2. Framework

We consider an incompressible fluid undergoing differential rotation about the vertical axis $Oz$ , with an angular velocity $\varOmega (r)$ that depends only on the cylindrical radial coordinate $r$ . The fluid is also assumed to be stably stratified in a central gravitational field, with a Brunt–Väisälä frequency $N(\rho )$ that depends solely on the spherical radial coordinate $\rho$ . The kinematic viscosity $\nu$ and the scalar diffusivity $\alpha$ are both taken to be constant. The model chosen for the stratification and angular rotation mimics configurations that can be found in stellar interiors (e.g. Dintrans et al. Reference Dintrans, Rieutord and Valdettaro1999). It can be rigorously justified provided that the centrifugal force remains negligible compared with gravity. However, this is not our purpose here, as other models can be adapted to the following derivation without difficulty.

We are interested in harmonic, axisymmetric perturbations of frequency $\varpi$ ( $\varpi \gt 0$ ). These perturbations are described by the velocity $\boldsymbol{U}$ , pressure $P$ and buoyancy $B$ , which are expressed in the form

(2.1) \begin{equation} (\boldsymbol {U}, P, B) = (\boldsymbol{u}, p, b)\textrm{e}^{-\textrm{i} \varpi t} +c.c., \end{equation}

where ‘c.c.’ denotes the complex conjugate. The governing equations for the amplitudes $\boldsymbol{u}$ , $p$ and $b$ are derived from the Boussinesq approximation of the Navier–Stokes equations. After non-dimensionalising all quantities using a characteristic length scale $L$ and a characteristic velocity scale $V$ , the linearised equations take the form

(2.2a) \begin{align} -{\textrm{i}} \varpi \boldsymbol{u} + 2 \varOmega (\hat {\boldsymbol{e}}_z \times \boldsymbol{u}) + r (\boldsymbol{u} \boldsymbol{\cdot }\boldsymbol{\nabla }\varOmega ) \hat {\boldsymbol{e}}_\phi & = -\boldsymbol{\nabla }p + b\, \hat {\boldsymbol{e}}_\rho + E{\nabla} ^2 \boldsymbol{u} , \end{align}

(2.2b) \begin{align} \boldsymbol{\nabla }\boldsymbol{\cdot }\boldsymbol{u} & = 0 , \end{align}

(2.2c) \begin{align} -{\textrm{i}} \varpi b + N^2 \boldsymbol{u}\boldsymbol{\cdot }\hat {\boldsymbol{e}}_\rho & = (E/Pr) {\nabla} ^2 b , \end{align}

where $Pr=\nu /\alpha$ is the Prandtl number and $E=\nu /(VL)$ the Ekman number. In these equations, $\varOmega$ and $N$ now denote the rotation rate and the Brunt–Väisälä frequency non-dimensionalised by $V/L$ . In the following, we assume $E\ll 1$ and $Pr \geqslant O(1)$ . The cylindrical and spherical vector bases are denoted by $(\hat {\boldsymbol{e}}_r, \hat {\boldsymbol{e}}_\phi , \hat {\boldsymbol{e}}_z)$ and $(\hat {\boldsymbol{e}}_\rho , \hat {\boldsymbol{e}}_\theta , \hat {\boldsymbol{e}}_\phi )$ , respectively.

2.1. Characteristics and local dispersion relation of the non-dissipative problem

In the non-dissipative limit ( $\nu =\alpha =0$ ), the system (2.2) can be reduced to a single second-order equation for the pressure $p$ (see, for instance, Friedlander & Siegmann Reference Friedlander and Siegmann1982; Mirouh et al. Reference Mirouh, Baruteau, Rieutord and Ballot2016). Retaining only second-order derivatives, this equation takes the form

(2.3) \begin{equation} \big(\varpi ^2 - N^2 \sin ^2\theta \big)\displaystyle \frac {\partial ^2p}{\partial r^2}+ 2N^2\cos \theta \, \sin \theta \frac {\partial ^2 p}{\partial z\partial r} + \big(\varpi ^2 - N^2 \cos ^2\theta - \xi ^2 \big)\displaystyle \frac {\partial ^2p}{\partial z^2} = 0, \end{equation}

where $\xi$ is the epicyclic frequency, defined by

(2.4) \begin{equation} \xi ^2(r) = 2 \varOmega (r) \omega (r), \end{equation}

and $\omega (r) = (1/r)\partial _r(r^2\varOmega )$ is the background vorticity. Equation (2.3) is a generalisation of the classical Poincaré equation, extended to account for both differential rotation and radial stratification. In the case of solid-body rotation and uniform stratification, it reduces to the standard form with $\xi ^2 = 4 \varOmega ^2$ .

This equation can also be derived from the full system (2.2) in the inviscid and non-diffusive limit ( $\nu =\alpha =0$ ) by assuming a local plane-wave solution of the form $\exp ({\textrm{i}} k_r r + {\textrm{i}} k_z z)$ with $k_r\gg 1$ and $k_z\gg 1$ . This leads to the local dispersion relation for inertia-gravity waves

(2.5) \begin{equation} \varpi ^2 = N^2\frac {(k_z \cos \theta - k_r \sin \theta )^2}{k^2} + \xi ^2 \frac {k_z^2}{k^2} , \quad k^2 = k_r^2 + k_z^2, \end{equation}

which corresponds to (2.3) under the substitution $\partial _r$ by ${\textrm{i}} k_r$ and $\partial _z$ by ${\textrm{i}} k_z$ . We assume a stable configuration in which both $N^2$ and $\xi ^2$ are positive. Equation (2.5) determines the allowable range of wave frequencies. At the equator ( $\theta =0$ ), the frequency ranges from 0 to $\sqrt {N^2+\xi ^2}$ , while at the poles ( $\theta = \pm \pi /2$ ), it ranges from $\min (N,\xi )$ to $\max (N;\xi )$ , depending on the relative magnitudes of $N$ to $\xi$ . From the dispersion relation (2.5), we can define both the phase velocity and group velocity. The phase velocity is given by

(2.6) \begin{align} \boldsymbol {v}^{(\varphi )} = \left(v^{(\varphi )}_r,v^{(\varphi )}_z \right)= \left (\frac {\varpi k_r}{ k^2}, \frac {\varpi k_z}{ k^2}\right ), \end{align}

while the group velocity is

(2.7) \begin{equation} \boldsymbol {v}^{(g)} =\left(v^{(g)}_{r},v^{(g)}_{z}\right) = \left (\displaystyle \frac {\partial \varpi }{\partial k_r},\displaystyle \frac {\partial \varpi }{\partial k_z}\right ), \end{equation}

with explicit expressions

(2.8) \begin{equation} v^{(g)}_{r}= - k_z\frac {N^2 k_\rho k_\theta + \xi ^2 k_r k_z}{\varpi } , \quad v^{(g)}_{z}= k_r\frac {N^2 k_\rho k_\theta + \xi ^2 k_r k_z}{\varpi }, \end{equation}

where

(2.9) \begin{align} k_\rho = k_r \cos \theta + k_z\sin \theta , \quad k_\theta = - k_r \sin \theta + k_z \cos \theta . \end{align}

The paths of characteristics associated with (2.3) are given by

(2.10a) \begin{align} \frac {\textrm{d}z}{\textrm{d}r} & = \frac {N^2 \sin \theta \cos \theta \pm \sqrt {\varDelta }}{\varpi ^2 - N^2\sin ^2\theta } , \end{align}

(2.10b) \begin{align} \frac {\textrm{d}r}{\textrm{d}z} & = \frac {N^2 \sin \theta \cos \theta \mp \sqrt {\varDelta }}{\varpi ^2 - N^2\cos ^2\theta -\xi ^2} , \end{align}

where

(2.11) \begin{equation} \varDelta = \varpi ^2\big(N^2+\xi ^2 - \varpi ^2\big) - N^2\xi ^2 \sin ^2\theta . \end{equation}

The curves defined by the vanishing of $\varDelta$ correspond to turning points, which play a key role in the wave dynamics. As explained in Friedlander & Siegmann (Reference Friedlander and Siegmann1982), these curves separate regions where (2.3) is hyperbolic ( $\varDelta \gt 0$ ) from those where it is elliptic ( $\varDelta \lt 0$ ). Only the hyperbolic regions ( $\varDelta \geqslant 0$ ) are relevant for wave propagation, as they admit characteristic paths. In a non-stratified fluid ( $N=0$ ), turning points occur at cylindrical locations where $\varpi = \xi (r)$ . At these points, the ray path becomes horizontal. Conversely, in a non-rotating fluid ( $\varOmega =0$ ) or in a irrotational flow ( $\omega =0$ ), the turning points lie at the spherical locations where $\varpi = N(\rho )$ . At these points, the ray path becomes radial.

Moreover, the slope of the characteristic can be related to the group velocity. From (2.8), we have

(2.12) \begin{equation} \frac {v^{(g)}_{z}}{v^{(g)}_r} = - \frac {k_r}{k_z} = \frac {N^2 \sin \theta \cos \theta \pm \sqrt {\varDelta }}{\varpi ^2 - N^2\sin ^2\theta } , \end{equation}

indicating that the group velocity points in the direction of the characteristic paths. Using the dispersion relation (2.5), the expression (2.11) for $\varDelta$ can also be rewritten as

(2.13) \begin{equation} \varDelta = \big(N^2 k_\rho k_\theta + \xi ^2 k_r k_z\big)^2. \end{equation}

This expression implies that the group velocity (2.8) vanishes precisely when $\Delta =0$ , i.e. at turning points.

