Wave equation

The wave equation models finite-speed propagation and second-order oscillation.

This is the causal three-dimensional fundamental solution of ${\partial }_t^2-c_0^2{\nabla }^2$. It concentrates on the outgoing sound cone and gives the retarded convolution of acoustic sources.

For $u_{tt}=\Delta u+\alpha u$ with a homogeneous Dirichlet boundary condition, let ${\lambda }_1>0$ be the first Dirichlet Laplacian eigenvalue. On a bounded one-dimensional interval, finite-energy displacement remains bounded for all time and all initial data exactly when $\alpha <{\lambda }_1$. The conserved energy $\frac{1}{2}(\Vert u_t{\Vert }_2^2+\Vert \nabla u{\Vert }_2^2-\alpha \Vert u{\Vert }_2^2)$ controls the gradient and, by the Poincare inequality and one-dimensional integration, the maximum norm. At equality a mode can grow linearly; above it a mode can grow exponentially. Bounded oscillations need not have a long-time limit.

The vector field method commutes a wave equation with spacetime symmetry generators, estimates the resulting commuted wave energies, and converts weighted L2 norms to pointwise decay using a Klainerman-Sobolev inequality. The commutation vector fields for the wave equation include translations, rotations, Lorentz boost vector fields and the scaling vector field. It is useful for almost global existence for wave equations and for small-data global existence under the classical null condition for wave equations.

A commuted wave energy controls derivatives of a solution after applying the commutation vector fields for the wave equation. A convenient equivalent norm is $A_m(t)={\sum }_{|I|\le m}\Vert \partial Z^{I}u(t){\Vert }_2$. The wave energy estimate controls its growth through commuted sources, while the Klainerman-Sobolev inequality gives pointwise decay for low-order derivatives.

In three spatial dimensions, with $Z$ the eleven translations, spatial rotations, boosts and scaling commutation vector fields for the wave equation, a sufficiently decaying smooth function satisfies $|u(t,x)|\le C(1+t+|x|{)}^{-1}(1+|t-|x||{)}^{-1/2}{\sum }_{|I|\le 2}\Vert Z^{I}u(t){\Vert }_2$. This weighted Sobolev inequality converts control of commuted wave energy into decay. It holds independently of any wave equation satisfied by $u$.

A first-order vector field $Z$ is a commutation field for the flat wave equation when $[\square ,Z]$ is zero or a controlled multiple of $\square$. Translations, spatial rotations and Lorentz boost vector fields commute with $\square$; the scaling vector field satisfies $[\square ,S]=2\square$. Repeated commutation preserves the differential order of a semilinear wave equation.

The vector field $S=t{\partial }_t+x\cdot \nabla$ generates simultaneous spacetime dilation. For the flat d'Alembert operator, $[\square ,S]=2\square$. Consequently it commutes with the homogeneous wave equation at the level of its solution set, although its operator commutator is nonzero.

For Minkowski spacetime with unit light speed, $L_i=t{\partial }_i+x_i{\partial }_t$ generates a Lorentz boost. It commutes with the d'Alembert operator and controls derivatives transverse to time slices in the vector field method for wave equations.

The vector field ${\Omega }_{ij}=x_i{\partial }_j-x_j{\partial }_i$ is an infinitesimal spatial rotation. It commutes with the flat d'Alembert operator and supplies angular derivatives in the Klainerman-Sobolev inequality.

The vector fields ${\partial }_t$ and ${\partial }_i$ generate time and space translations and commute with the flat d'Alembert operator. They form the unweighted part of the commutation vector fields for the wave equation.

For a real scalar wave equation, the kinetic and gradient energy is $E_{\mathrm{lin}}(t)=\frac{1}{2}\int (|u_t{|}^2+|\nabla u{|}^2)dx$. A potential term is added for a semilinear wave equation $u_{tt}-\Delta u+V^{\prime }(u)=0$, giving the conserved wave energy $E=E_{\mathrm{lin}}+\int V(u)dx$.

