2.1 Depth-averaged equations for the lower potential layer

Integrating the incompressibility, horizontal momentum and energy equations with respect to $z$ over the fluid depth and using the boundary conditions (2.4) and (2.5), we obtain the following exact integral relations for the lower layer:

(2.13) $$\begin{eqnarray}\displaystyle & \displaystyle \frac{\unicode[STIX]{x2202}h}{\unicode[STIX]{x2202}t}+\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}x}\left(\int _{0}^{h}u\,\text{d}z\right)=-M, & \displaystyle\end{eqnarray}$$

(2.14) $$\begin{eqnarray}\displaystyle & \displaystyle \frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}t}\left(\int _{0}^{h}u\,\text{d}z\right)+\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}x}\left(\int _{0}^{h}u^{2}\,\text{d}z+\int _{0}^{h}p\,\text{d}z\right)=p|_{z=h}h_{x}-Mu|_{z=h}, & \displaystyle\end{eqnarray}$$

(2.15) $$\begin{eqnarray}\displaystyle & & \displaystyle \frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}t}\left(\int _{0}^{h}\left(\frac{u^{2}}{2}+\unicode[STIX]{x1D700}^{2}\frac{w^{2}}{2}+z\right)\,\text{d}z\right)+\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}x}\left(\int _{0}^{h}u\left(\frac{u^{2}}{2}+\unicode[STIX]{x1D700}^{2}\frac{w^{2}}{2}+z\right)\,\text{d}z+\int _{0}^{h}pu\,\text{d}z\right)\nonumber\\ \displaystyle & & \displaystyle \qquad =-p|_{z=h}(h_{t}+M)-M\left.\left(\frac{u^{2}}{2}+\unicode[STIX]{x1D700}^{2}\frac{w^{2}}{2}+h\right)\right|_{z=h}.\end{eqnarray}$$

Introducing the average velocity in the lower layer,

(2.16) $$\begin{eqnarray}U=\frac{1}{h}\int _{0}^{h}u\,\text{d}z,\end{eqnarray}$$

we can rewrite the mass conservation law in the following form:

(2.17) $$\begin{eqnarray}\frac{\unicode[STIX]{x2202}h}{\unicode[STIX]{x2202}t}+\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}x}(hU)=-M.\end{eqnarray}$$

To obtain a closed system of governing equations, we need to know the pressure distribution in the layer. Integrating equation (2.1) from $0$ to $z$ , $0<z<h(t,x)$ , we obtain

(2.18) $$\begin{eqnarray}w(t,x,z)=-\int _{0}^{z}u_{x}(t,x,s)\,\text{d}s.\end{eqnarray}$$

In zero order with respect to $\unicode[STIX]{x1D700}$ , the vertical velocity is

(2.19) $$\begin{eqnarray}w(t,x,z)\approx -U_{x}z.\end{eqnarray}$$

A second-order approximation for the pressure comes from (2.10) where we have to replace $w$ by (2.19):

(2.20) $$\begin{eqnarray}\frac{\unicode[STIX]{x2202}p}{\unicode[STIX]{x2202}z}\approx -1-\unicode[STIX]{x1D700}^{2}\left(\frac{\unicode[STIX]{x2202}w}{\unicode[STIX]{x2202}t}+u\frac{\unicode[STIX]{x2202}w}{\unicode[STIX]{x2202}x}+w\frac{\unicode[STIX]{x2202}w}{\unicode[STIX]{x2202}z}\right)\approx -1+\unicode[STIX]{x1D700}^{2}z(U_{xt}+UU_{xx}-U_{x}^{2}).\end{eqnarray}$$

Integrating this from $h$ to $z$ ( $0<z<h$ ) we obtain

(2.21) $$\begin{eqnarray}p\approx p|_{z=h}-(z-h)+\unicode[STIX]{x1D700}^{2}(U_{xt}+UU_{xx}-U_{x}^{2})\frac{z^{2}-h^{2}}{2}.\end{eqnarray}$$

The integral of the pressure can thus be evaluated as

(2.22) $$\begin{eqnarray}\int _{0}^{h}p\,\text{d}z\approx p|_{z=h}\,h+\frac{h^{2}}{2}-\unicode[STIX]{x1D700}^{2}\frac{h^{3}}{3}(U_{xt}+UU_{xx}-U_{x}^{2}).\end{eqnarray}$$

One can also prove that, if the flow is weakly sheared, i.e. the dimensionless flow vorticity $\unicode[STIX]{x1D714}=u_{z}-\unicode[STIX]{x1D700}^{2}w_{x}=O(\unicode[STIX]{x1D700}^{\unicode[STIX]{x1D6FD}})$ with $1<\unicode[STIX]{x1D6FD}\leqslant 2$ , then

(2.23) $$\begin{eqnarray}\int _{0}^{h}u^{2}\,\text{d}z=hU^{2}+O(\unicode[STIX]{x1D700}^{2\unicode[STIX]{x1D6FD}})\end{eqnarray}$$

(for the proof, see Barros, Gavrilyuk & Teshukov (Reference Barros, Gavrilyuk and Teshukov2007) or Gavrilyuk, Kalisch & Khorsand (Reference Gavrilyuk, Kalisch and Khorsand2015)). In particular, in the potential case, $\unicode[STIX]{x1D6FD}=2$ . Analogous estimation of integrals can be given for the energy equation. Keeping only terms of order one and $\unicode[STIX]{x1D700}^{2}$ , we obtain the final system for a lower potential layer with mixing at the interface $z=h(t,x)$ :

(2.24) $$\begin{eqnarray}\displaystyle & \displaystyle \frac{\unicode[STIX]{x2202}h}{\unicode[STIX]{x2202}t}+\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}x}(hU)=-M, & \displaystyle\end{eqnarray}$$

(2.25) $$\begin{eqnarray}\displaystyle & \displaystyle \frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}t}(hU)+\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}x}\left(hU^{2}+\int _{0}^{h}p\,\text{d}z\right)=p|_{z=h}\,h_{x}-Mu|_{z=h}, & \displaystyle\end{eqnarray}$$

(2.26) $$\begin{eqnarray}\displaystyle & & \displaystyle \frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}t}\left(h\left(\frac{U^{2}}{2}+\unicode[STIX]{x1D700}^{2}\frac{U_{x}^{2}}{6}h^{2}+\frac{h}{2}\right)\right)+\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}x}\left(hU\left(\frac{U^{2}}{2}+\unicode[STIX]{x1D700}^{2}\frac{U_{x}^{2}}{6}h^{2}+\frac{h}{2}\right)+U\int _{0}^{h}p\,\text{d}z\right)\nonumber\\ \displaystyle & & \displaystyle \qquad =-p|_{z=h}(h_{t}+M)-M\left.\left(\frac{u^{2}}{2}+\unicode[STIX]{x1D700}^{2}\frac{U_{x}^{2}h^{2}}{2}+h\right)\right|_{z=h},\end{eqnarray}$$

where

(2.27) $$\begin{eqnarray}\int _{0}^{h}p\,\text{d}z=p|_{z=h}\,h+\frac{h^{2}}{2}-\unicode[STIX]{x1D700}^{2}\frac{h^{3}}{3}(U_{xt}+UU_{xx}-U_{x}^{2}).\end{eqnarray}$$