2.2. Local Frenet–Serret frame

Consider a characteristic path $\mathcal{C}$ parametrised by an arclength $x_\parallel \gt 0$ that increases in the direction of wave propagation. At any point along the path, the position vector is given by

(2.14) \begin{align} \boldsymbol{x} _{0}(x_\parallel ) =r_0(x_\parallel ) \hat {\boldsymbol{e}}_r +z_0(x_\parallel ) \hat {\boldsymbol{e}}_z . \end{align}

We define a local Frenet–Serret frame $(\hat {\boldsymbol{e}}_\parallel , \hat {\boldsymbol{e}}_\perp )$ in the $(r,z)$ plane as follows:

(2.15) \begin{equation} \displaystyle \frac {\partial \boldsymbol{x}_0}{\partial x_\parallel } = \hat {\boldsymbol{e}}_\parallel (x_\parallel ) = {r^{\prime}_0} (x_\parallel ) \hat {\boldsymbol{e}}_r + {z^{\prime}_0}(x_\parallel ) \hat {\boldsymbol{e}}_z , \quad \hat {\boldsymbol{e}}_{\perp } = -{z^{\prime}_0}(x_\parallel ) \hat {\boldsymbol{e}}_r + {r^{\prime}_0} (x_\parallel )\hat {\boldsymbol{e}}_z . \end{equation}

The vectors $\hat {\boldsymbol{e}}_{\parallel }$ and $\hat {\boldsymbol{e}}_{\perp }$ are orthonormal and satisfy the standard Frenet–Serret relations

(2.16) \begin{equation} \displaystyle \frac {\partial \hat {\boldsymbol{e}}_\parallel }{\partial x_\parallel } = \kappa (x_\parallel ) \hat {\boldsymbol{e}}_{\perp } (x_\parallel ) , \quad \displaystyle \frac {\partial \hat {\boldsymbol{e}}_\perp }{\partial x_\parallel } = -\kappa (x_\parallel ) \hat {\boldsymbol{e}}_{\parallel } (x_\parallel ) , \end{equation}

where $\kappa (x_\parallel )$ is the curvature of the characteristic path $\mathcal{C}$ at the point $\boldsymbol{x}_0(x_\parallel )$ . It is given by

(2.17) \begin{equation} \kappa (x_\parallel ) = {z^{\prime \prime}_0}(x_\parallel ) {r^{\prime}_0}(x_\parallel )- {r^{\prime \prime}_0}(x_\parallel ) {z^{\prime}}_0(x_\parallel ) . \end{equation}

The characteristic path $\boldsymbol{x_0}(x_\parallel )$ satisfies the dispersion relation, which can be written as

(2.18) \begin{equation} \varpi ^2 = \xi ^2_0 {r_0}^{\prime 2} + N^2_0 \displaystyle \frac {\big(r_0 r^{\prime}_0 + z_0 z^{\prime}_0\big)^2}{\rho _0^2} , \end{equation}

where $\rho _0=\sqrt {r_0^2 + z_0^2}$ is the local spherical radius, and $\xi _0=\xi (r_0)$ , $N_0=N(\rho _0)$ . The function $\varDelta (x_\parallel )$ can also be expressed along the characteristic path using (2.11) as

(2.19) \begin{equation} \varDelta (x_\parallel ) =\varpi ^2\big(N^2_0+\xi ^2_0 - \varpi ^2\big) - N^2_0\xi ^2_0 z_0^2/\rho _0^2 . \end{equation}

Another useful expression for $\varDelta$ , derived from the dispersion relation (2.18), is

(2.20) \begin{equation} \varDelta (x_\parallel ) = \frac {\left ( N^2_0(r_0 {r^{\prime}_0} + z_0 {z^{\prime}_0})(r_0 {z^{\prime}_0} - z_0 {r^{\prime}_0}) + \xi ^2_0 {r^{\prime}_0}{z^{\prime}_0}\rho _0^2 \right )^2 }{\rho _0^4} . \end{equation}

This allows us to write

(2.21) \begin{equation} \epsilon _1 \sqrt {\varDelta } = \frac { N^2_0 (r_0 {r^{\prime}_0} + z_0 {z^{\prime}_0})(r_0 {z^{\prime}_0} - z_0 {r^{\prime}_0}) }{\rho _0^2}+ \xi ^2_0 {r^{\prime}_0}{z^{\prime}_0}, \end{equation}

where $\epsilon _1= \textrm {sgn}[ N^2_0 (r_0 {r^{\prime}_0} + z_0 {z^{\prime}_0})(r_0 {z^{\prime}_0} - z_0 {r^{\prime}_0}) + \xi ^2_0 {r^{\prime}_0}{z^{\prime}_0}\rho _0^2]$ . From this, we can derive the following additional expressions:

(2.22a) \begin{align} {z^{\prime}_0}/{r^{\prime}_0} & = \frac {N^2_0 r_0 z_0/\rho _0^2 + \epsilon _1 \sqrt {\varDelta } }{\varpi ^2 -N^2_0 z_0^2/\rho _0^2} , \end{align}

(2.22b) \begin{align} {r^{\prime}_0}/{z^{\prime}_0} &= \frac {N^2_0 r_0 z_0/\rho _0^2 - \epsilon _1 \sqrt {\varDelta } }{\varpi ^2 -\xi _0^2 -N^2_0 r_0^2/\rho _0^2} . \end{align}

The characteristic path $\mathcal{C}$ may reach a turning point $\boldsymbol {x}_c= (r_c,z_c)$ at a specific value $x_{\parallel c}$ of the arclength parameter $x_\parallel$ , where $\varDelta =0$ . At this point, the characteristic path terminates and reflects, as illustrated in figure 1. It exhibits a generic cusp structure, which can be explicitly calculated under the assumption that the root of $\Delta$ is simple, that is,

(2.23) \begin{align} \varDelta (x_{\parallel c})=\varDelta _c=0 \quad \textrm{and} \quad \Delta ^{\prime}(x_{\parallel c})=\Delta ^{\prime}_c \neq 0. \end{align}

Figure 1.

(a) Frames associated with the characteristic path close to a turning point $\boldsymbol {x}_c$ . (b) Scaling of the turning-point region for a viscous beam.

Let us now investigate the form of the path in the local frame $(\hat {\boldsymbol{e}}_{\parallel c}, \hat {\boldsymbol{e}}_{\perp c})$ attached to the turning point $\boldsymbol {x}_c$ (see figure 1 a). If we express the path as $\boldsymbol {x}_0= \boldsymbol {x}_c + y_{\parallel 0} \hat {\boldsymbol{e}}_{\parallel c} + y_{\perp 0} \hat {\boldsymbol{e}}_{\perp c}$ , the coordinates $(y_{\parallel 0}, y_{\perp 0})$ satisfy, for $x_\parallel$ close to $x_{\parallel c}$

(2.24) \begin{equation} y_{\parallel 0} \sim -|x_\parallel - x_{\parallel c}| , \quad y_{\perp 0} \sim \pm \frac {2}{3}\alpha _c |x_\parallel - x_{\parallel c} |^{3/2} , \end{equation}

which leads to the following relation between $y_{\perp 0}$ and $y_{\parallel 0}$ :

(2.25) \begin{equation} y_{\perp 0} \sim \pm \frac {2}{3}\alpha _c (-y_{\parallel 0} )^{3/2}. \end{equation}

Here, $\alpha _c$ is given by

(2.26) \begin{equation} \alpha _c = -\frac {\epsilon _1 z_c^{\prime 2}\sqrt {-\varDelta ^{\prime}_c}}{\varpi ^2 - \xi _c^2 - N^2_cr_c^2/\rho _c^2} = -\frac {\epsilon _1 r_c^{\prime 2}\sqrt {-\varDelta ^{\prime}_c}}{\varpi ^2 - N^2_cz_c^2/\rho _c^2} = -\frac {\epsilon _1\sqrt {-\varDelta ^{\prime}_c}}{2\varpi ^2 - N^2_c -\xi ^2_c} , \end{equation}

where the subscript $c$ denotes values evaluated at the turning point $x_{\parallel c}$ . In (2.24b ) and (2.25), the upper sign corresponds to the incident beam ( $x_\parallel \lt x_{\parallel c}$ ) while the lower sign corresponds to the reflected beam ( $x_\parallel \gt x_{\parallel c}$ ). If $N_c^2 r_c z_c\neq 0$ , none of the denominators in (2.26) vanish, since

(2.27) \begin{align} \big(\varpi ^2 - \xi _c^2 - N^2_cr_c^2/\rho _c^2\big)\big(\varpi ^2 - N^2_cz_c^2/\rho _c^2\big) = N^4_cr_c^2z_c^2/\rho _c^4 \gt 0 , \end{align}

and

(2.28) \begin{align} 2\varpi ^2 - \xi _c^2 - N^2_c= \big(\varpi ^2 - \xi _c^2 - N^2_cr_c^2/\rho _c^2\big) + \big(\varpi ^2 - N^2_cz_c^2/\rho _c^2\big) . \end{align}

If $N_c^2 r_c z_c = 0$ , then either $r^{\prime}_c=0$ and $\varpi ^2- \xi _c^2- N_c^2r_c^2/\rho _c^2=0$ , or $z^{\prime}_c=0$ and $\varpi ^2- N_c^2z_c^2/\rho _c^2=0$ . In either case, the third expression of $\alpha _c$ remains valid.

It is important to note that the vectors $\hat {\boldsymbol{e}}_\perp$ and $\hat {\boldsymbol{e}}_\parallel$ associated with the reflected beam do not match $\hat {\boldsymbol{e}}_{\perp c}$ and $\hat {\boldsymbol{e}}_{\parallel c}$ as $x_\parallel \rightarrow x_{\parallel c}$ . As a result, the functions ${r^{\prime}_0}$ and ${z^{\prime}_0}$ are not continuous across $x_{\parallel c}$ . Throughout this analysis, all quantities evaluated at $x_{\parallel c}$ should be understood as taken in the limit $ x_\parallel \rightarrow x_{\parallel c}^-$ , i.e. for the incident beam.

Finally, it is interesting to note that the curvature $\kappa$ also diverges at $\boldsymbol {x}_c$ as

(2.29) \begin{equation} \kappa \sim - \frac {\alpha _c}{2\sqrt { |x_{\parallel c} - x_\parallel |} }. \end{equation}

2.3. Governing equations in the local frame of the characteristics

As we are interested in solutions localised around the characteristic path $\mathcal{C}$ , it is useful to express the governing equations in the local coordinate system associated with this path. We define the local variables $(x_\parallel , x_\perp )$ of a point $\boldsymbol {x}$ close to $\mathcal{C}$ by writing

(2.30) \begin{equation} \boldsymbol{x} = \boldsymbol{x}_0(x_\parallel ) + x_\perp \hat {\boldsymbol{e}}_\perp , \quad \textrm{with} \quad \boldsymbol{x}_0(x_\parallel ) = r_0(x_\parallel ) \hat {\boldsymbol{e}}_r + z_0(x_\parallel ) \hat {\boldsymbol{e}}_z, \end{equation}

where $\boldsymbol {x}_0(x_\parallel )$ is the closest point to $\boldsymbol{x}$ on $\mathcal{C}$ . The differential vector increment can then be expressed as

(2.31) \begin{equation} \begin{aligned} {\textrm d}\boldsymbol{x} &= \displaystyle \frac {\partial \boldsymbol{x}_0}{\partial x_\parallel } {\textrm d}x_\parallel + \displaystyle \frac {\partial \boldsymbol{x}_0}{\partial \phi }{\textrm d}\phi + x_\perp \displaystyle \frac {\partial \hat {\boldsymbol{e}}_{\perp }}{\partial x_\parallel } {\textrm d}x_\parallel + \hat {\boldsymbol{e}}_\perp {\textrm d}x_\perp \\ & = (1 - \kappa x_\perp ) {\textrm d}x_\parallel \hat {\boldsymbol{e}}_\parallel + {\textrm d}x_\perp \hat {\boldsymbol{e}}_{\perp } + r_0 {\textrm d}\phi \, \hat {\boldsymbol{e}}_{\phi } , \end{aligned} \end{equation}

which ensures that $(x_\parallel , x_\perp ,\phi )$ forms an orthogonal curvilinear coordinate system with the metric components $(h_{\parallel },h_\perp ,h_\phi ) = (1 - \kappa (x_\parallel ) x_\perp ,1,r_0(x_\parallel ))$ .