For $y_{tt}=c^2{y}_{xx}$ on $0<x<L$, initially at rest with zero displacement, apply $y(0,t)=a$ for $t>0$ and keep $y(L,t)=0$. A geometric series in the Laplace transform givesHere $H$ is the Heaviside step function, $L,c>0$, and only finitely many terms contribute at any finite time. The second step in each pair is a sign-reversed reflection from the fixed endpoint. The formula solves the wave equation classically away from the fronts and distributionally across them; the midpoint is a square wave of period $2L/c$.

The wave speed $c$ is the propagation speed in a wave equation such as $u_{tt}=c^2{\nabla }^{2}u$.

In three spatial dimensions, the solution with initial position $u_0$ and velocity $u_1$ iswhere $M_t$ is the spherical mean.

The strong Huygens principle says that the three-dimensional wave solution at $(t,x)$ depends only on initial data on the sphere $|y-x|=|t|$. A compactly supported disturbance therefore leaves no tail in the interior of its light cone.

Finite propagation speed means that data outside the backward light cone of a spacetime point cannot affect the solution at that point.

For the homogeneous speed-one wave equation, integrate its wave energy density over $B(x_{\ast },T-t)$. The moving boundary contributes $-e$, while the usual flux contributes $u_t{\partial }_{n}u$. Their sum is $-\frac{1}{2}(u_t-{\partial }_{n}u{)}^2-\frac{1}{2}|{\nabla }_{\tan }u{|}^2\le 0$. Zero Cauchy data in the initial ball therefore force zero solution in its backward light cone.

For Cauchy data prescribed on $x=0$ over $|y|<a<1$, put $S=\mathrm{artanh}a$. The maximal region determined solely by those data isIn the travel-time coordinate for a one-dimensional variable-speed wave equation $(x,s)$ this is the ordinary characteristic diamond $|x|+|s|<S$, as follows from finite propagation speed.

A radiation field is the rescaled limit of a wave along null infinity. For a three-dimensional wave, the outgoing field is the limit of $ru(t,r\omega )$ with $t-r$ fixed, while the incoming field fixes $t+r$.

A scattering map sends incoming asymptotic data to outgoing asymptotic data for the same solution.

The one-dimensional wave equation has left-moving and right-moving travelling-wave solutions. Their superposition is described by the D'Alembert formula and remains valid distributionally for locally integrable profiles.

The d'Alembert operator is the Lorentzian wave operator $\square ={\eta }^{\mu \nu }{\partial }_{\mu }{\partial }_{\nu }$ in flat spacetime, or $\square ={\nabla }^{\mu }{\nabla }_{\mu }$ on scalar fields in curved spacetime.

A semilinear wave equation has the formso its highest-order part is linear while the forcing may depend nonlinearly on the solution and its lower derivatives.

A wave map is a harmonic map with Lorentzian domain: a critical point of the action obtained by contracting the pullback target metric with the domain Lorentzian metric. For target sphere $S^2$ in Euclidean space and $\square =-{\partial }_t^2+\Delta$, its extrinsic equation is $\square \varphi =\varphi (|{\varphi }_t{|}^2-|\nabla \varphi {|}^2)$, subject to $|\varphi |=1$. The derivative contractions are null forms for wave equations. Wave map Cauchy data must include a tangent initial velocity.

Smooth compatible data sufficiently small in high weighted Sobolev norms relative to a constant map produce global smooth wave maps in three and four spatial dimensions. In four dimensions, derivative decay $(1+t{)}^{-3/2}$ is time-integrable and closes commuted wave energy estimates. In three dimensions the weaker $(1+t{)}^{-1}$ decay requires the cancellation of null forms for wave equations, exploited by the vector field method for wave equations. These classical localized-data statements do not assert global regularity from small supercritical energy alone.