The value of the velocity $u|_{z=h}$ is not yet determined. Let us remark that we have formally three governing equations for only two unknowns $h$ and $U$ . The compatibility condition between the energy, momentum and mass equations gives us only one possibility (for proof, see appendix A):

(2.28) $$\begin{eqnarray}u|_{z=h}=U.\end{eqnarray}$$

2.2 Depth-averaged equations for the upper turbulent layer

Consider the following hydrostatic equations in the upper layer having the same density as the lower layer. In dimensionless variables, the equations are

(2.29) $$\begin{eqnarray}\displaystyle & \displaystyle \frac{\unicode[STIX]{x2202}u}{\unicode[STIX]{x2202}x}+\frac{\unicode[STIX]{x2202}w}{\unicode[STIX]{x2202}z}=0, & \displaystyle\end{eqnarray}$$

(2.30) $$\begin{eqnarray}\displaystyle & \displaystyle \frac{\unicode[STIX]{x2202}u}{\unicode[STIX]{x2202}t}+\frac{\unicode[STIX]{x2202}u^{2}}{\unicode[STIX]{x2202}x}+\frac{\unicode[STIX]{x2202}uw}{\unicode[STIX]{x2202}z}=-\frac{\unicode[STIX]{x2202}p}{\unicode[STIX]{x2202}x},\quad p(t,x,z)=h+\unicode[STIX]{x1D702}-z. & \displaystyle\end{eqnarray}$$

The hydrostatic distribution of the pressure already takes into account the dynamic condition at the free surface. They admit the energy conservation law:

(2.31) $$\begin{eqnarray}\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}t}\left(\frac{u^{2}}{2}+z\right)+\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}x}\left(u\left(\frac{u^{2}}{2}+z\right)+pu\right)+\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}z}\left(w\left(\frac{u^{2}}{2}+z\right)+pw\right)=0.\end{eqnarray}$$

The corresponding kinematic boundary conditions at the free surface and at the internal surface are

(2.32a,b ) $$\begin{eqnarray}\left.\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}t}(h+\unicode[STIX]{x1D702})+u\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}x}(h+\unicode[STIX]{x1D702})-w\,\right|_{z=h+\unicode[STIX]{x1D702}}=0,\quad \frac{\unicode[STIX]{x2202}h}{\unicode[STIX]{x2202}t}+\left.u\frac{\unicode[STIX]{x2202}h}{\unicode[STIX]{x2202}x}-w\,\right|_{z=h}=-M.\end{eqnarray}$$

Averaging the incompressibility equation, we obtain

(2.33) $$\begin{eqnarray}\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}t}\unicode[STIX]{x1D702}+\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}x}(\unicode[STIX]{x1D702}\bar{u})=M,\end{eqnarray}$$

with

(2.34) $$\begin{eqnarray}\bar{u}(t,x)=\frac{1}{\unicode[STIX]{x1D702}}\int _{h}^{h+\unicode[STIX]{x1D702}}u(t,x,s)\,\text{d}s.\end{eqnarray}$$

Averaging of the horizontal momentum equation gives us

(2.35) $$\begin{eqnarray}\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}t}\left(\int _{h}^{h+\unicode[STIX]{x1D702}}u\,\text{d}z\right)+\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}x}\left(\int _{h}^{h+\unicode[STIX]{x1D702}}(u^{2}+p)\,\text{d}z\right)=M\,u|_{z=h}-p|_{z=h}\,h_{x},\end{eqnarray}$$

or

(2.36) $$\begin{eqnarray}\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}t}(\unicode[STIX]{x1D702}\bar{u})+\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}x}\left(\unicode[STIX]{x1D702}\bar{u}^{2}+\unicode[STIX]{x1D702}e+\frac{\unicode[STIX]{x1D702}^{2}}{2}\right)=M\,u|_{z=h}-p|_{z=h}\,h_{x},\end{eqnarray}$$

with the specific turbulent energy $e$ defined as

(2.37) $$\begin{eqnarray}e=\frac{1}{\unicode[STIX]{x1D702}}\int _{h}^{h+\unicode[STIX]{x1D702}}(u-\bar{u})^{2}\,\text{d}z.\end{eqnarray}$$

The quantity $e$ actually measures the flow dispersiveness. In Antuono & Brocchini (Reference Antuono and Brocchini2013) the evaluation of this term has been given for different flow regimes. Instead, we will obtain the evolution equation for $e$ . The averaged energy equation is

(2.38) $$\begin{eqnarray}\displaystyle & & \displaystyle \frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}t}\left(\int _{h}^{h+\unicode[STIX]{x1D702}}\left(\frac{u^{2}}{2}+z\right)\,\text{d}z\right)+\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}x}\left(\int _{h}^{h+\unicode[STIX]{x1D702}}\left(u\left(\frac{u^{2}}{2}+z\right)+pu\right)\,\text{d}z\right)\nonumber\\ \displaystyle & & \displaystyle \qquad =M\left(\left.\frac{u^{2}}{2}\right|_{z=h}+h\right)+p|_{z=h}(h_{t}+M).\end{eqnarray}$$

We suppose that the shear effects in the upper layer are now more important. More exactly, the dimensionless flow vorticity is $\unicode[STIX]{x1D714}=u_{z}-\unicode[STIX]{x1D700}^{2}w_{x}\approx u_{z}$ . Let us assume that initially $u_{z}=O(\unicode[STIX]{x1D700}^{\unicode[STIX]{x1D6FD}})$ with $0<\unicode[STIX]{x1D6FD}<1$ . One can prove that this property is then valid for any time. It also implies that

(2.39) $$\begin{eqnarray}\int _{h}^{h+\unicode[STIX]{x1D702}}(u-\bar{u})^{2}\,\text{d}z=O(\unicode[STIX]{x1D700}^{2\unicode[STIX]{x1D6FD}}).\end{eqnarray}$$

This term is now more important than the dispersion, which is of order $\unicode[STIX]{x1D700}^{2}$ , and should be kept in the momentum equation. The vorticity is ‘weak’ in the sense that it is not of order one, but it could be quite ‘large’ (‘almost’ of order one) when $\unicode[STIX]{x1D6FD}$ is small. Analogous estimation of integrals can be given for the energy equation (for details, see Teshukov Reference Teshukov2007; Richard & Gavrilyuk Reference Richard and Gavrilyuk2012, Reference Richard and Gavrilyuk2013). Keeping only terms of order one and $\unicode[STIX]{x1D700}^{2\unicode[STIX]{x1D6FD}}$ , we obtain the energy equation for the upper turbulent layer in the form