The governing equations (2.2) become, in this local coordinate system

(2.32a) \begin{align} -{\textrm{i}} \varpi u_\parallel - 2 \varOmega (r) {r^{\prime}_0} u_\phi &= -\frac {1}{h_\parallel }\displaystyle \frac {\partial p}{\partial x_\parallel } + b\, \hat {\boldsymbol{e}}_\rho \boldsymbol{\cdot }\hat {\boldsymbol{e}}_\parallel + E({\nabla} ^2 \boldsymbol{u})_\parallel , \end{align}

(2.32b) \begin{align} -{\textrm{i}} \varpi u_\perp + 2 \varOmega (r) {z^{\prime}_0} u_\phi &= -\displaystyle \frac {\partial p}{\partial x_\perp } + b\, \hat {\boldsymbol{e}}_\rho \boldsymbol{\cdot }\hat {\boldsymbol{e}}_\perp + E({\nabla} ^2 \boldsymbol{u})_\perp , \end{align}

(2.32c) \begin{align} -{\textrm{i}} \varpi u_\phi + \omega (r) ({r^{\prime}_0} u_\parallel - {z^{\prime}_0} u_\perp ) &= E ({\nabla} ^2 \boldsymbol{u})_\phi , \end{align}

(2.32d) \begin{align} \displaystyle \frac {\partial (r_0 u_\parallel )}{\partial x_\parallel } + r_0 \displaystyle \frac {\partial (h_\parallel u_\perp )}{\partial x_\perp } & = 0 , \end{align}

(2.32e) \begin{align} -{\textrm{i}} \varpi b + N^2(\rho )( u_\parallel \,\hat {\boldsymbol{e}}_\rho \boldsymbol{\cdot }\hat {\boldsymbol{e}}_\parallel + u_\perp \,\hat {\boldsymbol{e}}_\rho \boldsymbol{\cdot }\hat {\boldsymbol{e}}_\perp ) & = (E/Pr) {\nabla} ^2 b , \end{align}

with the geometric relations:

(2.33) \begin{equation} \hat {\boldsymbol{e}}_\rho \boldsymbol{\cdot }\hat {\boldsymbol{e}}_\parallel = \frac {r_0 r^{\prime}_0 +z_0 {z^{\prime}_0}}{\rho } , \quad \hat {\boldsymbol{e}}_\rho \boldsymbol{\cdot }\hat {\boldsymbol{e}}_\perp = \frac { z_0 {r^{\prime}_0}-r_0 {z^{\prime}_0} + x_\perp }{\rho } , \end{equation}

and the coordinate transformations:

(2.34) \begin{equation} r=r_0 -{z^{\prime}_0} x_\perp , \quad z=z_0 + {r^{\prime}_0} x_\perp , \quad \rho =\sqrt {r^2+z^2} . \end{equation}

3. Viscous solutions localised around a characteristic path

In this section, we develop an approximation for a viscous/diffusive solution that is localised on the characteristic path $\mathcal{C}$ . Our goal is to extend the solutions derived by Moore & Saffman (Reference Moore and Saffman1969) and Thomas & Stevenson (Reference Thomas and Stevenson1972), originally obtained for purely rotating or purely stratified fluid, to the more general case of a non-uniformly rotating and stratified fluid.

Guided by their approach, we introduce a stretched transverse coordinate

(3.1) \begin{equation} \bar {x}_{\perp }= \frac {x_\perp }{E^{1/3}} , \end{equation}

and make the following ansatz:

(3.2a) \begin{align} u_\parallel &= u_\parallel ^{(0)}(x_\parallel , \bar {x}_{\perp }) + \ldots , \end{align}

(3.2b) \begin{align} u_\perp &= E^{1/3} u_\perp ^{(1)}(x_\parallel , \bar {x}_{\perp })+ \ldots , \end{align}

(3.2c) \begin{align} u_\phi & = u_\phi ^{(0)}(x_\parallel , \bar {x}_{\perp }) + \ldots , \end{align}

(3.2d) \begin{align} b &= b^{(0)}(x_\parallel , \bar {x}_{\perp }) + \ldots , \end{align}

(3.2e) \begin{align} p &= E^{1/3} p^{(1)} (x_\parallel , \bar {x}_{\perp }) + \ldots . \end{align}

This ansatz assumes that the solution is mainly oriented along $\boldsymbol {e}_\parallel$ (in the (x,z) plane) and depends on a single fast variable $\tilde {x}_\perp$ , corresponding to a beam width $E^{1/3}$ smaller than the background variation scale.

From (2.32c , e ), the leading-order relations between $u_\parallel ^{(0)}$ , $u_\phi ^{(0)}$ and $b^{(0)}$ are

(3.3a) \begin{align} u_\phi ^{(0)} &= -\frac { {\textrm{i}} \omega _0 {r^{\prime}_0} }{\varpi } u_{\parallel }^{(0)} , \end{align}

(3.3b) \begin{align} b^{(0)} &= -\frac {{\textrm{i}} N^2_0}{\varpi } \frac {r_0{r^{\prime}_0} + z_0 {z^{\prime}_0}}{\rho _0} u_\parallel ^{(0)} , \end{align}

with $\rho _0=\sqrt {r_0^2+z_0^2}$ , where we recall $N_0=N(\rho _0)$ and $\omega _0=\omega (r_0)$ . From the leading-order balance of (2.32b ), we obtain

(3.4) \begin{align} \displaystyle \frac {\partial p^{(1)}}{\partial \bar {x}_{\perp }} = -2 \varOmega _0 z^{\prime}_0 u_\phi ^{(0)} - \frac {r_0z^{\prime}_0 - z_0r^{\prime}_0}{\rho _0} b^{(0)}. \end{align}

Substituting expressions from (3.3a , b ) yields

(3.5) \begin{equation} \displaystyle \frac {\partial p^{(1)}}{\partial \bar {x}_{\perp }} = \frac {{\textrm{i}}}{\varpi }\left ( \xi ^2_0{r^{\prime}_0} z^{\prime}_0 + N^2_0 \frac {(r_0{r^{\prime}_0} + z_0 {z^{\prime}_0})(r_0z^{\prime}_0 - z_0r^{\prime}_0)}{\rho _0^2} \right )u_\parallel ^{(0)} . \end{equation}

Now consider the operation ${\textrm{i}} \varpi$ (2.32a ) $- 2 \varOmega {r^{\prime}_0}$ (2.32c ) − $\hat {\boldsymbol{e}}_\rho \boldsymbol{\cdot }\hat {\boldsymbol{e}}_\parallel$ (2.32e ), which gives

(3.6) \begin{equation} \begin{array}{l} \left ( \varpi ^2 - \xi ^2 {r^{\prime 2}_0} -N^2 \displaystyle \frac {(r_0 r^{\prime}_0 + z_0 z^{\prime}_0)^2}{\rho ^2} \right ) \big(u_\parallel ^{(0)} + E^{1/3} u_\parallel ^{(1)}\big)\big(1- \kappa _0 \bar {x}_{\perp } E^{1/3}\big) \\ \hspace {1cm} = - E^{1/3} \left (\xi ^2_0r^{\prime}_0z^{\prime}_0 + N^2_0 \displaystyle \frac {(r_0 r^{\prime}_0 + z_0 z^{\prime}_0)(r_0{z^{\prime}_0}-z_0{r^{\prime}_0})}{\rho _0^2} \right ) u_\perp ^{(1)} - {\textrm{i}} E^{1/3} \varpi \displaystyle \frac {\partial p^{(1)}}{\partial x_\parallel } \\ \hspace {1.5cm} + E^{1/3} \left ( {\textrm{i}} \varpi \displaystyle \frac {\partial ^2u_\parallel ^{(0)}}{\partial \bar {x}_{\perp }^2} - 2\varOmega _0r^{\prime}_0 \displaystyle \frac {\partial ^2u_\phi ^{(0)}}{\partial \bar {x}_{\perp }^2} - \frac {1}{Pr}\frac {(r_0 r^{\prime}_0 + z_0 z^{\prime}_0)}{\rho _0} \displaystyle \frac {\partial ^2b^{(0)}}{\partial \bar {x}_{\perp }^2} \right ) +O(E^{2/3}) . \end{array} \end{equation}

Due to the dispersion relation (2.18), the left-hand side reduces to

(3.7) \begin{align} E^{1/3} \left [ \partial _r(\xi ^2)_{_{\! 0}} {r^{\prime 2}_0}{z^{\prime}_0} + \partial _\rho (N^2/\rho ^2)_{_{\! 0}} (r_0 r^{\prime}_0 + z_0 z^{\prime}_0)^2(r_0{z^{\prime}_0}-z_0{r^{\prime}_0}) \right ] \bar {x}_{\perp } u_\parallel ^{(0)} . \end{align}

Taking the derivative with respect to $\bar {x}_{\perp }$ and substituting $\partial _{\bar {x}_{\perp }} p^{(1)}$ from (3.5), $u_\phi ^{(0)}$ and $b^{(0)}$ from (3.3a , b ), and using the expression for $\partial _{\bar {x}_{\perp }}u_\perp ^{(1)}$ from (2.32d )