For smooth localized compatible wave map Cauchy data and compact Riemannian target, one-dimensional wave maps stay smooth globally. In null coordinates, the equation is $D_v{\varphi }_u=D_u{\varphi }_v=0$. Compatibility of the target metric with the covariant derivative makes $|{\varphi }_u{|}^2$ independent of $v$ and $|{\varphi }_v{|}^2$ independent of $u$. These transported derivative bounds and higher energy estimates prevent finite-time breakdown.

A time-independent wave map is a harmonic map of its spatial domain into the target. For target $S^2$, its equation is $\Delta Q+|\nabla Q{|}^{2}Q=0$. In two dimensions inverse stereographic projection gives a smooth nonconstant example of finite wave map energy $4\pi$.

The conserved energy of a wave map from Minkowski spacetime to a Riemannian target is $E=\frac{1}{2}\int (|{\varphi }_t{|}^2+|\nabla \varphi {|}^2)dx$, with norms measured in the target metric. Contraction of the wave map equation with the velocity and integration by parts gives conservation.

Under ${\varphi }_{\lambda }(t,x)=\varphi (t/\lambda ,x/\lambda )$, the conserved wave map energy scales as ${\lambda }^{n-2}E$. Energy is subcritical for $n=1$, critical for $n=2$, and supercritical for $n\ge 3$. The critical derivative index of the homogeneous Sobolev space is $s_c=n/2$. Small energy alone is not an appropriate general small-data regularity hypothesis in supercritical dimensions.

Initial position ${\varphi }_0$ for a wave map takes values in the target, and initial velocity ${\varphi }_1$ belongs to its tangent space at ${\varphi }_0(x)$. For $S^2$ this means $|{\varphi }_0|=1$ and ${\varphi }_0\cdot {\varphi }_1=0$. Localized data agree with a constant map and zero velocity outside a compact set; the map itself need not be a compactly supported ambient function.

A quadratic derivative expression is a null form when its symbol vanishes on parallel null covectors. The Lorentz contraction $Q_0(f,g)=-f_t{g}_t+\nabla f\cdot \nabla g$ is the basic example. Its cancellation suppresses interactions of derivatives pointing along the same light ray. It is the algebraic structure behind the classical null condition for wave equations.

For a quadratic derivative source $A^{\alpha \beta }{\partial }_{\alpha }u{\partial }_{\beta }u$, the classical null condition is $A^{\alpha \beta }{\ell }_{\alpha }{\ell }_{\beta }=0$ whenever $-{\ell }_0^2+|{\ell }_{\mathrm{sp}}{|}^2=0$. It removes the leading interaction of parallel lightlike derivatives. In three dimensions this structure gives global smooth solutions for sufficiently small, localized data through the vector field method for wave equations. This PDE condition is distinct from the null condition asserting that a single vector is lightlike.

The focusing power wave equation is $u_{tt}-\Delta u=u|u{|}^{p-1}$. Its conserved energy has the negative potential term $-|u{|}^{p+1}/(p+1)$ and is not positive definite. Localized ordinary differential equation blowup for a wave equation produces smooth compactly supported data with finite maximal lifespan for suitable focusing powers.

A spatially constant blowup profile for a wave equation solves an ordinary differential equation. Choose smooth compactly supported initial data which agree with the profile on a ball larger than its blowup time times the wave speed. Finite propagation speed preserves agreement in an interior light cone, forcing finite-time breakdown of the localized solution. For $u_{tt}-\Delta u=u^3$, the profile $u(t)=\sqrt{2}/(1-t)$ is one example.

For a smooth semilinear wave equation whose nonlinearity is a smooth function of the solution and its first derivatives, local Sobolev space theory gives a continuation criterion: a finite maximal existence time requires divergence of a sufficiently high solution norm. A bounded base norm $H^{s+1}\times H^s$, with $s>n/2$, allows restarting local existence. Higher regularity persists by differentiated energy estimates. Precise norms can be weakened for particular nonlinearities.