(2.40) $$\begin{eqnarray}\displaystyle & & \displaystyle \frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}t}\left(\unicode[STIX]{x1D702}\left(\frac{\bar{u}^{2}}{2}+\frac{e}{2}+\frac{\unicode[STIX]{x1D702}+2h}{2}\right)\right)+\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}x}\left(\unicode[STIX]{x1D702}\bar{u}\left(\frac{\bar{u}^{2}}{2}+e+\frac{\unicode[STIX]{x1D702}+2h}{2}\right)+\frac{\unicode[STIX]{x1D702}^{2}}{2}\bar{u}\right)\nonumber\\ \displaystyle & & \displaystyle \qquad =M\left(\left.\frac{u^{2}}{2}\right|_{z=h}+h\right)+p|_{z=h}(h_{t}+M).\end{eqnarray}$$

Since the pressure distribution in the upper layer is linear with $z$ (the pressure is hydrostatic), one obviously has

(2.41) $$\begin{eqnarray}p|_{z=h}=\unicode[STIX]{x1D702}.\end{eqnarray}$$

Remark 1. The rate at which the potential layer is ‘contaminated’ by the vorticity (the term $M$ ) should be specified. We take it in the form

(2.42) $$\begin{eqnarray}M=\unicode[STIX]{x1D70E}\sqrt{e},\quad \unicode[STIX]{x1D70E}=\text{const}.\end{eqnarray}$$

The justification for (2.42) can be done in the following way. Consider two co-flowing horizontal fluid layers having the same velocity. The upper layer is turbulent, and the lower one is potential. We denote the classical Reynolds averaging of the dimensionless Euler equations (with $\unicode[STIX]{x1D700}=1$ ) by angle brackets $\langle \cdots \rangle$ , with a classical representation for any $f$ in the form $f=\langle f\rangle +\tilde{f}$ , where $\langle \tilde{f}\rangle =0$ . The corresponding dimensionless Reynolds equations are

(2.43) $$\begin{eqnarray}\displaystyle \left.\begin{array}{@{}c@{}}\langle u\rangle _{x}+\langle w\rangle _{z}=0,\quad \langle u\rangle _{t}+(\langle u\rangle ^{2}+\langle \tilde{u} ^{2}\rangle +\langle p\rangle )_{x}+(\langle u\rangle \langle w\rangle +\langle \tilde{u} \tilde{w}\rangle )_{z}=0,\\ \langle w\rangle _{t}+(\langle u\rangle \langle w\rangle +\langle \tilde{u} \tilde{w}\rangle )_{x}+(\langle w\rangle ^{2}+\langle \tilde{w}^{2}\rangle +\langle p\rangle )_{z}=-1,\\ \langle E\rangle _{t}+(\langle u\rangle \langle E\rangle +\langle \tilde{u} \tilde{E}\rangle +\langle p\rangle \langle u\rangle +\langle \tilde{p}\tilde{u} \rangle )_{x}\qquad \;\\ \qquad +(\langle w\rangle \langle E\rangle +\langle \tilde{w}\tilde{E}\rangle +\langle p\rangle \langle w\rangle +\langle \tilde{p}\tilde{w}\rangle )_{z}=-\langle w\rangle .\end{array}\right\} & & \displaystyle\end{eqnarray}$$

Here

(2.44) $$\begin{eqnarray}\displaystyle \left.\begin{array}{@{}c@{}}\displaystyle E=\frac{u^{2}+w^{2}}{2},\quad \langle E\rangle =\frac{\langle u\rangle ^{2}+\langle w\rangle ^{2}}{2}+\frac{k}{2},\quad k=\langle \tilde{u} ^{2}\rangle +\langle \tilde{w}^{2}\rangle ,\\ \displaystyle \tilde{E}=\langle u\rangle \tilde{u} +\langle w\rangle \tilde{w}+\frac{\tilde{u} ^{2}+\tilde{w}^{2}}{2}-\frac{k}{2},\end{array}\right\} & & \displaystyle\end{eqnarray}$$

the subscripts $t$ , $x$ and $z$ mean differentiation with respect to $t$ , $x$ and $z$ , respectively, and $k$ equals twice the kinetic turbulent energy. Consider now particular solutions of (2.43) where all the variables depend only on $(t,z)$ , and the term $\langle u\rangle \langle w\rangle$ is negligible with respect to the Reynolds shear stress. Also, we will assume that the third-order correlations and the correlation $\langle \tilde{p}\tilde{w}\rangle$ are negligible. Such a solution is suitable for the description of a horizontal mixing layer. Equations (2.43) imply that

(2.45a,b ) $$\begin{eqnarray}\langle u\rangle _{t}+(\langle \tilde{u} \tilde{w}\rangle )_{z}=0,\quad \left(\frac{\langle u\rangle ^{2}}{2}+\frac{k}{2}\right)_{t}+(\langle \tilde{u} \tilde{w}\rangle \langle u\rangle )_{z}=0.\end{eqnarray}$$

System (2.45) simplifying (2.43) with the closure $-\langle \tilde{u} \tilde{w}\rangle =\unicode[STIX]{x1D70E}k$ , where $\unicode[STIX]{x1D70E}=\text{const}.$ is the shear coefficient, is hyperbolic, with the characteristic slopes $\pm \unicode[STIX]{x1D70E}\sqrt{2k}$ . In a vertically sheared flow, the sign of $\unicode[STIX]{x1D70E}$ coincides with the sign of the vertical mean velocity gradient. The value of $\unicode[STIX]{x1D70E}$ is approximately $0.15$ (Townsend Reference Townsend1956; Bradshaw, Ferriss & Atwell Reference Bradshaw, Ferriss and Atwell1967). The mixing front between the turbulent layer and the potential layer can be seen as a shock separating the two regions. Applying the Rankine–Hugoniot relations coming from (2.45) we obtain

(2.46a,b ) $$\begin{eqnarray}D[\langle u\rangle ]+\unicode[STIX]{x1D70E}[k]=0,\quad D\left[\frac{\langle u\rangle ^{2}}{2}+\frac{k}{2}\right]+\unicode[STIX]{x1D70E}[k\langle u\rangle ]=0.\end{eqnarray}$$

Here the square brackets mean the jump of variables, and $D$ is the front velocity. Let $u^{-},k^{-}$ and $u^{+},k^{+}=0$ be the values of the unknowns in the turbulent and potential regions, respectively, with $u^{-}>u^{+}$ . Then we have

(2.47a,b ) $$\begin{eqnarray}k^{-}=(u^{-}-u^{+})^{2},\quad D=-\unicode[STIX]{x1D70E}\frac{(u^{-}-u^{+})^{2}}{u^{-}-u^{+}}=-\unicode[STIX]{x1D70E}\sqrt{k^{-}}.\end{eqnarray}$$

This is exactly the empirical relation (2.42) for $M$ , if we identify the turbulent energy $k$ with the energy $e$ measuring the flow dispersiveness, and the shock front velocity $D$ with $M$ . The ‘minus’ sign corresponds to the fact that $\unicode[STIX]{x1D702}$ is growing at the same time as $z$ is decreasing. Such a shock is stable in the sense of Lax (the characteristics overlap at the shock). A detailed discussion can also be found in Liapidevskii & Teshukov (Reference Liapidevskii and Teshukov2000).