(3.8) \begin{equation} \displaystyle \frac {\partial u_\perp ^{(1)}}{\partial \bar {x}_{\perp }} = -\frac {1}{r_0} \displaystyle \frac {\partial \big(r_0 u_\parallel ^{(0)}\big)}{\partial x_\parallel }, \end{equation}

we finally obtain a single equation for $u_\parallel ^{(0)}$

(3.9) \begin{equation} \begin{array}{l} \left (\xi ^2_0r^{\prime}_0z^{\prime}_0 + N^2_0 \displaystyle \frac {(r_0 r^{\prime}_0 + z_0 z^{\prime}_0)(r_0{z^{\prime}_0}-z_0{r^{\prime}_0})}{\rho _0^2} \right ) \displaystyle \frac {1}{r_0} \displaystyle \frac {\partial (r_0 u_\parallel ^{(0)})}{\partial x_\parallel } = \\ \hspace {2cm} -\displaystyle \frac {\partial }{\partial x_\parallel }\left [ \left ( \xi ^2_0{r^{\prime}_0} z^{\prime}_0 + N^2_0 \frac {(r_0{r^{\prime}_0} + z_0 {z^{\prime}_0})(r_0z^{\prime}_0 - z_0r^{\prime}_0)}{\rho _0^2} \right )u_\parallel ^{(0)} \right ] \\ \hspace {2cm}- \displaystyle \frac {{\textrm{i}}}{\varpi }\left (2\xi ^2_0{r^{\prime 2}_0} + \left (1+\displaystyle \frac {1}{Pr} \right ) N^2_0 \displaystyle \frac {(r_0 r^{\prime}_0 + z_0 z^{\prime}_0)^2}{\rho _0^2} \right ) \displaystyle \frac {\partial ^3u_\parallel ^{(0)}}{\partial \bar {x}_{\perp }^3} \\ \hspace {1cm} + \left [\partial _r(\xi ^2)_{_{\! 0}} {r^{\prime 2}_0}{z^{\prime}_0} + \partial _\rho (N^2/\rho ^2)_{_{\! 0}} (r_0 r^{\prime}_0 + z_0 z^{\prime}_0)^2(r_0{z^{\prime}_0}-z_0{r^{\prime}_0}) \right ] \displaystyle \frac {\partial \big(\bar {x}_{\perp } u_\parallel ^{(0)}\big)}{\partial \bar {x}_{\perp }}. \end{array} \end{equation}

If we now introduce the new function $\tilde {u}_\parallel$ and the new variables $\tilde {x}_\parallel$ and $\tilde {x}_\perp$ defined by

(3.10a) \begin{align} u_\parallel ^{(0)}(\bar {x}_{\perp }, x_\parallel ) &= \frac {F(x_\parallel )}{\sqrt {r_0| K(x_\parallel )|}} \tilde {u}_\parallel (\tilde {x}_\perp ,\tilde {x}_\parallel ) , \end{align}

(3.10b) \begin{align} \tilde {x}_\parallel &= \int _{x_\parallel ^{(0)}}^{x_\parallel } \displaystyle \frac {D F^3}{2 K} {\textrm d}x_\parallel +\tilde {x}_\parallel ^{(0)}, \end{align}

(3.10c) \begin{align} \tilde {x} _\perp &= F(x_\parallel ) \bar {x}_{\perp } , \end{align}

with

(3.11a) \begin{align} K(x_\parallel ) &= \xi ^2_0r^{\prime}_0z^{\prime}_0 + N^2_0 \displaystyle \frac {(r_0 r^{\prime}_0 + z_0 z^{\prime}_0)(r_0{z^{\prime}_0}-z_0{r^{\prime}_0})}{\rho _0^2} , \end{align}

(3.11b) \begin{align} H(x_\parallel ) &= \partial _r(\xi ^2)_{_{\! 0}} {r^{\prime 2}_0}{z^{\prime}_0} + \partial _\rho (N^2/\rho ^2)_{_{\! 0}} (r_0 r^{\prime}_0 + z_0 z^{\prime}_0)^2(r_0{z^{\prime}_0}-z_0{r^{\prime}_0}) ,\\[-12pt]\nonumber \end{align}

(3.11c) \begin{align} D(x_\parallel ) &= \displaystyle \frac {2\xi ^2_0{r^{\prime 2}_0}\rho _0^2 + (1+P{\kern-1pt}r^{-1}) N^2_0 (r_0 r^{\prime}_0 + z_0 z^{\prime}_0)^2}{ \varpi \rho _0^2}, \end{align}

(3.11d) \begin{align} F(x_\parallel ) &= \exp \left (\int _{x_\parallel ^{(0)}}^{x_\parallel } \frac {H}{2K}{\textrm d}x_\parallel \right ) , \end{align}

then (3.9) becomes the parameter-free equation

(3.12) \begin{equation} \displaystyle \frac {\partial \tilde {u}_\parallel }{\partial \tilde {x}_{\parallel }} = -{\textrm{i}} \displaystyle \frac {\partial ^3\tilde {u}_\parallel }{\partial \tilde {x}_{\perp }^3} . \end{equation}

In (3.10b ) and (3.11d ), the parameter $x_\parallel ^{(0)}\gt 0$ corresponds to the arclength at which we start considering the solution, and $\tilde {x}_\parallel ^{(0)}$ is an arbitrary constant. The functions $K$ , $H$ and $D$ can also be expressed in terms of the angles $\theta = \textrm {atan}(z_0/r_0)$ and $\phi =\textrm {atan}({z^{\prime}_0}/{r^{\prime}_0})$ as

(3.13a) \begin{align} K &= \xi ^2_0\cos \phi \sin \phi + N^2_0 \cos (\phi -\theta )\sin (\phi -\theta ), \end{align}

(3.13b) \begin{align} H & = \partial _r(\xi ^2)_{_{\! 0}} \cos ^2\phi \sin \phi + \rho _0^2\partial _\rho (N^2/\rho ^2)_{_{\! 0}}\cos ^2(\phi -\theta )\sin (\phi -\theta ), \end{align}

(3.13c) \begin{align} D & = \displaystyle \frac {2\xi ^2_0\cos ^2\phi + (1+P{\kern-1pt}r^{-1}) N^2_0 \cos ^2(\phi -\theta )}{ \varpi }. \end{align}

The function $H$ contains the terms associated with the spatial variations of the background rotation and stratification. It can also be written as

(3.14) \begin{equation} H= -\left .\displaystyle \frac {\partial \varpi _{nv}^2}{\partial x_\perp }\right |_0 , \end{equation}

where

(3.15) \begin{align} \varpi _{nv}^2(r,z,\phi ) = \xi ^2(r)\cos ^2\phi + N^2(\rho )(r\cos \phi + z \sin \phi )^2/\rho ^2 \end{align}

is the local inviscid dispersion relation, and $r$ and $z$ depend on $x_\perp$ according to (2.34).

Equation (3.12) is the same as the one obtained by Moore & Saffman (Reference Moore and Saffman1969) and Thomas & Stevenson (Reference Thomas and Stevenson1972). It can be easily solved using a Fourier transform in $\bar {x}_{\perp }$ , yielding a general solution

(3.16) \begin{equation} \tilde {u}_\parallel (\tilde {x}_{\perp },\tilde {x}_\parallel ) = \int _{-\infty }^{\infty } a(k)\textrm{e}^{-k^3 \tilde {x}_{\parallel }}\textrm{e}^{\textrm{i} k \tilde {x}_{\perp }} {\textrm d}k . \end{equation}

A localised solution was also obtained by Kistovich & Chashechkin (Reference Kistovich and Chashechkin1998) for a two-dimensional vertically stratified fluid. They used a global WKB ansatz, taking advantage of the separability of the problem in that case. If we apply our analysis to this configuration for which $N=N(z)$ and $\xi =0$ , we obtain

(3.17a) \begin{align} \varpi ^2 & = N_0^2 {z_0}^{\prime 2} ,\quad K= N_0^2 {z^{\prime}_0}{r^{\prime}_0}= \varpi ^2 {r^{\prime}_0}/{z^{\prime}_0},\quad H= \partial _z (N^2)_{_{\! 0}} {r^{\prime}_0}{z_0}^{\prime 2}, \end{align}

(3.17b) \begin{align} D & = (1+P{\kern-1pt}r^{-1}) \varpi ,\quad F= N_0= \frac {\varpi }{|{z^{\prime}_0}|} , \end{align}

which gives

(3.18a) \begin{align} u_\parallel ^{(0)} &= |{z^{\prime}_0}{r^{\prime}_0}|^{-1/2} \tilde {u}_\parallel (\tilde {x}_\perp ,\tilde {x}_\parallel ) ,\quad \end{align}

(3.18b) \begin{align} \tilde {x} _\perp &= \frac {\varpi }{|{z^{\prime}_0}|} \bar {x}_{\perp } , \quad \tilde {x}_\parallel = \displaystyle \frac { (1+P{\kern-1pt}r^{-1}) \varpi ^2 }{2}\int _{x_\parallel ^{(0)}}^{x_\parallel } \displaystyle \frac { {\textrm d}x_\parallel }{ r^{\prime}_0 {z_0}^{\prime 2}} +\tilde {x}_\parallel ^{(0)} . \end{align}

These expressions are identical to (2.5) and (2.6) of Kistovich & Chashechkin (Reference Kistovich and Chashechkin1998), once the solution is expressed in terms of $\bar {x}_{\perp }$ .

In the denominator of (3.10b ), we recognise the expression (2.21) for $\epsilon _1\sqrt {\varDelta }$ . For $x_\parallel \gt x_\parallel ^{(0)}$ , the sign of $\tilde {x}_\parallel$ depends on the value of $\epsilon _1$ . The solutions propagating along $\mathcal{C}$ in the direction of increasing $x_\parallel$ consist of positive wavenumbers ( $a(k\lt 0)=0$ ) if $\epsilon _1=1$ , and negative wavenumbers ( $a(k\gt 0)=0$ ) if $\epsilon _1=-1$ . Hence, the solutions propagating along $\mathcal{C}$ in the direction of increasing $x_\parallel$ can be written as

(3.19) \begin{equation} \tilde {u}_\parallel (\tilde {x}_{\perp },\tilde {x}_\parallel ) = \int _{0}^{\infty } b(k)\textrm{e}^{-k^3 \epsilon _1 \tilde {x}_{\parallel }}\textrm{e}^{\textrm{i} k \epsilon _1 \tilde {x}_{\perp }} {\textrm d}k . \end{equation}

When the incident beam reaches the turning point $x_{\parallel c}$ , the viscous solution given by (3.10a – c ) and (3.11a – d ) breaks down, as the amplitude prefactor in the expression (3.10a ) for $u_\parallel ^{(0)}$ diverges (this term is proportional to $\Delta ^{-1/4}$ , see expression (2.21)). While the denominators in $\tilde {x} _\parallel$ and $F$ in (3.10b ) and (3.11d ) vanish at this point, the integrals remain convergent. The behaviour near the turning point can be obtained by taking $x_\parallel \rightarrow x_{\parallel c}$ in (3.10a – c ). We obtain

(3.20) \begin{equation} u_\parallel ^{(0)} \sim \frac { F_c }{\sqrt {r_c} \left | -\Delta ^{\prime}_c(x_{\parallel c} - x_\parallel ) \right |^{1/4}} \tilde {u}_\parallel (F_c\bar {x}_{\perp }, \tilde {x}_{\parallel c}) . \end{equation}

This expression can be written in terms of the local variable $(y_\parallel , y_\perp )$ when $x_\parallel \rightarrow x_{\parallel c}$ using the approximations

(3.21) \begin{equation} x_{\parallel c} - x_\parallel \sim - y_{\parallel 0} = -y_\parallel , \quad x_\perp \sim y_\perp - y_{\perp 0} \sim y_\perp - \frac {2}{3}\alpha _c (-y_\parallel )^{3/2} . \end{equation}

Since $x_\perp =O(E^{1/3})$ , the second relation implies that $y_\perp = O(E^{1/3})$ and $y_\parallel = O(E^{2/9})$ in the vicinity of the turning point. We will use this scaling to analyse the behaviour of the solution near the turning-point region in the next section.