Almost global existence means a lifespan which grows exponentially in the inverse size of small initial data, typically $T_{\epsilon }\ge \exp (c/\epsilon )-1$ for derivative-quadratic semilinear wave equations in three dimensions. Commuted wave energy and the Klainerman-Sobolev inequality lead to a logarithmic accumulation $\epsilon \log (1+T)$ in a bootstrap argument. This implies existence up to every fixed inverse power ${\epsilon }^{-N}$ when the data are small enough depending on $N$.

The defocusing power-type wave equation has the conserved nonnegative energy

For $u_{tt}-\Delta u+u^3=0$ in three spatial dimensions, conserved positive wave energy $E$ controls $\Vert \nabla u{\Vert }_2$. The Sobolev inequality gives $\Vert u^2\nabla u{\Vert }_2\lesssim \Vert \nabla u{\Vert }_2^2\Vert D^{2}u{\Vert }_2\lesssim E\Vert D^{2}u{\Vert }_2$. Differentiated wave energy estimates and the Gronwall inequality then give $\Vert u(t){\Vert }_{{\dot{H}}^2}+\Vert u_t(t){\Vert }_{{\dot{H}}^1}\le CD{e}^{CEt}$, where $D$ bounds the corresponding initial Sobolev norms.

Multiplying a defocusing wave equation by $\nabla \psi \cdot \nabla u+(\Delta \psi )u/2$ and integrating by parts gives a weighted spacetime identity. For a radial solution and radial weight, its coercive terms are

In three dimensions, $\Delta |x|=2/|x|$ away from the origin and $-\Delta (1/|x|)=4\pi {\delta }_0$. Hence $-{\Delta }^2|x|=8\pi {\delta }_0$ as a distribution.

For a finite-energy defocusing power-wave solution in three dimensions,Hardy's inequality bounds the Morawetz action by the conserved energy.

For a radial function $u(t,r)$ in three dimensions, setting $w=ru$ removes the radial first derivative because $r\Delta u=w_{rr}$. The defocusing power equation becomeson the half-line.

In three dimensions set $w=ru$ for a radial function satisfying $\square u=F$. If the source has support in $s\le r\le s+1$ and $|F(s,r)|\le (1+s{)}^{-2}$, the Duhamel principle bounds its contribution to $w(t,t+a)$ by $\frac{1}{2}\log (1+t)$ for $1/2\le a\le 1$. Its backward characteristic interval meets the source in a set of length at most one, and $r|F|\le (1+s{)}^{-1}$. Division by $r\simeq 1+t$ gives $|u|\lesssim \log (2+t)/(1+t)$ for compactly supported initial data.

For $q=w_t+w_r$ and a radial weight $\psi$,

Suppose the inhomogeneous linear wave equation has an energy estimate in a Banach solution space $X_T$, and its nonlinearity obeystogether with the corresponding locally Lipschitz estimate. For a sufficiently large closed ball and sufficiently small $T$, the map that solves the linear equation with source $F(w)$ is a contraction. The contraction mapping theorem then gives a local weak solution.

Small displacements in an isotropic elastic solid satisfy a vector wave equation governed by density and the two Lamé moduli.

In a homogeneous elastic layer $-H<z<H$ with rigid boundaries, in-plane propagating modes mix P- and SV-waves. For phase speed $c>c_P>c_S$, one even-horizontal, odd-vertical family obeys

A shear-horizontal wave mode has ${\omega }_n(k)=c_s\sqrt{k^2+(n\pi /h{)}^2}$. At $x=Vt$ with $0<|V|<c_s$, a nondegenerate stationary Fourier phase gives generic $t^{-1/2}$ decay. Faster rays have no stationary point; compactly supported data give eventual zero by finite propagation speed.

Divergence of the elastic equation isolates dilatational motion, while curl isolates rotational motion, producing P and S wave equations with different speeds.

A P-wave is longitudinal, carries nonzero dilatation, and has speed $c_P=\sqrt{(\lambda +2\mu )/\rho }$ in an isotropic solid.

In a longitudinal plane wave, displacement is parallel to the wavevector and the curl vanishes.