Remark 2. Besides the dissipation term coming from the mixing at the boundary between the layers, it is necessary to add the bulk dissipation term in the turbulent layer, describing the turbulent energy dissipation toward smaller scales (smaller than the scale of large vortices defined by $\unicode[STIX]{x1D702}$ ). The most simple and classical expression for the energy dissipation rate $d$ can be taken in the form (Townsend Reference Townsend1956)

(2.48) $$\begin{eqnarray}d=\unicode[STIX]{x1D70E}\frac{\unicode[STIX]{x1D705}}{2}e^{3/2}=M\frac{\unicode[STIX]{x1D705}}{2}e.\end{eqnarray}$$

Here the coefficient $\unicode[STIX]{x1D705}/2$ is of order one (it is a ‘calibration’ coefficient to the model); the factor $1/2$ is added for convenience. The shear coefficient $\unicode[STIX]{x1D70E}$ has an effect on the length of surface waves, while the dissipation coefficient $\unicode[STIX]{x1D705}$ is responsible for the rate of turbulent layer thickness development. The variation of $\unicode[STIX]{x1D705}$ between $2$ and $6$ did not show a qualitative and quantitative modification of solution properties. The value $\unicode[STIX]{x1D705}=3$ was chosen for comparison of the model predictions to experimental data. As before, the energy $e$ is given by (2.37).

2.3 Final two-layer system

The final dimensionless system for the thickness of the upper turbulent layer $\unicode[STIX]{x1D702}$ , its average velocity $\bar{u}$ , energy $e$ , thickness of the low potential layer $h$ and its average velocity $U$ can be written as

(2.49) $$\begin{eqnarray}\frac{\unicode[STIX]{x2202}\unicode[STIX]{x1D702}}{\unicode[STIX]{x2202}t}+\frac{\unicode[STIX]{x2202}\unicode[STIX]{x1D702}\bar{u}}{\unicode[STIX]{x2202}x}=M,\end{eqnarray}$$

(2.50) $$\begin{eqnarray}\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}t}(\unicode[STIX]{x1D702}\bar{u})+\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}x}\left(\unicode[STIX]{x1D702}\bar{u}^{2}+\unicode[STIX]{x1D702}e+\frac{\unicode[STIX]{x1D702}^{2}}{2}\right)=MU-\unicode[STIX]{x1D702}h_{x},\end{eqnarray}$$

(2.51) $$\begin{eqnarray}\displaystyle & & \displaystyle \frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}t}\left(\unicode[STIX]{x1D702}\left(\frac{\bar{u}^{2}}{2}+\frac{e}{2}+\frac{\unicode[STIX]{x1D702}+2h}{2}\right)\right)+\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}x}\left(\unicode[STIX]{x1D702}\bar{u}\left(\frac{\bar{u}^{2}}{2}+\frac{e}{2}+\frac{\unicode[STIX]{x1D702}+2h}{2}\right)+\left(\unicode[STIX]{x1D702}e+\frac{\unicode[STIX]{x1D702}^{2}}{2}\right)\bar{u}\right)\nonumber\\ \displaystyle & & \displaystyle \qquad =M\left(\frac{U^{2}}{2}+h\right)+\unicode[STIX]{x1D702}(h_{t}+M)-d,\end{eqnarray}$$

(2.52) $$\begin{eqnarray}\frac{\unicode[STIX]{x2202}h}{\unicode[STIX]{x2202}t}+\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}x}(hU)=-M,\end{eqnarray}$$

(2.53) $$\begin{eqnarray}\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}t}(hU)+\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}x}\left(hU^{2}+\int _{0}^{h}p\,\text{d}z\right)=\unicode[STIX]{x1D702}h_{x}-MU,\end{eqnarray}$$

(2.54) $$\begin{eqnarray}\displaystyle & & \displaystyle \frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}t}\left(h\left(\frac{U^{2}}{2}+\unicode[STIX]{x1D700}^{2}\frac{U_{x}^{2}}{6}h^{2}+\frac{h}{2}\right)\right)+\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}x}\left(hU\left(\frac{U^{2}}{2}+\unicode[STIX]{x1D700}^{2}\frac{U_{x}^{2}}{6}h^{2}+\frac{h}{2}\right)+U\int _{0}^{h}p\,\text{d}z\right)\nonumber\\ \displaystyle & & \displaystyle \qquad =-\unicode[STIX]{x1D702}(h_{t}+M)-M\left(\frac{U^{2}}{2}+\unicode[STIX]{x1D700}^{2}\frac{U_{x}^{2}h^{2}}{2}+h\right),\end{eqnarray}$$

with

(2.55a-c ) $$\begin{eqnarray}\int _{0}^{h}p\,\text{d}z=\unicode[STIX]{x1D702}h+\frac{h^{2}}{2}-\unicode[STIX]{x1D700}^{2}\frac{h^{3}}{3}(U_{xt}+UU_{xx}-U_{x}^{2}),\quad M=\unicode[STIX]{x1D70E}\sqrt{e},\quad d=M\frac{\unicode[STIX]{x1D705}}{2}e.\end{eqnarray}$$

It is shown in appendix A that equation (2.54) (the energy balance for the lower layer) is not independent – it is a consequence of (2.52) and (2.53). It is added just to have a ‘symmetric’ presentation of the governing equations. Summing the momentum equations for each layer (2.50) and (2.53), and the energy balances (2.51) and (2.54), one immediately obtains the total momentum conservation law,

(2.56) $$\begin{eqnarray}\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}t}(\unicode[STIX]{x1D702}\bar{u}+hU)+\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}x}\left(\unicode[STIX]{x1D702}\bar{u}^{2}+\unicode[STIX]{x1D702}e+\frac{\unicode[STIX]{x1D702}^{2}}{2}+hU^{2}+\int _{0}^{h}p\,\text{d}z\right)=0,\end{eqnarray}$$

and the total energy balance equation,

(2.57) $$\begin{eqnarray}\displaystyle & & \displaystyle \frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}t}\left(\unicode[STIX]{x1D702}\left(\frac{\bar{u}^{2}}{2}+\frac{e}{2}+\frac{\unicode[STIX]{x1D702}+2h}{2}\right)+h\left(\frac{U^{2}}{2}+\unicode[STIX]{x1D700}^{2}\frac{U_{x}^{2}}{6}h^{2}+\frac{h}{2}\right)\right)\nonumber\\ \displaystyle & & \displaystyle \qquad +\,\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}x}\left(\unicode[STIX]{x1D702}\bar{u}\left(\frac{\bar{u}^{2}}{2}+\frac{e}{2}+\frac{\unicode[STIX]{x1D702}+2h}{2}\right)+\left(\unicode[STIX]{x1D702}e+\frac{\unicode[STIX]{x1D702}^{2}}{2}\right)\bar{u}\right.\nonumber\\ \displaystyle & & \displaystyle \left.\qquad +\,hU\left(\frac{U^{2}}{2}+\unicode[STIX]{x1D700}^{2}\frac{U_{x}^{2}}{6}h^{2}+\frac{h}{2}\right)+U\int _{0}^{h}p\,\text{d}z\right)\nonumber\\ \displaystyle & & \displaystyle \quad =-M\unicode[STIX]{x1D700}^{2}\frac{U_{x}^{2}h^{2}}{2}-d.\end{eqnarray}$$

Since $M>0$ (the thickness of the upper layer is increasing because of the mixing) and $d>0$ , the total energy is decreasing.