Note that, in expression (3.19), if we choose

(3.22) \begin{equation} b(k) =b^S(k) \equiv k^{m-1} \frac {\textrm{e}^{-\textrm{i} m \pi /2}}{(m-1)!} , \quad \textrm{with} \quad m\gt 0, \end{equation}

we recover the similarity solution derived by Moore & Saffman (Reference Moore and Saffman1969) and Thomas & Stevenson (Reference Thomas and Stevenson1972)

(3.23) \begin{equation} \tilde {u}_\parallel = \tilde {u}_\parallel ^{S} \equiv |\tilde {x}_\parallel |^{-m/3} h_m(\eta ), \quad \eta = \epsilon _1 \frac {\tilde {x}_\perp }{|\tilde {x}_\parallel |^{1/3}} = \epsilon _1 \frac {\bar {x}_{\perp }}{|\tilde {x}_\parallel /F^3|^{1/3}} , \end{equation}

where the function $h_m(\eta )$ is defined as

(3.24) \begin{equation} h_m(\eta ) = \frac {\textrm{e}^{-\textrm{i} m \pi /2}}{(m-1)!} \int _0^{\infty } \textrm{e}^{-p^3} \textrm{e}^{\textrm{i} p \eta } p^{m-1} {\textrm d}p. \end{equation}

The normalisation of the function $h_m$ ensures the following asymptotic behaviour:

(3.25a) \begin{align} h_m(\eta ) & \begin{array}[t]{c} \sim \\ \scriptstyle {\eta \rightarrow +\infty } \end{array} |\eta |^{-m} , \end{align}

(3.25b) \begin{align} h_m(\eta ) & \begin{array}[t]{c} \sim \\ \scriptstyle {\eta \rightarrow -\infty } \end{array} \textrm{e}^{-\textrm{i} m \pi }|\eta |^{-m} . \end{align}

This implies that the similarity solution $\tilde {u}_\parallel ^{S}$ behaves as

(3.26a) \begin{align} & \tilde {u}_\parallel ^S \begin{array}[t]{c} \sim \\ \scriptstyle {\epsilon _1\tilde {x}_{\perp } \rightarrow +\infty } \end{array} | \tilde {x}_{\perp } |^{-m} , \end{align}

(3.26b) \begin{align} & \tilde {u}_\parallel ^S \begin{array}[t]{c} \sim \\ \scriptstyle {\epsilon _1\tilde {x}_{\perp } \rightarrow -\infty } \end{array} \textrm{e}^{-\textrm{i} m \pi }|\tilde {x}_{\perp }|^{-m} . \end{align}

These behaviours demonstrate that expression (3.23) defines a viscous solution associated with an inviscid singularity of the form $|x_\perp |^{-m}$ along the characteristic path $\mathcal{C}$ . The definition (3.23) of the similarity variable $\eta$ shows that this singularity originates at $\tilde {x}_\parallel =0$ . The width of the wave beam is given by $|\tilde {x}_\parallel |^{1/3} /F$ , where $\tilde {x}_\parallel$ and $F$ evolve according to (3.10b ) and (3.11d ), respectively.

4. Solution close to the turning point and reflected beam

We have seen in the previous section that the solution described by (3.10a , b ) and (3.19) breaks down upon reaching a turning point. In this section, our objective is to derive a new approximation that remains valid in the vicinity of a turning point. To this end, we adopt the local frame ( $\hat {\boldsymbol{e}}_{\parallel c}$ , $\hat {\boldsymbol{e}}_{\perp c}$ ) and the local coordinates ( $y_\parallel$ , $y_\perp$ ).

In this frame, if we denote the velocity components along $\hat {\boldsymbol{e}}_{\parallel c}$ and $\hat {\boldsymbol{e}}_{\perp c}$ by $v_\parallel$ and $v_\perp$ , respectively, the governing equations become

(4.1a) \begin{align} -{\textrm{i}} \varpi v_\parallel - 2 \varOmega r^{\prime}_c u_\phi &= -\displaystyle \frac {\partial p}{\partial y_\parallel } + b\, \hat {\boldsymbol{e}}_\rho \boldsymbol{\cdot }\hat {\boldsymbol{e}}_{\parallel c} + E{\nabla} ^2 v_\parallel , \end{align}

(4.1b) \begin{align} -{\textrm{i}} \varpi v_\perp + 2 \varOmega z^{\prime}_c u_\phi &= -\displaystyle \frac {\partial p}{\partial y_\perp } + b\, \hat {\boldsymbol{e}}_\rho \boldsymbol{\cdot }\hat {\boldsymbol{e}}_{\perp c} + E{\nabla} ^2 v_\perp , \end{align}

(4.1c) \begin{align} -{\textrm{i}} \varpi u_\phi + \omega \big(r^{\prime}_c v_\parallel - z^{\prime}_c v_\perp \big) &= E {\nabla} ^2 u_\phi , \end{align}

(4.1d) \begin{align} \displaystyle \frac {\partial rv_\parallel }{\partial y_\parallel } + r \displaystyle \frac {\partial v_\perp }{\partial y_\perp } &= 0 , \end{align}

(4.1e) \begin{align} -{\textrm{i}} \varpi b + N^2( v_\parallel \,\hat {\boldsymbol{e}}_\rho \boldsymbol{\cdot }\hat {\boldsymbol{e}}_{\parallel c} + v_\perp \,\hat {\boldsymbol{e}}_\rho \boldsymbol{\cdot }\hat {\boldsymbol{e}}_{\perp c} ) &= (E/Pr) {\nabla} ^2 b . \end{align}

We now introduce the following ansatz:

(4.2) \begin{equation} v_\parallel = v_\parallel ^{(0)} , \quad v_\perp = E^{1/9} v_\perp ^{(1)} , \quad u_\phi = u_\phi ^{(0)} , \quad b = b^{(0)}, \, p= E^{4/9} p^{(4)}, \end{equation}

together with the rescaled local variables

(4.3) \begin{equation} \bar {y}_\parallel = E^{-2/9} y_\parallel , \quad \bar {y}_\perp = E^{-1/3} y_\perp . \end{equation}

Under this scaling, the governing equations reduce to

(4.4a) \begin{align} & -{\textrm{i}} \varpi v_\parallel ^{(0)} - 2 \varOmega r^{\prime}_c u_\phi ^{(0)} = - E^{2/9} \displaystyle \frac {\partial p^{(4)}}{\partial \bar {y}_\parallel } + b^{(0)}( r r^{\prime}_c+z z^{\prime}_c)/\rho ,\quad \end{align}

(4.4b) \begin{align} & -{\textrm{i}} \varpi E^{1/9} v_\perp ^{(1)} + 2 \varOmega z^{\prime}_c u_\phi ^{(0)} = - E^{1/9}\displaystyle \frac {\partial p^{(4)}}{\partial \bar {y}_\perp } + b^{(0)}(zr^{\prime}_c -rz^{\prime}_c)/\rho ,\quad \end{align}

(4.4c) \begin{align} & -{\textrm{i}} \varpi u_\phi ^{(0)}+ \omega \big(r^{\prime}_c v_\parallel ^{(0)} - z^{\prime}_c v_\perp ^{(1)} E^{1/9}\big) = 0, \end{align}

(4.4d) \begin{align} & \displaystyle \frac {\partial v_\parallel ^{(0)}}{\partial \bar {y}_\parallel } + \displaystyle \frac {\partial v_\perp ^{(1)}}{\partial \bar {y}_\perp } =0 ,\quad \end{align}

(4.4e) \begin{align} & -{\textrm{i}} \varpi b^{(0)} + (N^2/\rho )\big( v_\parallel ^{(0)} ( rr^{\prime}_c+zz^{\prime}_c) + E^{1/9} v_\perp ^{(1)} (zr^{\prime}_c -rz^{\prime}_c) \big) = 0. \end{align}

Note that viscous terms do not appear in these equations.

From (4.4c , e ), we express $u_\phi ^{(0)}$ and $b^{(0)}$ in terms of $v_\parallel ^{(0)}$ and $v_\perp ^{(1)}$

(4.5a) \begin{align} u_\phi ^{(0)} & = \frac {\omega }{{\textrm{i}} \varpi }\big(r^{\prime}_c v_\parallel ^{(0)} - z^{\prime}_c E^{1/9} v_\perp ^{(1)} \big) , \end{align}

(4.5b) \begin{align} b^{(0)} & = \frac { N^2}{{\textrm{i}} \varpi \rho }\big( v_\parallel ^{(0)} ( rr^{\prime}_c+zz^{\prime}_c) - E^{1/9} v_\perp ^{(1)} (rz^{\prime}_c -zr^{\prime}_c) \big) . \end{align}

Substituting these into (4.4a , b ) yields

(4.6a) \begin{align} -\frac {\varDelta ^{\prime}_c}{{\textrm{i}}\varpi (2\varpi ^2 -\xi _c^2 -N^2_c )} \bar {y}_\parallel v_\parallel ^{(0)} & = - \displaystyle \frac {\partial p^{(4)}}{\partial \bar {y}_\parallel } , \end{align}

(4.6b) \begin{align} \frac {2\varpi ^2 -\xi _c^2 -N^2_c }{{\textrm{i}} \varpi } v_\perp ^{(1)} & = -\displaystyle \frac {\partial p^{(4)}}{\partial \bar {y}_\perp } , \\[9pt] \nonumber \end{align}

where we have used the following estimates, deduced from (2.18), (2.19) and (2.20) near the turning point $r_c$ :

