R this system in the exterior region turn out to be unstable, which corresponds with our conventions to ∂2 /∂r2 <0. R In the Schwarzschild case, i.e. for a = 0, there are only orbits on the sphere at r = 3M, which is called the photon sphere. For nonzero a, the orbiting null geodesics fill up an open region in spacetime which we shall also refer to as the photon sphere in the Kerr case. As a 0, the photon sphere tends to r = 3M. → 8 L.ANDERSSONANDP.BLUE There are many descriptions of the Kerr spacetime and its geodesics, including [3, 23, 40]. In Boyer-Lindquist coordinates, the d’Alembertian (cid:3)g = α α takes the form ∇ ∇ 1 1 (cid:3)g = ∂r∆∂r+ (r;M,a;∂t,∂φ,Q) , (1.5) Σ ∆R (cid:18) (cid:19) where is given by (1.4) with the conserved quantities E,L ,Q replaced by their z R corresponding operators ∂ ,∂ , and the second-order Carter operator3 Q, t φ 1 cos2θ Q= θ∂ sinθ∂ + ∂2+a2sin2θ∂2, sin θ θ sin2θ φ t (r;M,a;∂ ,∂ ,Q)= (r2+a2)2∂2 4aMr∂ ∂ +(∆ a2)∂2+∆Q. (1.6) R t φ − t − t φ − φ We have used some unusual sign conventions in defining to avoid factors of i R when replacing the constants of motion by differential operators. It is clear from the above that ∂ , ∂ , and Q are symmetry operators for the t φ wave equation on Kerr. We denote the set of order-n generators of the symmetry algebra generated by these operators by Sn ={∂tnt∂φnφQnQ|nt+nφ+2nQ =n;nt,nφ,nQ ∈N}. (1.7) In particular, S0 = Id , S1 = ∂t,∂φ . { } { } Of particular importance in our analysis will be the set of second-order symmetry operators, S2 ={∂t2,∂t∂φ,∂φ2,Q}={Sa}, and underlined indices always refer to the index in this set. Thefunction ispolynomialinitslastthreevariables,so (r;M,a;∂ ,∂ ,Q)is t φ R R well defined. Furthermore, it can be written as a linear combination of the second- order symmetries with coefficients which are rational in r, M, and a, (r;M,a;∂ ,∂ ,Q)= aS . t φ a R R 1.3. Strategy of proof and further results. Recallfromearlierintheintroduc- tion that there are three major problems in the Kerr spacetime: (1) No positive, conserved energy: There is no timelike, Killing vector. In particular, the vector field ∂ , which is Killing, is only timelike outside the t ergosphere, r >M +√M2 a2cos2θ. − (2) Lack of sufficient classical symmetries: The higher energies generated the Lie derivatives in the ∂ and ∂ directions do not control enough directions t φ to control Sobolev norms or the function. (3) Complicated trapping: There are null geodesics which orbit the black hole, in the sense that they neither escape to null infinity nor enter the black hole. Since solutions to the wave equation can follow null geodesics for an arbitrarily long time, this presents an obstacle to decay. Furthermore, thereareorbitinggeodesicsoccuringoverarangeofrintheKerrspacetime (with a > 0), which makes the situation more complicated than in the | | Schwarzschildspacetime(a=0), wherethereareonlyorbitinggeodesicsat r =3M. To overcome the first problem, we first observe that the vector ∂ is timelike for t sufficiently large r; that, if a ω = H r2+a2 3Since the Carter operator Q and the Carter constant are closely related, we use the same notationforboth. HIDDEN SYMMETRIES AND DECAY FOR THE WAVE EQUATION ON KERR 9 denotes the angular velocity of the horizon, then the vector ∂ +ω ∂ is null on t H φ thehorizonandtimelikeforsufficientlysmallr >r ; thattheregionswhere∂ and + t ∂ +ω ∂ are timelike overlap when a is sufficiently small; and that both ∂ and t H φ t | | ∂ +ω ∂ are Killing. Thus, if we let t H φ T =∂ +χω ∂ , (1.8) χ t H φ whereχisidentically1forr <r forsomeconstantr ,identically0forr >r +M, χ χ χ and smoothly decreases on [r ,r +M], then, for sufficiently small a, this vector- χ χ field will be timelike everywhere and will be Killing outside the fixed region r ∈ [r ,r +M]. Thus, toprovetheboundednessofthispositive, itwillbesufficientto χ χ controlthebehaviourofsolutionsinthisfixedregionthroughaMorawetzestimate. Toovercomeproblem(2), wenotethatthesecond-orderoperatorQisasymme- tryandisaweaklyellipticoperator. UsingQ,∂2,and∂2 assymmetriestogenerate φ t higher energies, we can control energies of the the spherical Laplacian of ψ. These control Sobolev norms which are sufficiently strong to control ψ 2. | | To handle the complicated trapping, we will use our extension of the vector-field method to include hidden symmetries. To construct a Morawetz multiplier, we wouldliketoconstructavectorfieldwithaweightthatchangessignattheorbiting geodesic, but this is not possible using a classical vector-field. If we introduce = aS =∂2+∂2+Qtogiveusanellipticoperatorandanextra,free,underlined L L a t φ index, we can take as our collection of Morawetz vector fields Aab = zw˜ (a b)∂ , 0 r − R L 1 qab = z ∂ w˜ (a b), A − 2 r R0 L z(cid:16) (cid:16) (cid:17)(cid:17) ˜ a =∂ a , 0 r R ∆R (cid:16) (cid:17) with z and w smooth, positive functions to be chosen. Applying the analogy of the deformation formula, the difference between the energies on one hypersurface and another is E (Σ ) E (Σ )= Pα g d4x. (Aab,qAab) 2 − (Aab,qAab) 1 Z (cid:16)∇α (Aab,qAab)(cid:17)p−| | Ignoring several distracting details, the deformation is of the form 1 z ˜ a ˜ b αβ(∂ S ψ)(∂ S ψ) 0 0 α a β b 2 R R L z1/2 +z1/2∆3/2 ∂ w ˜ a b(∂ S ψ)(∂ S ψ) − r ∆1/2R0 L r a r b (cid:18) (cid:18) (cid:19)(cid:19) 1 + (∂ ∆∂ z(∂ w˜ a)) b(S ψ)(S ψ). r r r 0 a b 4 R L In the first line, one factor of ˜ arises from the wave equation, and the other from 0 R our choice of the Morawetz vector field Aab, which allows us to construct a perfect square to obtain positivity. In the second line, the term involves two derivatives of ˜. Nearthephotonorbits, theconvexitypropertiesof , whichensuredthatthe −R R orbits are unstable, ensure that this term is positive. We are free to choose z and w to get positivity away from the photon orbits. The fourth term is lower-order, since it involves fewer derivatives. For small a, with v denoting terms of the form S ψ, and with our choices of z a and w, the sum of the second and third terms is of the form ∆2 9r2 46Mr+54M2 M (∂ v)2+ − v2 (1.9) r2(r2+a2) r r4 (cid:18) (cid:19) withsmallperturbationsonthecoefficients. Thecoefficientonv2 ispositiveoutside a compact interval in (r , ). As shown in [5], it is sufficient to prove a Hardy + ∞