2.4 An equivalent form of the momentum equation for the lower layer

The flux in the momentum equation (2.53) for the lower layer contains the higher-order derivatives of the velocity $U$ . It appears that for numerical purposes the following new variable

(2.58) $$\begin{eqnarray}K=U-\frac{\unicode[STIX]{x1D700}^{2}}{3h}(h^{3}U_{x})_{x}\end{eqnarray}$$

is more convenient for the model formulation. The variable $K$ is nothing but the tangential velocity along the free surface (cf. Gavrilyuk & Teshukov Reference Gavrilyuk and Teshukov2001; Bardos & Lannes Reference Bardos and Lannes2012; Gavrilyuk et al. Reference Gavrilyuk, Kalisch and Khorsand2015). For this, we replace the momentum equation by an equivalent equation (for proof see appendix B):

(2.59) $$\begin{eqnarray}K_{t}+\left(KU+h+\unicode[STIX]{x1D702}-\frac{U^{2}}{2}-\frac{\unicode[STIX]{x1D700}^{2}}{2}U_{x}^{2}h^{2}-\unicode[STIX]{x1D700}^{2}hMU_{x}\right)_{x}=M\left(\frac{K-U}{h}+\unicode[STIX]{x1D700}^{2}h_{x}U_{x}\right).\end{eqnarray}$$

In the case where the mixing is absent ( $M=0$ ), this equation is reduced to the conservative one:

(2.60) $$\begin{eqnarray}K_{t}+\left(KU+h+\unicode[STIX]{x1D702}-\frac{U^{2}}{2}-\frac{\unicode[STIX]{x1D700}^{2}}{2}U_{x}^{2}h^{2}\right)_{x}=0.\end{eqnarray}$$

2.5 Production of the turbulent energy

An evolution equation for the turbulent energy $e$ in the upper layer can also be obtained. Consider the subsystem for the upper layer (2.49)–(2.51). Using the equations of mass (2.49) and momentum (2.50), one can obtain from (2.51):

(2.61) $$\begin{eqnarray}\frac{\unicode[STIX]{x1D702}}{2}\frac{\bar{\text{D}}e}{\text{D}t}+e\left(M-\frac{\bar{\text{D}}\unicode[STIX]{x1D702}}{\text{D}t}\right)=M\left(\frac{(U-\bar{u})^{2}}{2}-\frac{e}{2}\right)-d,\quad \frac{\bar{\text{D}}}{\text{D}t}=\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}t}+\bar{u}\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}x}.\end{eqnarray}$$

This is equivalent to

(2.62) $$\begin{eqnarray}\frac{\unicode[STIX]{x1D702}}{2}\frac{\bar{\text{D}}e}{\text{D}t}-e\frac{\bar{\text{D}}\unicode[STIX]{x1D702}}{\text{D}t}=M\left(\frac{(U-\bar{u})^{2}}{2}-\frac{3}{2}e\right)-d,\end{eqnarray}$$

which implies that

(2.63) $$\begin{eqnarray}\frac{\bar{\text{D}}}{\text{D}t}\left(\frac{e}{\unicode[STIX]{x1D702}^{2}}\right)=\frac{M}{\unicode[STIX]{x1D702}^{3}}\left((U-\bar{u})^{2}-(3+\unicode[STIX]{x1D705})e\right).\end{eqnarray}$$

For numerics, a conservative equation for $q$ will be used. Replacing $e$ by $q^{2}$ and using the equation for $\unicode[STIX]{x1D702}$ , one can derive a conservative equation for $q=\sqrt{e}$ :

(2.64) $$\begin{eqnarray}\frac{\unicode[STIX]{x2202}q}{\unicode[STIX]{x2202}t}+\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}x}(q\bar{u})=\frac{\unicode[STIX]{x1D70E}}{2\unicode[STIX]{x1D702}}((U-\bar{u})^{2}-(1+\unicode[STIX]{x1D705})q^{2}).\end{eqnarray}$$

2.6 Flow over topography

The derivation of the governing equations for the flow over topography is analogous. However, the final system contains second-order material derivatives of $b$ along the mean velocity of the lower layer. The equations can be simplified considerably under the following assumption. Let the dimensionless bottom variation be weak: $z=b(\unicode[STIX]{x1D700}^{\unicode[STIX]{x1D6FC}}x),\unicode[STIX]{x1D6FC}>0$ . Then one can obviously neglect the terms containing second-order material derivatives of $b$ and keep in the right-hand side of the dimensionless momentum equation for the upper layer only the term $-\unicode[STIX]{x1D702}b_{x}$ , and in the dimensionless equation for $K$ the term $-b_{x}$ . Such a mild slope approximation was used, for example, in Serre (Reference Serre1953) and Liapidevskii and Gavrilova (Reference Liapidevskii and Gavrilova2008).

3.1 The structure of stationary solutions

The mixing process is dissipative, so the stationary solution is rather an undular bore, and not a sequence of solitary (or periodic) waves. The upper layer develops quite slowly if the flow is supercritical everywhere (the determinant (3.8) is positive). A typical flow picture is shown in figure 2.

Figure 2.

A typical supercritical stationary flow for Froude number $F=1.2$ . The lower boundary separates potential and turbulent layers. The mixing region of thickness $\unicode[STIX]{x1D702}$ is gradually increasing. The parameters are as follows: $H_{0}=0.1~\text{m}$ , $U_{0}=1.19~\text{m}~\text{s}^{-1}$ , $g=9.81~\text{m}~\text{s}^{-2}$ , $\unicode[STIX]{x1D70E}=0.15$ and $\unicode[STIX]{x1D705}=0$ .

There exists a critical Froude number $F_{\star }$ such that the flow becomes critical (the determinant (3.8) vanishes) at the crest of the leading wave. This value can easily be determined numerically. For example, $F_{\star }\approx 1.338$ for $\unicode[STIX]{x1D705}=0$ .

One can also find another critical value $F_{\star \star }$ for which the determinant vanishes at the crest of the second wave. For example, $F_{\star \star }\approx 1.283$ for $\unicode[STIX]{x1D705}=0$ . A sequence of the corresponding critical values can thus be constructed. The smaller the Froude number (but always larger than one), the larger is the domain where the solution is supercritical. The critical Froude numbers increase with increasing $\unicode[STIX]{x1D705}$ . For example, for $\unicode[STIX]{x1D705}=3$ , one has $F_{\star }\approx 1.3832$ and $F_{\star \star }\approx 1.3201$ .

When the supercritical–subcritical transition happens, i.e. the determinant (3.8) changes sign from positive to negative, the solution may contain a jump of the variables of the upper turbulent layer. The thickness of the upper layer in the subcritical local zone will rapidly increase with increasing distance downstream because of the mixing process. The pressure distribution approaches the hydrostatic one because the turbulent layer will expand. The transition can occur not necessarily at the leading wave, but, for example, at the second wave. Physically, this corresponds to a possibility of the roller appearance at the second wave of the wave train, and not at the first one. This fact was observed, in particular, in the experiments by Ryabenko (Reference Ryabenko1990) performed in a wide channel. The wide channel experiments correspond better to our modelling because the sidewall effects can be neglected. For narrow channels, the sidewall effects are dominant, and the roller appears first always at the leading wave (see the classification of the flow patterns for undular hydraulic jumps in Chanson & Montes (Reference Chanson and Montes1995)).