(4.7) \begin{align} & 2\varOmega \omega r^{\prime}_cz^{\prime}_c + N^2 (rr^{\prime}_c+zz_c)(rz^{\prime}_c-zr^{\prime}_c)/\rho ^2 = O(E^{2/9}), \end{align}

(4.8) \begin{align} & - \varpi ^2 + 2\varOmega \omega z_c^{\prime 2}+ N^2(rz^{\prime}_c-zr^{\prime}_c)^2/\rho ^2 \sim - 2\varpi ^2 + \xi _c^2 + N_c^2, \end{align}

(4.9) \begin{align} & -\varpi ^2 + 2\varOmega \omega r_c^{\prime 2}+ N^2(rr^{\prime}_c+zz^{\prime}_c)^2/\rho ^2 \sim E^{2/9} \frac {\varDelta ^{\prime}_c}{2\varpi ^2 - \xi _c^2 -N^2_c} \bar {y}_\parallel . \end{align}

Eliminating $p^{(4)}$ and $v_\perp ^{(1)}$ using (4.4d ) finally gives a single equation for $v_\parallel ^{(0)}$

(4.10) \begin{equation} \displaystyle \frac {\partial ^2v_\parallel ^{(0)}}{\partial \bar {y}_\parallel ^2} - \frac {\varDelta ^{\prime}_c\, \bar {y}_\parallel }{ (2\varpi ^2 - \xi _c^2 -N^2_c )^2} \displaystyle \frac {\partial ^2v_\parallel ^{(0)}}{\partial \bar {y}_\perp ^2} =0 . \end{equation}

This equation can be solved using a Fourier transform in the $\bar {y}_\perp$ direction

(4.11) \begin{equation} v_\parallel ^{(0)} = \int _{-\infty }^{+\infty } \tilde {v}_\parallel \textrm{e}^{\textrm{i} k \bar {y}_\perp }{\textrm d}k, \end{equation}

such that $\tilde {v}_{\parallel }$ satisfies the Airy equation

(4.12) \begin{equation} \displaystyle \frac {\partial ^2\tilde {v}_{\parallel }}{\partial \tilde {y}_\parallel ^2} - \tilde {y}_\parallel \tilde {v}_\parallel =0, \end{equation}

with

(4.13) \begin{equation} \tilde {y}_{\parallel } = \gamma |k|^{2/3} \bar {y}_\parallel , \end{equation}

and

(4.14) \begin{equation} \gamma = \left (\frac {-\varDelta ^{\prime}_c}{(2\varpi ^2 -\xi _c^2 -N^2_c )^2}\right )^{1/3} . \end{equation}

Finding an Airy equation near a turning point is not a surprise (Wasow Reference Wasow1985). Such an equation was derived in the context of non-rotating fluids by Sutherland (Reference Sutherland2010) for plane waves and by Kistovich & Chashechkin (Reference Kistovich and Chashechkin1998) for concentrated beams.

The solution to the Airy (4.12) that remains bounded for large $|\tilde {y}_\parallel |$ is

(4.15) \begin{equation} \tilde {v}_\parallel = c(k) \textrm {Ai}\left (\tilde {y}_\parallel \right ), \end{equation}

where $\textrm {Ai}$ is an Airy function (Abramowitz & Stegun Reference Abramowitz and Stegun1965, p. 446), and $c(k)$ is a function to be determined by matching.

Using the asymptotic expansion of the Airy function as $\tilde {y}_\parallel \rightarrow -\infty$ (Abramowitz & Stegun Reference Abramowitz and Stegun1965, p. 448), we obtain

(4.16) \begin{equation} \tilde {v}_\parallel \sim \frac {c(k)}{\sqrt {\pi } (-\tilde {y}_\parallel )^{1/4}} \sin \left (\frac {2}{3}(-\tilde {y}_\parallel )^{3/2} +\frac {\pi }{4}\right ). \end{equation}

This leads to the following asymptotic form for $v_\parallel ^{(0)}$ :

(4.17) \begin{equation} \begin{array}{l} \displaystyle v_\parallel ^{(0)} \sim \int _{-\infty }^{+\infty } \displaystyle \frac {C(k)}{(-\bar {y}_\parallel )^{1/4}}\left \{\exp \left [{\textrm{i}} k \left (\bar {y}_\perp + \frac {2 |k|}{3k}\gamma ^{3/2} (-\bar {y}_\parallel )^{3/2} \right ) -{\textrm{i}} \pi /4\right ] \right . \, \\[12pt] \hspace {0.8cm} \left .+ \exp \left [{\textrm{i}} k \left (\bar {y}_\perp - \displaystyle \frac {2 |k|}{3k}\gamma ^{3/2} (-\bar {y}_\parallel )^{3/2} \right ) +{\textrm{i}} \pi /4\right ] \right \} {\textrm d}k, \end{array} \end{equation}

with

(4.18) \begin{equation} C(k)= \frac {c(k)}{2\sqrt {\pi }\gamma ^{1/4}|k|^{1/6} } . \end{equation}

Each wave beam corresponds to a prescribed sign of the axial wavenumber $k$ . Expression (4.17) is therefore a sum of four wave beams, corresponding to positive and negative wavenumbers propagating along two distinct paths: $\bar {y}_\perp = ({2}/{3})\gamma ^{3/2}(-\bar {y}_\parallel )^{3/2}$ and $\bar {y}_\perp = - ({2}/{3})\gamma ^{3/2}(-\bar {y}_\parallel )^{3/2}$ . Only one of these corresponds to the incident beam. Using (3.20) and (3.21), the expression for the incident beam in terms of the local variable is

(4.19) \begin{equation} u_\parallel ^{(0)} \sim \frac { \displaystyle \int _0^{+\infty } b(k /F_c) \textrm{e}^{-k^3\epsilon _1\tilde {x}_{\parallel c}/F_c^3 } \textrm{e}^{\textrm{i} k \epsilon _1 ( \bar {y}_\perp - \frac {2}{3}\alpha _c (-\bar {y}_\parallel )^{3/2})} {\textrm d}k}{r_c^{1/2}(-\varDelta ^{\prime}_c)^{1/4}(-\bar {y}_\parallel )^{1/4}E^{1/18}}. \end{equation}

From (2.26), we can write

(4.20) \begin{equation} \alpha _c = - \epsilon _1 \epsilon _2 \gamma ^{3/2} , \end{equation}

with $\epsilon _2 = \textrm {sgn}(2\varpi ^2 - \xi _c^2 - N_c^2)$ . We then immediately see that the incident beam corresponds to the second term in (4.17) if $\epsilon _2 \gt 0$ , but to the first term if $\epsilon _2 \lt 0$ . The condition of matching then imposes the following relation between $C(k)$ in (4.17) and $b(k)$ in (4.19):

(4.21) \begin{equation} C(k) = \frac {H(\epsilon _1 k) b(\epsilon _1 k /F_c) }{r_c^{1/2}(-\varDelta ^{\prime}_c)^{1/4} E^{1/18}} \textrm{e}^{-k^3 \tilde {x}_{\parallel c} /F_c^3 } \textrm{e}^{\textrm{i} \epsilon _2 \pi /4} , \end{equation}

where $H(z)$ is the Heaviside function.

Expression (4.21) for $C(k)$ implies the following expression for $c(k)$ appearing in (4.15):

(4.22) \begin{equation} c(k) = \frac {H(\epsilon _1 k) b(\epsilon _1 k /F_c) 2 \sqrt {\pi }\gamma ^{1/4} |k|^{1/6} }{r_c^{1/2}(-\varDelta ^{\prime}_c)^{1/4} E^{1/18}} \textrm{e}^{-k^3 \tilde {x}_{\parallel c} /F_c^3} \textrm{e}^{\textrm{i} \epsilon _2 \pi /4} . \end{equation}

For the similarity solution obtained with $b(k)$ given by (3.22), we obtain the following expression for the solution in the turning-point region:

(4.23) \begin{equation} v_\parallel = C_t V_m^{S}(Y_\parallel ,Y_\perp ) , \end{equation}

where

(4.24) \begin{equation} V_m^{S}(Y_\parallel ,Y_\perp ) \equiv \int _0^{\infty } k^{m-1+1/6} \textrm {Ai}(k^{2/3}Y_\parallel ) \textrm{e}^{\textrm{i} k Y_\perp } \textrm{e}^{-k^3} {\textrm d}k, \end{equation}

and

(4.25a) \begin{align} C_t & = \frac { 2 \sqrt {\pi }\gamma ^{1/4} F_c^{7/6} }{\tilde {x}_{\parallel c}^{(m+1/6)/3} r_c^{1/2}(-\varDelta ^{\prime}_c)^{1/4} E^{1/18}} \frac {\textrm{e}^{-\textrm{i} m \pi /2}}{(m-1)!} \textrm{e}^{\textrm{i} \epsilon _2 \pi /4} \, , \end{align}

(4.25b) \begin{align} Y_\parallel & = \frac {\gamma F_c^{2/3} y_\parallel }{|\tilde {x}_{\parallel c}|^{2/9} E^{2/9}} \, , \end{align}

(4.25c) \begin{align} Y_\perp & = \frac {\epsilon _1F_c y_\perp }{|\tilde {x}_{\parallel c}|^{1/3} E^{1/3}} \, . \end{align}

We recall that $\gamma$ is defined in (4.14), $\varDelta$ in (2.11), $\tilde {x}_\parallel$ in (3.10b ) and $F$ in (3.11d ).

Contour plots of the real part, imaginary part, norm and phase of the function $V_1^S$ in the $(Y_\parallel ,Y_\perp )$ plane are shown in figure 2. The real part $\textrm{Re}\, e(V_m^S)$ is even with respect to the variable $Y_\perp$ while the imaginary part $\textrm{Im}\,m(V_m^S)$ is odd in $Y_\perp$ . As time evolves, the velocity field is expected to oscillate between these two components. As seen in figure 2(c), $|V_m^S|$ is localised around the ray trajectory defined by $Y_\perp =\pm (-Y_\parallel )^{3/2}$ for $Y_\parallel \leqslant 0$ . For positive $Y_\parallel$ , $|V_m^{S}|$ becomes evanescent, i.e. it decays exponentially as $Y_\parallel$ increases, consistent with the known behaviour of the Airy function Ai for positive arguments. This agrees with the physical expectation that no waves propagate beyond the turning point located at the origin. In figure 2(d), we can observe how the phase of the beam changes as it passes through the turning point.