An S-wave is transverse, carries rotation without dilatation, and has speed $c_S=\sqrt{\mu /\rho }$.

In a transverse plane wave, displacement is perpendicular to the wavevector and its divergence vanishes.

An SV-wave is polarized in the vertical plane containing its wavevector and perpendicular to that wavevector.

An SH-wave is polarized horizontally and perpendicular to the vertical propagation plane.

For a layer $0<z<h$ with rigid lower boundary and traction-free upper boundary, an SH mode
$u_y=Y(z)e^{i(\kappa x-\omega t)}$ hasand dispersionIts cutoff is ${\omega }_n=c_s{q}_n$, its phase speed is $\omega /\kappa$, and its group speed is $c_s^2\kappa /\omega$.

At a perfectly bonded interface, all displacement and traction components are continuous.

An obliquely incident in-plane P-wave generally produces reflected and transmitted P and SV waves; SH polarization decouples.

Phase matching fixes a common frequency and tangential wavenumber for every incident, reflected, and transmitted elastic mode.

For two inviscid media, continuity of normal displacement and normal stress determines the reflected and transmitted longitudinal amplitudes.

Equal normal velocities and the membrane tension-mass balance determine complex pressure amplitudes. With identical fluids on both sides the reflected and transmitted squared amplitudes sum to one, expressing conserved mean energy flux.

Equating the absolute reflected and transmitted displacement amplitudes, then using acoustic Snell's law, gives an algebraic condition on incidence angle and the two moduli-over-speed impedances.

A stretched uniform string obeys y_tt=c^2 y_xx with wave speed equal to the square root of tension over density.

Zero values at both endpoints of a length-$L$ string allow the D'Alembert formula to be written with opposite travelling waves and a $2L$-periodic profile. This represents reflection with sign reversal at each fixed endpoint. For positive-time boundary conditions, periodicity concerns the arguments visited by the string solution; unused profile values can be defined by periodic extension.

With fixed endpoints, $y_{tt}-y_{xx}+y_t=0$ separates into sine modes whose amplitudes satisfy

Critical damping occurs when a second-order mode has a repeated real characteristic root. It returns to equilibrium without oscillating and as quickly as possible among nonoscillatory regimes.

For initial sine coefficient $a_n$ and zero initial velocity, put ${\omega }_n^2=n^2{\pi }^2-1/4$. The modal amplitude is

For fixed endpoints andintegration by parts and the damped wave equation give

A normal mode oscillates at one frequency with a fixed spatial eigenfunction.

The temporal frequency of a normal mode. In $\dot{z}=iAz$, a real eigenvalue of $A$ is a signed eigenfrequency; its positive imaginary part after damping gives exponential decay in $e^{i\lambda t}$.

Orthogonality separates total quadratic wave energy into a sum of independent modal energies.

A node is a point where a standing-wave eigenfunction vanishes at all times.

Fourier transformation in space turns $u_{tt}=c^2{u}_{xx}$ into the independent oscillators ${\hat{u}}_{tt}+c^2{k}^2\hat{u}=0$.

For zero initial displacement and unit delta initial velocity, the spatial Fourier transform of the wave equation gives the displayed multiplier. A compact cutoff function around zero frequency produces an ordinary smooth function after inversion. The remaining sine splits into two oscillatory integrals with phases $x\cdot \xi \pm ct|\xi |$ and symbols proportional to $(1-\chi )/|\xi |$ of order $-1$. Their stationary directions lie on $|x|=c|t|$, proving a light-cone bound for singular support. This does not claim vanishing of smooth tails inside the cone.