Now we will show how to construct another solution containing a jump riding down the forward slope of the leading wave even for $F<F_{\star }$ . The construction of such a solution can be done as follows. Take any Froude number $F<F_{\star }$ . We put the jump at a point $x=x_{0}$ , which should be properly chosen. For given values $\unicode[STIX]{x1D702}_{0}=\unicode[STIX]{x1D702}(x_{0})$ , $\bar{u}_{0}=\bar{u}(x_{0})$ and $q_{0}=q(x_{0})$ , one can use the Rankine–Hugoniot relations (3.9):

(3.21) $$\begin{eqnarray}\left.\begin{array}{@{}c@{}}\displaystyle \unicode[STIX]{x1D702}\bar{u}=\unicode[STIX]{x1D702}_{0}\bar{u}_{0}=Q,\quad \unicode[STIX]{x1D702}\bar{u}^{2}+\unicode[STIX]{x1D702}q^{2}+\frac{g}{2}\unicode[STIX]{x1D702}^{2}=\unicode[STIX]{x1D702}_{0}{\bar{u}_{0}}^{2}+\unicode[STIX]{x1D702}_{0}q_{0}^{2}+\frac{g}{2}\unicode[STIX]{x1D702}_{0}^{2},\\ \displaystyle \frac{\bar{u}^{2}}{2}+\frac{3}{2}q^{2}+g\unicode[STIX]{x1D702}=\frac{\bar{u}_{0}^{2}}{2}+\frac{3}{2}q_{0}^{2}+g\unicode[STIX]{x1D702}_{0}.\end{array}\right\}\end{eqnarray}$$

Eliminating the velocities, one can obtain a quadratic equation for the layer thickness after the jump:

(3.22) $$\begin{eqnarray}\frac{g}{2}\unicode[STIX]{x1D702}_{0}^{2}\unicode[STIX]{x1D702}^{2}-\unicode[STIX]{x1D702}\left(Q^{2}+3\unicode[STIX]{x1D702}_{0}\left(\frac{1}{2}g\unicode[STIX]{x1D702}_{0}^{2}+\unicode[STIX]{x1D702}_{0}q_{0}^{2}\right)\right)+2\unicode[STIX]{x1D702}_{0}Q^{2}=0.\end{eqnarray}$$

The equation has two positive roots. Only one of them (the minimal one), having the property $\unicode[STIX]{x1D702}<2\unicode[STIX]{x1D702}_{0}$ , must be chosen (see the discussion in Richard & Gavrilyuk (Reference Richard and Gavrilyuk2012)). Since the value of $\unicode[STIX]{x1D702}_{0}$ is almost negligible, the jump of $\unicode[STIX]{x1D702}$ is almost invisible. However, the corresponding turbulent energy considerably increases. The flow in the turbulent region becomes subcritical. The thickness of the turbulent layer increases, the flow in this layer decelerates and the determinant (3.8) in the equation for $\unicode[STIX]{x1D702}$ vanishes at some point $x=x_{1}$ . If at the same point the numerator vanishes, the solution can be extended again to the supercritical regime for $x>x_{1}$ . The value of $x=x_{0}$ can always be chosen, and thus a new solution containing a local subcritical zone between $x_{0}$ and $x_{1}$ can be constructed. Such a new solution is shown in figure 3 for the Froude number $F=1.2$ . The local subcritical zone can be associated with wave breaking.

It is interesting to note that the local wavelength of the stationary wave train strongly decreases with increasing distance downstream. This is in accordance with the experimental observations of Lemoine (Reference Lemoine1948, p. 185). In particular, he commented on this fact: que l’on voit $\unicode[STIX]{x1D706}$ décroître en s’éloignant du front d’onde [we see that $\unicode[STIX]{x1D706}$ decreases away from the wavefront]. Here $\unicode[STIX]{x1D706}$ denotes the wavelength. A physical explanation of this fact is the following. The appearance of a localized subcritical turbulent region is a highly dissipative phenomenon. It intensifies the mixing process and thus leads to a drastic decrease of the amplitude and wavelength.

Figure 3.

A new supercritical–subcritical stationary solution for Froude number $F=1.2$ . The subcritical turbulent region rapidly increases after the jump. The flow continuously passes to the supercritical regime. The flow parameters are as follows: $H_{0}=0.1~\text{m}$ , $U_{0}=1.19~\text{m}~\text{s}^{-1}$ , $g=9.81~\text{m}~\text{s}^{-2}$ , $\unicode[STIX]{x1D70E}=0.15$ and $\unicode[STIX]{x1D705}=0$ . The supercritical–subcritical transition is at the point $x_{0}\approx -0.124$ ; the subcritical–supercritical transition is at the point $x_{1}\approx 0.025$ .

These two solutions have completely different properties. The first one, without the jump, has long modulations and quite narrow mixing region (figure 2). The second one, with the jump, has shorter modulations at the same space interval, smaller amplitude and thicker mixing zone (see figure 3). The same boundary conditions at negative infinity can thus produce completely different stationary solutions. Such a non-uniqueness of the stationary solutions is often associated with the hysteresis phenomenon, where one or another solution can appear through non-stationary scenarios.

The model also contains a criterion of wave breaking, which is characterized by the Froude number between $1.3$ and $1.4$ depending on the energy dissipation parameter $\unicode[STIX]{x1D705}$ . An interesting feature of this model is the possibility of existence of a local subcritical zone $[x_{0},x_{1}]$ , which can be associated with wave breaking. The non-stationary wave breaking will be studied in the next sections.

4.1 Numerical method

For numerical treatment, the following dimensional form of the governing equations is used:

(4.1) $$\begin{eqnarray}\left.\begin{array}{@{}c@{}}\displaystyle \frac{\unicode[STIX]{x2202}\unicode[STIX]{x1D702}}{\unicode[STIX]{x2202}t}+\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}x}(\unicode[STIX]{x1D702}\bar{u})=\unicode[STIX]{x1D70E}q,\quad \frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}t}(h+\unicode[STIX]{x1D702})+\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}x}(hU+\unicode[STIX]{x1D702}\bar{u})=0,\\ \displaystyle \frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}t}(\unicode[STIX]{x1D702}\bar{u})+\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}x}\left((\bar{u}^{2}+q^{2})\unicode[STIX]{x1D702}+\frac{g\unicode[STIX]{x1D702}^{2}}{2}\right)=\unicode[STIX]{x1D70E}qU+\unicode[STIX]{x1D6FC}_{1},\\ \displaystyle \frac{\unicode[STIX]{x2202}q}{\unicode[STIX]{x2202}t}+\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}x}(q\bar{u})=\frac{\unicode[STIX]{x1D70E}}{2\unicode[STIX]{x1D702}}((U-\bar{u})^{2}-(1+\unicode[STIX]{x1D705})q^{2}),\\ \displaystyle \frac{\unicode[STIX]{x2202}K}{\unicode[STIX]{x2202}t}+\frac{\unicode[STIX]{x2202}}{\unicode[STIX]{x2202}x}\left(KU+g(h+\unicode[STIX]{x1D702})-\frac{U^{2}}{2}+\unicode[STIX]{x1D6FC}_{2}\right)=\unicode[STIX]{x1D70E}q\left(\frac{K-U}{h}+\unicode[STIX]{x1D6FC}_{3}\right)-gb_{x},\end{array}\right\}\end{eqnarray}$$