The other term in (4.17) corresponds to the reflected beam. It has the same form as the incident beam expression (4.19), but it is now localised along the path $\bar {y} = -(2/3)\alpha _c(-\bar {y}_\parallel )^{3/2}$ . If we express this term in the coordinates associated with the reflected beam, which is such that $\hat {\boldsymbol{e}} _\parallel$ and $\hat {\boldsymbol{e}}_\perp$ are reversed compared with those of the incident beam (see figure 1), we obtain

(4.26) \begin{equation} u_{\parallel R}^{(0)} \sim - \frac { \textrm{e}^{\textrm{i} \epsilon _2 \pi /2} \displaystyle \int _0^{+\infty } b(k /F_c ) \textrm{e}^{-k^3\epsilon _1\tilde {x}_{\parallel c} /F_c^3 } \textrm{e}^{\textrm{i} k \epsilon _1 ( \bar {y}_\perp + \tfrac {2}{3}\alpha _c (-\bar {y}_\parallel )^{3/2})} {\textrm d}k}{r_c^{1/2}(-\varDelta ^{\prime}_c)^{1/4}(-\bar {y}_\parallel )^{1/4}E^{1/18}} . \end{equation}

This yields an expression for the reflected beam of the same form as (3.10) as we move away from the turning point

(4.27a) \begin{align} u_{\parallel R}^{(0)}(\bar {x}_{\perp }, x_\parallel )& = \frac {F(x_\parallel )}{\sqrt {r_0| K(x_\parallel )|}} \tilde {u}_{\parallel R} (\tilde {x}_\perp ,\tilde {x}_\parallel ) ,\quad \end{align}

(4.27b) \begin{align} \tilde {x}_\parallel & = \int _{x_{\parallel c}}^{x_\parallel } \displaystyle \frac {D F^3}{2 K} {\textrm d}x_\parallel +\tilde {x}_{\parallel c}, \end{align}

(4.27c) \begin{align} \tilde {x} _\perp & = F(x_\parallel ) \bar {x}_{\perp } , \end{align}

(4.27d) \begin{align} F(x_\parallel ) & = \exp \left ( \int _{x_{\parallel c}}^{x_\parallel } \frac {H}{2K}{\textrm d}x_\parallel \right ), \end{align}

where $K$ , $H$ and $D$ are defined in (3.11a – c ) and $\tilde {u}_{\parallel R} (\tilde {x}_{\perp },\tilde {x}_\parallel )$ is given by

(4.28) \begin{equation} \tilde {u}_{\parallel R} = \textrm{e}^{-\textrm{i} \epsilon _2 \pi /2} \int _{0}^{\infty } b(k)\textrm{e}^{-k^3 \epsilon _1 \tilde {x}_{\parallel }}\textrm{e}^{-\textrm{i} k \epsilon _1 \tilde {x}_{\perp }} {\textrm d}k . \end{equation}

This matches the form of the incident beam expression in (3.19), if we replace $\tilde {x}_{\perp }$ by $-\tilde {x}_{\perp }$ and include a phase shift of $-\epsilon _2 \pi /2$ . Note that the sign of this phase shift is given by the sign of $N^2+\xi ^2-2\varpi ^2$ at the turning point. For a non-stratified fluid $\varpi = \xi _c$ and for a non-rotating or irrotational flow, $\varpi =N_c$ . In both cases, this sign is negative ( $\epsilon _2 =1$ ), so the phase shift is $-\pi /2$ .

Figure 2.

Contour plot of the function $V_m^S$ in the $(Y_\parallel ,Y_\perp )$ plane for $m=1$ . Dashed black lines indicate the ray trajectory ( $Y_\perp = \pm (-Y_\parallel )^{3/2}$ ) as it reflects at the turning point (black circle at the origin). (a) Real part of $V_m^S$ . (b) Imaginary part of $V_m^{S}$ . (c) Absolute value of $V_m^{S}$ . (d) Phase of $V_m^{S}$ .

5. Comparison with numerical simulations

In this section, we compare the asymptotic predictions of the previous sections with numerical simulations of the linear viscous problem. We consider a non-stratified fluid ( $N=0$ ) that is rotating with a non-uniform angular velocity profile $\varOmega (r)$ . As in the theoretical analysis, the problem is assumed to be axisymmetric about the rotation axis, and we use cylindrical coordinates $(r,z)$ . We focus on the wave dynamics generated by localised forcing at a point $(r_f,z_f)$ within a finite radial interval $(r_i,r_o)$ distant from the axis (i.e. $r_i\gt 0$ ).

Figure 3.

(a) Schematic of the numerical domain. The blue line shows the ray path from the forcing point (red circle) across the turning point (red square) and the ensuing reflected beam. Slices from $S_1$ to $S_4$ correspond to the locations where the comparisons between asymptotics and numerics are performed. The grey area is the damping region necessary to prevent reflections on the boundaries. (b) Transverse velocity $U_{\phi }$ once the harmonic response is reached for $E=10^{-7}$ . A movie showing the time evolution of the global solution and the rescaled local solution can be found in Supplemental movie is available at https://doi.org/10.1017/jfm.2026.11219.

Instead of solving the harmonic perturbation (2.2), we solve the time-dependent linearised equations for the perturbations in the rotating frame with an external forcing term

(5.1a) \begin{align} \frac {\partial \boldsymbol{U}}{\partial t} + 2 \varOmega (\hat {\boldsymbol{e}}_z \times \boldsymbol{U}) + r (\boldsymbol{U} \boldsymbol{\cdot }\boldsymbol{\nabla }\varOmega ) \hat {\boldsymbol{e}}_\phi &= -\boldsymbol{\nabla }P + E {\nabla} ^2 \boldsymbol{U} + \boldsymbol{f}-\frac {\chi }{\tau } \boldsymbol{U} , \end{align}

(5.1b) \begin{align} \boldsymbol{\nabla }\boldsymbol{\cdot }\boldsymbol{U} &= 0 , \end{align}

where $\boldsymbol{U}$ and $P$ are the perturbation velocity and pressure, respectively. The radius $r_f$ and the local rotation rate $\varOmega (r_f)$ are used to non-dimensionalise the variables. The Ekman number is then defined as $E=\nu /(\varOmega (r_f)r_f^2)$ . Note that, as before, we have used the same notation $\varOmega$ for both the dimensional and dimensionless rotation rates.

The external forcing $\boldsymbol{f}$ in (5.1a ) is prescribed as

(5.2) \begin{equation} \boldsymbol{f}=\sin (\varpi t)\exp \big(-(R/R_0)^2\big)\hat {\boldsymbol{e}}_\phi , \end{equation}

where $R=\sqrt {(r-1)^2+(z-z_f)^2)}$ is the distance from the forcing point located at $(1,z_f)$ , and $R_0$ is the characteristic size of the forcing region. We choose $R_0$ small enough for the forcing to be considered nearly point-like, while still allowing the numerical resolution of the inner viscous structure of the fluid response around that point. Typically, we choose $R_0=10^{-3}$ . Note that this forcing procedure does not give direct access to the structure of the emitted shear layer, in particular, we cannot control the singularity strength given by the parameter $m$ in (3.24). As shown below, our particular forcing procedure leads to a shear layer which is accurately modelled as a Moore and Saffman beam with $m=2$ .

Since we solve these equations in a finite square domain $[r_i,r_o]\times [z_i,z_o]$ with no-slip boundary conditions on all sides, we add the last term in (5.1a ) in order to prevent reflections at the boundaries of the domain. This relaxation term ensures a fast decay of waves before they can reach and reflect on the boundaries. The function $\chi$ is zero in the bulk of the domain and reaches unity as soon as the distance from the walls is $10^{-1}$ or less, as indicated in grey in figure 3. The parameter $\tau$ is chosen to ensure a gradual absorption of the waves and minimise reflected waves, typically $\tau =4$ (larger values lead to reflection from the damping layer itself while smaller values lead to waves reflecting from the boundaries).

In the following, we use the following set of parameters. The numerical domain is defined by $r_i=0.8$ and $r_o=1.8$ while the vertical extent is $z\in [0,1]$ . The forcing is centred at $(r_f,z_f)=(1,0.2)$ . The rotation rate of the fluid is non-uniform and varies linearly with the horizontal coordinate according to

(5.3) \begin{equation} \varOmega (r)=1-\alpha (r-1). \end{equation}

Note that this linear variation of the rotation rate is only valid away from the rotation axis. The differential rotation parameter $\alpha$ must be carefully chosen: it must ensure that a turning point lies within the computational domain, while also avoiding centrifugal instability. The latter condition is satisfied by requiring $\xi ^2(r)\gt 0$ throughout the domain, where $\xi$ is the epicyclic frequency defined in (2.4). For the linear rotation profile defined in (5.3), both constraints are satisfied over the radial domain $r\in [0.8,1.8]$ with $\alpha =1/2$ . The Ekman number is fixed to $E=10^{-7}$ and the wave frequency is set to $\varpi = \sqrt {3/2}$ , so that the initial propagation angle at the forcing point is $45^\circ$ . Finally, the remaining parameters are $R_0=10^{-3}$ , which defines the size of the forcing region, and $\tau =4$ , which sets the damping time scale in the outer sponge layer to effectively suppress boundary reflections.

We solve this initial value problem using the open-source spectral element solver Nek5000 (Fischer Reference Fischer1997; Deville, Fischer & Mund Reference Deville, Fischer and Mund2002), which allows for local mesh refinement around the forcing region and the turning point. The domain is discretised using a number $\mathcal{E}=10\,404$ of tetragonal elements while the velocity is discretised within each element using Lagrange polynomial interpolants of order 14. No dealiasing is used since the equations solved are effectively linear. Convergence has been tested by gradually increasing the polynomial order for a fixed number of elements. We focus on the time-harmonic solution obtained after the initial transient has decayed.