For real wave speed $c$ and $t\ge 0$, define the complex-frequency cosine through the entire series $C_t(z)={\sum }_{j\ge 0}(-1{)}^j(ct{)}^{2j}(z\cdot z{)}^j/(2j)!$. The expression is independent of the square-root branch even though the square root itself need not be entire. Writing $z=a+ib$ gives $|\mathrm{Im}\sqrt{z\cdot z}|\le |b|$, and hence $|C_t(z)|\le e^{|c|t|\mathrm{Im}z|}$. The Paley–Wiener–Schwartz theorem therefore makes its inverse Fourier transform a distribution supported in the radius-$|c|t$ ball. Multiplication by $C_t$ evolves initial displacement with zero initial velocity in the wave equation, giving finite propagation speed for compactly supported initial distributions.

The Cauchy problem $u_{tt}=c^2{u}_{xx}$ has

A rectangular initial displacement with zero initial velocity splits into two equal half-height copies moving in opposite directions. The D'Alembert formula reads $u(x,t)=[u_0(x+ct)+u_0(x-ct)]/2$. Overlap adds the heights; separated copies each retain the original width. For discontinuous fronts the formula defines a weak solution of the wave equation.

For $u_{tt}-u_{xx}=0$ with $u(0,x)=0$ and $u_t(0,x)=f(x)$,

If $u(x,0)=0$, $u_t(x,0)=g(x)$, and the odd function $g$ is supported on $[-a,a]$, thenThe interval $[x-t,x+t]$ then contains the entire support, whose integral vanishes by oddness.

The wavenumber is the spatial angular frequency of a monochromatic wave. A factor $e^{ikx}$ has wavelength $2\pi /|k|$.

The wavelength is the spatial period of a monochromatic wave. For angular wavenumber $k$, it is $\lambda =2\pi /|k|$.

A dispersion relation gives angular frequency $\omega$ as a function of wavenumber $k$ for monochromatic waves admitted by a linear wave equation.

The response to a localized perturbation must be followed in space as well as time. Growth at a fixed position is absolute wave-packet instability; growth transported away while the original position decays is convective wave-packet instability. The Briggs-Bers criterion evaluates the relevant complex spatial branches of a dispersion relation, while a real-wavenumber temporal mode test alone does not generally distinguish those behaviors.

A criterion for absolute wave-packet instability based on analytic continuation of the spatial roots of a dispersion relation. Start the temporal inversion contour above the spectrum, then lower it while deforming the spatial contour. A spatial pinch point occurs when branches originating in opposite spatial half-planes obstruct that deformation. A double root $D=D_k=0$ is only an algebraic candidate. The opposite-half-plane condition is stated in the primary study doi.org/10.1017/jfm.2016.195, section 4.2.

A collision of spatial branches which pinches the spatial inversion contour as the temporal contour is lowered from above all singularities. The branches must originate on opposite sides of the spatial contour. A collision of two upper-half-plane branches or two lower-half-plane branches can solve the double-root equations without causing absolute wave-packet instability.

The real dispersion above is temporally stable for every real $k$, but its analytic continuation has stationary points at $k=\pm i$ with $\omega =i/2$. Writing $\omega =i\Omega$ gives $k^2=-1\pm \sqrt{1/2-\Omega }$; as $\Omega$ decreases to $1/2$, both branches at $+i$ approach from the upper half-plane and both at $-i$ from the lower. These are not spatial pinch points. This provides a concrete counterexample to treating growing algebraic double roots as sufficient for absolute wave-packet instability.

A localized disturbance grows along some moving rays but decays at its original fixed location. This wave-packet meaning differs from buoyancy-driven stellar convection and is therefore given a distinct canonical title. Changing the observation frame can change the absolute/convective classification. The Briggs-Bers criterion uses the laboratory-frame impulse response to distinguish this case from absolute wave-packet instability.

A localized perturbation has positive asymptotic exponential growth at a fixed observation point in the chosen reference frame. The relevant spatial pinch point of the Briggs-Bers criterion has positive imaginary frequency. Merely finding a growing temporal mode or a formal complex saddle is not sufficient; the saddle must contribute to the localized impulse response.