where

(4.2a-c ) $$\begin{eqnarray}\unicode[STIX]{x1D6FC}_{1}=-g\unicode[STIX]{x1D702}(h+b)_{x},\quad \unicode[STIX]{x1D6FC}_{2}=-\frac{U_{x}^{2}h^{2}}{2}-h\unicode[STIX]{x1D70E}qU_{x},\quad \unicode[STIX]{x1D6FC}_{3}=h_{x}U_{x},\end{eqnarray}$$

and the velocity $U$ is determined by the equation

(4.3) $$\begin{eqnarray}K=U-\frac{1}{3h}(h^{3}U_{x})_{x}.\end{eqnarray}$$

For a given $K$ , equation (4.3) is an ordinary differential equation for $U$ . The function $z=b(x)$ defines the bottom variation; $\unicode[STIX]{x1D70E}$ and $\unicode[STIX]{x1D705}$ are constants. The preceding system (4.1) can be written in conservative form as

(4.4) $$\begin{eqnarray}\frac{\unicode[STIX]{x2202}\boldsymbol{u}}{\unicode[STIX]{x2202}t}+\frac{\unicode[STIX]{x2202}\boldsymbol{f}}{\unicode[STIX]{x2202}x}=\boldsymbol{g},\end{eqnarray}$$

where

(4.5) $$\begin{eqnarray}\left.\begin{array}{@{}c@{}}\displaystyle \boldsymbol{u}=(\unicode[STIX]{x1D702},\,h+\unicode[STIX]{x1D702},\,\unicode[STIX]{x1D702}\bar{u},\,q,\,K)^{\text{T}},\\ \displaystyle \boldsymbol{f}=\left(\unicode[STIX]{x1D702}\bar{u},\,hU+\unicode[STIX]{x1D702}\bar{u},\,(\bar{u}^{2}+q^{2})\unicode[STIX]{x1D702}+\frac{g\unicode[STIX]{x1D702}^{2}}{2},q\bar{u},KU+g(h+\unicode[STIX]{x1D702})-\frac{U^{2}}{2}+\unicode[STIX]{x1D6FC}_{2}\right)^{\text{T}},\\ \displaystyle \boldsymbol{g}=\left(\unicode[STIX]{x1D70E}q,\,0,\,\unicode[STIX]{x1D70E}qU+\unicode[STIX]{x1D6FC}_{1},\,\frac{\unicode[STIX]{x1D70E}}{2\unicode[STIX]{x1D702}}((U-\bar{u})^{2}-(1+\unicode[STIX]{x1D705})q^{2}),\,\unicode[STIX]{x1D70E}q\left(\frac{K-U}{h}+\unicode[STIX]{x1D6FC}_{3}\right)-gb_{x}\right)^{\text{T}}.\end{array}\right\}\end{eqnarray}$$

Following Le Metayer et al. (Reference Le Metayer, Gavrilyuk and Hank2010), we divide the numerical resolution of system (4.1) into three successive steps:

1. (1) the numerical approximations of the terms $\unicode[STIX]{x1D6FC}_{i},\;i=1,2,3$ , containing spatial derivatives (see relations (4.2));

2. (2) the time evolution of the conservative variables $\boldsymbol{u}$ using the method based on modification of Godunov’s scheme; and

3. (3) the resolution of the ordinary differential equation (4.3) to obtain the values of velocity $U$ from variables $h$ and $K$ .

To solve the differential balance laws (4.4) numerically, we implement here the Nessyahu & Tadmor (Reference Nessyahu and Tadmor1990) second-order central scheme (see also Russo Reference Russo, Rionero and Romano2005):

(4.6) $$\begin{eqnarray}\left.\begin{array}{@{}c@{}}\begin{array}{@{}c@{}}\displaystyle \boldsymbol{u}_{j}^{n+1/2}=\boldsymbol{u}_{j}^{n}-\unicode[STIX]{x1D6EC}\boldsymbol{f}_{j}^{\prime }/2+\boldsymbol{g}(\boldsymbol{u}_{j}^{n})\unicode[STIX]{x0394}t/2\quad (\unicode[STIX]{x1D6EC}=\unicode[STIX]{x0394}t/\unicode[STIX]{x0394}x),\end{array}\\ \begin{array}{@{}rcl@{}}\displaystyle \boldsymbol{u}_{j+1/2}^{n+1}\, & =\, & (\boldsymbol{u}_{j}^{n}+\boldsymbol{u}_{j+1}^{n})/2+(\boldsymbol{u}_{j}^{\prime }-\boldsymbol{u}_{j+1}^{\prime })/8-\unicode[STIX]{x1D6EC}(\boldsymbol{f}(\boldsymbol{u}_{j+1}^{n+1/2})-\boldsymbol{f}(\boldsymbol{u}_{j}^{n+1/2}))\\ \displaystyle \, & \, & +\,(\boldsymbol{g}(\boldsymbol{u}_{j}^{n})+\boldsymbol{g}(\boldsymbol{u}_{j+1}^{n}))\unicode[STIX]{x0394}t/2.\end{array}\end{array}\right\}\end{eqnarray}$$

Here $\boldsymbol{u}_{j}^{n}=\boldsymbol{u}(t^{n},x_{j})$ , $\unicode[STIX]{x0394}x$ is the spatial grid spacing and $\unicode[STIX]{x0394}t$ is the time step satisfying the Courant condition, defined by the velocity of characteristics of the dispersionless part of the model. The computational domain on the $x$ axis is divided into $N$ cells, the cell centres being denoted by $x_{j}$ . Values $\boldsymbol{u}_{j}^{\prime }/\unicode[STIX]{x0394}x$ and $\boldsymbol{f}_{j}^{\prime }/\unicode[STIX]{x0394}x$ are approximations of the first-order derivatives with respect to $x$ , calculated according to the ‘MinMod’ procedure. At $t=0$ the initial data $\boldsymbol{u}_{j}^{0}$ are specified. The boundary conditions $\boldsymbol{u}_{1-k}^{n}$ and $\boldsymbol{u}_{N+k}^{n}$ ( $k=1,2$ ) should also be prescribed. Since the state $\boldsymbol{u}_{j}^{n}$ is known using formulae (4.6), one can obtain the conservative variable $\boldsymbol{u}$ at time $t^{n+1}$ . Scheme (4.6) does not require an exact or approximate solution of the Riemann problem, which is very convenient in our case, since system (4.1) includes five equations. It should be noted that other schemes (in particular, on non-staggered grids) based on the local Lax–Friedrichs flux are also appropriate here.