Our analysis focuses on the wave beam emitted from the source located at $(r_f,z_f)$ , propagating toward the upper-right quadrant of the domain, see figures 3(a) and 3(b). The equation of the characteristic along which the beam is localised, is derived from (2.10) and takes the form

(5.4) \begin{equation} z^{\prime}(r)=\frac {{\textrm d}z}{{\textrm d}r} = \frac {\sqrt {\Delta }}{\varpi ^2}= \sqrt {5 - 5r +r^2} . \end{equation}

Integrating this expression yields the path of the characteristic

(5.5) \begin{equation} z(r) = \frac {3}{4} + \left (-\frac {5}{4} + \frac {r}{2}\right ) \sqrt {5-5 r + r^2} - \frac {5}{8} \log \left (5-2 r -2\sqrt {5-5 r + r^2}\right ) + z_f , \end{equation}

shown in blue in figure 3(a). The turning point is located at $r_c=(5-\sqrt {5})/2$ . For the present forcing, the structure of the beam is found to be well described by a Moore–Saffman solution with singularity index $m=2$ . The evolution of this solution along the characteristic is described by (3.10), (3.19), (3.22), which lead to the following expression for the parallel velocity component:

(5.6) \begin{equation} u_\parallel \sim A \frac {C_0(r)}{|\tilde {x}_{\parallel }|^{2/3}}h_2(\eta ) , \quad \eta = \frac { F(r)x_\perp }{(E \tilde {x}_\parallel )^{1/3}}, \end{equation}

where

(5.7a) \begin{align} F(r)= \xi (r) & = \varpi \sqrt {1+ z^{\prime 2}(r)} = \sqrt {\frac {3}{2}} \sqrt {6 - 5r +r^2}, \end{align}

(5.7b) \begin{align} C_0(r) = \frac {\sqrt {1+ z^{\prime 2}(r)}}{\sqrt {r z^{\prime}(r)}} & =\frac { \sqrt {6 - 5r +r^2}}{\sqrt {r}(5-5r+r^2)^{1/4}} , \end{align}

(5.7c) \begin{align} \tilde {x}_\parallel = \varpi ^2 \int _{x_\parallel ^{(0)}}^{x_\parallel } \frac {\big(1+ z^{\prime 2}(r)\big)^{3/2}}{ z^{\prime}(r)} {\textrm d}x_\parallel + \tilde {x}_\parallel ^{(0)} & = \frac {3}{2} \int _{1}^r\frac {\big(6 -5r+r^2\big)^2}{\sqrt {5-5r+r^2}}{\textrm d}r + \tilde {x}_\parallel ^{(0)}. \end{align}

The azimuthal velocity $u_\phi$ is related to $u_\parallel$ via (3.3a )

(5.8) \begin{equation} u_\phi = -{\textrm{i}} C_1(r) u_\parallel , \end{equation}

with

(5.9) \begin{equation} C_1(r) = \frac {\omega (r) }{\varpi \sqrt {1+z^{\prime 2}}} = \sqrt {\frac {3}{2}}\frac {(2-r)}{\sqrt {6 -5 r+r^2}} . \end{equation}

The physical azimuthal velocity is then given by

(5.10) \begin{equation} U_\phi (t) = 2|A|\frac {C_0(r)C_1(r)}{|\tilde {x}_{\parallel }|^{2/3}}\textrm{Re}\,\big(h_2(\eta ) \textrm{e}^{-\textrm{i}(\varpi t -\phi _A + \pi /2)}\big ) . \end{equation}

If we consider the quantity defined by

(5.11) \begin{equation} U = \sqrt { U_\phi ^2(t) + U_\phi ^2(t+\pi /(2\varpi ))} = 2 |A| \frac {C_0C_1}{|\tilde {x}_{\parallel }|^{2/3}} |h_2(\eta )|, \end{equation}

we obtain a function that does not depend on time anymore. For each $r$ , its maximum is reached on the characteristic curve $z=z(r)$ (corresponding to $\eta =0$ ) and is given by

(5.12) \begin{equation} U_{max }(r) = 2 |h_2(0)| |A| C_0(r)C_1(r)/|\tilde {x}_{\parallel }(r)|^{2/3}. \end{equation}

This expression can be used to estimate the parameters $|A|$ and $\tilde {x}_\parallel ^{(0)}$ in (5.7c ). The best fit for $r$ between 1.05 and 1.3 is obtained for $|A|\approx 0.0125$ and $\tilde {x}_\parallel ^{(0)} = 0.010$ , as illustrated in figure 4. Albeit a phase factor, the solution is then fully determined.

Figure 4.

Plot of the numerical signal $(\sqrt { U_\theta ^2(t) + U_\theta ^2(t+\pi /(2\varpi ))})$ on the characteristic $z=z(r)$ (grey line), together with $U_{max } (r)$ (black dashed line) given by (5.12) for $|A| = 0.0125$ and $\tilde {x}_\parallel ^{(0)} = 0.01$ . Both incident and reflected beams are shown. The solid curves depart from the dashed curves upon entering the turning-point region.

When the beam reaches the turning point, the approximation (4.23) with $m=2$ has to be used

(5.13) \begin{equation} u_\parallel \sim A C_t V_2(Y_\parallel ,Y_\perp ) . \end{equation}

The amplitude $C_t$ , and local variables $Y_\parallel$ and $ Y_\perp$ , are defined in this case by

(5.14a) \begin{align} C_t& = - \frac { 4 \sqrt {\pi } (3/2)^{7/12}}{5^{1/12} \sqrt {3(5-\sqrt {5})}} \frac {\textrm{e}^{\textrm{i}\pi /4}}{\tilde {x}_{\parallel c}^{13/18} E^{1/18}} \approx - 2.7276 \frac { \textrm{e}^{\textrm{i}\pi /4}}{\tilde {x}_{\parallel c}^{13/18} E^{1/18}}, \end{align}

(5.14b) \begin{align} Y_\parallel & = \frac {5^{1/6} (3/2)^{1/3}(r-r_c)}{|\tilde {x}_{\parallel c}|^{2/9} E^{2/9}} , \end{align}

(5.14c) \begin{align} Y_\perp & = \frac { (3/2)^{1/2}(z-z_c)}{|\tilde {x}_{\parallel c}|^{1/3} E^{1/3}}. \\[9pt] \nonumber \end{align}

The physical azimuthal velocity near the turning point is then given by

(5.15) \begin{equation} U_\phi (t)= 2 |A| |C_t| C_1(r_t) \textrm{Re}\,\big (V_2(Y_\parallel ,Y_\perp ) \textrm{e}^{-\textrm{i} (\varpi t - \phi _A- 3\pi /4)}\big ) . \end{equation}

The reflected beam has a similar expression to the incident beam, but we have to change $x_\perp$ in $-x_\perp$ and add a phase shift of $-\pi /2$ . Moreover, the relation between $u_\parallel$ and $u_\phi$ is now $u_\phi = {\textrm{i}} C_1 u_\parallel$ , so we obtain

(5.16) \begin{equation} U_{\phi R}(t) = 2|A|\frac {C_0(r)C_1(r)}{|\tilde {x}_{\parallel }|^{2/3}}\textrm{Re}\,\big (h_2(-\eta ) \textrm{e}^{-\textrm{i}(\varpi t -\phi _A ) }\big ) . \end{equation}

The functions $C_0$ and $C_1$ remain unchanged for the reflected beam. However, the coordinate $\tilde {x}_\parallel$ differs, and is now given by

(5.17) \begin{equation} \tilde {x}_\parallel = \frac {3}{2} \int _{r}^{r_c}\frac {\big(6 -5r+r^2\big)^2}{\sqrt {5-5r+r^2}}{\textrm d}r + \hat {x}_{\parallel c} + \tilde {x}_\parallel ^{(0)}, \end{equation}

where

(5.18) \begin{equation} \hat {x}_{\parallel c} = \frac {3}{2} \int _{1}^{r_c}\frac {\big(6 -5r+r^2\big)^2}{\sqrt {5-5r+r^2}}{\textrm d}r \approx 1.7663 . \end{equation}

In figure 4, we see that the amplitude of the reflected beam along the characteristic is perfectly described by the quantity $U_{\textit{max}}(r)$ in (5.12), using the values of $|A|$ and $\tilde {x}_{\parallel }^{(0)}$ estimated for the incident beam.

For the comparison of the beam transverse structure, we select four distinct times $t_N= t_o + N \pi /(2\varpi ), N=0,1,2,3$ spanning one oscillation period. The reference time $t_o$ and phase $\phi _A$ are chosen such that

(5.19) \begin{align} \max _{r_1\lt r\lt r_2, z=z_c} |U_\phi (t_o)| \end{align}

is the smallest together with the condition that $U_\phi (t_1) \gt 0$ at the turning point. Theoretically, the first condition is satisfies if $ \textrm{e}^{-\textrm{i}(\varpi t_o - \phi _A- 3\pi /4)} =\pm {\textrm{i}}$ since $V_2(Y_\parallel ,0)$ is real for all $Y_\parallel$ . The second condition implies that the positive sign must be chosen, which determines the phase as $\varpi t_o - \phi _A=\pi /4$ . The solution is now fully specified for all time.

Figure 5.

Structure of incident and reflected beams. Plot of $U_\phi (t_N)$ , $N=0,1,2,3$ normalised by $M=2|A|C_0C_1/(\tilde {x}_\parallel )^{2/3}$ , as a function of the rescaled transverse variable $\eta = \tilde {x}_{\perp } /(E\tilde {x}_{\parallel })^{1/3}$ , in section $S_1$ for the incident beam (a), in section $S_2$ for the reflected beam (b), as shown in figure 3(a). The solid lines represent the theoretical solutions, while the dashed lines correspond to the numerical results. The colours indicate different times: $t=t_0$ (blue), $t=t_1$ (red), $t=t_2$ (green) and $t=t_3$ (black). In these plots, $|A| =0.0125$ and $\tilde {x}_\parallel ^{(0)} = 0.01$ .

Figure 5 compares the numerical results with the theoretical solution for both incident and reflected beams. We observe very good agreement at all four times $t_N$ . The reflected beams curves display the expected phase shift: the structure of the reflected beam at time $t_N$ corresponds to that of the incident beam at time $t_{N-1}$ .

The solution in the turning-point region is shown in figure 6 in both a vertical section (a) and a horizontal section (b) (sections $S_3$ and $S_4$ shown in figure 3 a). Again, we observe good agreement between the numerical solution and the theoretical prediction. Note, however, that the numerical solution in the horizontal section does not vanish for $t=t_0$ and $t=t_2$ , in contrast to the theoretical solution. Capturing this variation would require incorporating higher-order corrections into the theoretical model.

Figure 6.

Structure of the solution in the turning-point region. Plot of $U_\phi (t_N)$ , $N=0,1,2,3$ normalised by $2|A|C_tC_1(r_t)$ in the vertical section $S_3$ (a) as a function of $Y_\perp$ and in the horizontal section $S_4$ (b) as a function of $Y_\parallel$ , see figure 3(a). As in figure 5, the solid lines represent the theoretical solutions and the dashed lines the numerical results. The colours correspond to the same times as before: $t=t_0$ (blue), $t=t_1$ (red), $t=t_2$ (green) and $t=t_3$ (black). The parameters $|A| =0.0125$ and $\tilde {x}_\parallel ^{(0)} = 0.01$ are also the same as in figure 5.