With normal modes $e^{ikx-i\omega t}$, a finite upper bound on $\mathrm{Im}\omega$ allows the inverse temporal Laplace transform contour to start above all temporal singularities. An unbounded growth rate at high real wavenumber prevents the standard causal Green function construction used in the Briggs-Bers criterion. For $\mathrm{Im}\omega =\alpha k^2-\beta k^4-\gamma$ with nonzero real $\beta$, the bound holds exactly when $\beta >0$.

Wave dispersion means that phase velocity depends on wavenumber or, equivalently, wavelength. Wave packets generally travel with group velocity, which need not equal phase velocity. The linear rotating shallow-water dispersion relation ${\omega }^2=f^2+c^2{k}^2$ has faster long-wave phase motion but slower long-wave packet motion when $f\ne 0$.

A nondispersive wave has constant phase velocity on a propagation branch. For $\omega =ck$, its group velocity is also $c$. A frequency relation with a nonzero offset, such as $\omega =ck+b$, has constant group velocity but a phase velocity $c+b/k$ depending on wavenumber; constant group velocity alone is therefore not sufficient for nondispersion.

A dispersion diagram plots the branches of a dispersion relation, usually with wavenumber on the horizontal axis and angular frequency on the vertical axis. Its slopes, intersections, and gaps display group velocities, mode conversion, and forbidden frequency ranges.

Ray tracing follows the paths along which a slowly modulated wave packet propagates. For a local dispersion relation $\omega =\Omega (\mathbf{k};\mathbf{x},t)$, the ray velocity is the group velocity ${\nabla }_{\mathbf{k}}\Omega$.

A local dispersion relation $\omega =\Omega (\mathbf{k};\mathbf{X},T)$ yieldsThe spatial and temporal derivatives on the right hold the wavevector fixed. Translation symmetries conserve the corresponding wavevector components, while time independence conserves frequency.

A simple turning point for a stratified wavespeed occurs at $c(Y_s)=\omega /k$, where $Q={\omega }^2/c^2-k^2$ has a nonzero derivative. If $c_Y(Y_s)>0$, write $q=-Q^{\prime }(Y_s)>0$; then $\ell \sim [q(Y_s-Y){]}^{1/2}$ below it. The vertical group velocity vanishes and the stationary wave amplitude in a stratified wavespeed diverges as $(Y_s-Y{)}^{-1/4}$.

In ${\epsilon }^2{\Psi }^{\prime \prime }+2{\epsilon }^2(c^{\prime }/c){\Psi }^{\prime }+Q(Y)\Psi =0$, with $Q\sim -q(Y-Y_s)$ and $q>0$, set $Y-Y_s=({\epsilon }^2/q{)}^{1/3}\eta$. The leading equation is ${\Psi }_{\eta \eta }-\eta \Psi =0$, the Airy ordinary differential equation. The first derivative and nonlinear coefficient corrections are smaller by $O({\epsilon }^{2/3})$.

The local inner solution is a combination of Airy functions $\mathrm{Ai},\mathrm{Bi}$. Decay in the forbidden region removes $\mathrm{Bi}$. The remaining Airy turning-point connection formula is oscillatory on the allowed side and represents an incident/reflected pair, not a single upward branch. Without the forbidden-side condition the local differential equation alone leaves both coefficients free.

A simple ray turning point has outer amplitude proportional to $|Y-Y_s{|}^{-1/4}$. At the Airy scaling at a variable-speed wave turning point, $|Y-Y_s|=O({\epsilon }^{2/3})$, this gives an inner wavefield of size $O({\epsilon }^{-1/6})$ when the incident action normalization is order one. The enhancement is finite for nonzero $\epsilon$ and is resolved by the Airy function.

For the zero-frequency dispersion relationwrite $\mathbf{k}=\kappa (\cos \varphi ,\sin \varphi )$. Thenwhere $\psi$ is the direction of the group velocity. Maximizing the angular slope gives a downstream wake wedge of semi-angle

For a normal mode proportional to $e^{\sigma t}$, the real part $\mathrm{Re}\sigma$ is its exponential growth rate. A positive growth rate signals linear instability.