The numerical estimation of the first-order derivatives of any function $\unicode[STIX]{x1D711}$ at $x=x_{j}$ is given by the following relation:

(4.7a,b ) $$\begin{eqnarray}\left(\frac{\unicode[STIX]{x2202}\unicode[STIX]{x1D711}}{\unicode[STIX]{x2202}x}\right)_{j}=\frac{\unicode[STIX]{x1D711}_{j+1/2}-\unicode[STIX]{x1D711}_{j-1/2}}{\unicode[STIX]{x0394}x},\quad \unicode[STIX]{x1D711}_{j\pm 1/2}=\frac{\unicode[STIX]{x1D711}_{j\pm 1}+\unicode[STIX]{x1D711}_{j}}{2}.\end{eqnarray}$$

This discretization is used to determine the terms $\unicode[STIX]{x1D6FC}_{i}$ in (4.2). As a result, we obtain the flux $\boldsymbol{f}_{j}^{n}$ at time $t^{n}$ .

The last ingredient concerns the determination of the velocity $U_{j}$ at the next time step when the vector of conservative variables $\boldsymbol{u}_{j}^{n+1}$ is known. Applying finite difference discretization (4.7), one can rewrite equation (4.3) in the following form:

(4.8) $$\begin{eqnarray}a_{j-1}U_{j-1}-c_{j}U_{j}+b_{j}U_{j+1}=-f_{j}\quad (j=1,\ldots ,N),\end{eqnarray}$$

where

(4.9a-d ) $$\begin{eqnarray}a_{i}=h_{j-1/2}^{3},\quad b_{j}=h_{j+1/2}^{3},\quad c_{j}=a_{j}+b_{j}+3\unicode[STIX]{x0394}x^{2}h_{j},\quad f_{j}=3\unicode[STIX]{x0394}x^{2}h_{j}K_{j}.\end{eqnarray}$$

All the variables in the preceding relation are those at time $t^{n+1}$ . As a consequence, the variables $h_{j}$ and $K_{j}$ are known and come from the time evolution of conservative variables with the help of the numerical scheme presented above. The only unknowns are the terms $U_{j}$ at each node. Relations represent a tridiagonal system of linear equations that can be solved by a direct (Gauss) method. We apply here a simplified form of Gaussian elimination, which is known as the Thomas algorithm.

The next section is devoted to numerical results obtained with the present method for different applications.

4.2 Water hammer problem with a free surface (Favre waves)

Pioneering experimental work concerning the development of undular bores was performed by Favre (Reference Favre1935), who studied their formation in long open channels due to a rapid opening or closing of a gate. Similar experiments with a detailed investigation of the formation of surge waves in open channels were performed by Treske (Reference Treske1994) and by Soares-Frazão & Zech (Reference Soares-Frazão and Zech2002). The analytical approach to the study of undular bores was developed in El, Grimshaw & Kamchatnov (Reference El, Grimshaw and Kamchatnov2005) and El, Grimshaw & Smyth (Reference El, Grimshaw and Smyth2006) within a framework of the Serre–Su–Gardner–Green–Naghdi equations (Serre Reference Serre1953; Su & Gardner Reference Su and Gardner1969; Green & Naghdi Reference Green and Naghdi1976). This problem was studied numerically in Soares-Frazão & Guinot (Reference Soares-Frazão and Guinot2008) and Tissier et al. (Reference Tissier, Bonneton, Marche, Chazel and Lannes2011). Laboratory experiments (Treske Reference Treske1994; Chanson Reference Chanson2009) showed that the wave Froude number $F$ , which is defined as

(4.10) $$\begin{eqnarray}F=\frac{U_{0}-D}{\sqrt{gH_{0}}},\end{eqnarray}$$

controls the bore shape. In the supercritical regime ( $F>1$ ), undulations start developing at the bore front in the near-critical state ( $F\approx 1$ ). However, when the Froude number is approximately between 1.3 and 1.4, the transition from an undular bore to a bore consisting of a steep front (breaking bore) occurs. Figure 4 shows a definition sketch for the Favre waves in the undular regime. Constants $U_{0}$ and $H_{0}$ are the upstream flow velocity and depth. In the following, $D$ denotes the propagation speed of the wavefront; $a_{m}$ represents the jump height; and $a_{max}$ and $a_{min}$ are the maximum and minimum amplitude of the first wave. Below, we present some results related to the numerical modelling of Favre waves on the basis of (4.1)–(4.3) in the domain of Froude number between $1$ and $1.3$ corresponding to the domain of existence of the undular bore. We also demonstrate the transition from undular to purely breaking bores with increasing Froude number, and, in particular, compare numerical results for the first wave amplitude with experimental data (Treske Reference Treske1994).

Figure 4.

Definition sketch for the Favre waves.

We perform calculations in dimensionless variables in the domain $x\in [0,100]$ for $N=1000$ . We take here $g=1$ , $\unicode[STIX]{x1D70E}=0.15$ and $\unicode[STIX]{x1D705}=3$ . To start the calculations, we specify the unknown variables $\boldsymbol{u}$ in the node points $x=x_{j}$ at $t=0$ as follows: $\unicode[STIX]{x1D702}=0.01$ , $h+\unicode[STIX]{x1D702}=1$ , $\bar{u}=K=U_{0}$ and $q=0$ . On the left boundary we assume that $\boldsymbol{u}_{1-k}^{n}=\boldsymbol{u}_{1}^{n}$ . These conditions allow us to calculate until the perturbations reach the boundary of the computational domain. To satisfy the impermeability condition ( $\bar{u}=U=0$ ) on the right wall $x=100$ , we use the reflecting boundary conditions: $\boldsymbol{u}_{N+k}^{n}=\boldsymbol{u}_{N-k}^{n}$ except for variables $\unicode[STIX]{x1D702}\bar{u}$ and $K$ which are treated as $(\unicode[STIX]{x1D702}\bar{u})_{N+k}^{n}=-(\unicode[STIX]{x1D702}\bar{u})_{N-k}^{n}$ and $K_{N+k}^{n}=-K_{N-k}^{n}$ .

It should be noted that, instead of the upstream velocity $U_{0}$ , it is more convenient to prescribe the upstream Froude number $F$ . A simple relation between them can be obtained. Indeed, according to conservation of mass and momentum, we have the following relations (in dimensional variables):

(4.11) $$\begin{eqnarray}\left.\begin{array}{@{}c@{}}\displaystyle (U_{0}-D)H_{0}=(U_{1}-D)H_{1},\\ \displaystyle (U_{0}-D)^{2}H_{0}+\frac{gH_{0}^{2}}{2}=(U_{1}-D)^{2}H_{1}+\frac{gH_{1}^{2}}{2},\end{array}\right\}\end{eqnarray}$$

where $U_{0}$ and $U_{1}$ are the upstream and downstream velocities, respectively, and $H_{0}$ and $H_{1}=H_{0}+a_{m}$ are the corresponding water depths. In our case, $U_{1}=0$ . The Bélanger formula

(4.12) $$\begin{eqnarray}\frac{H_{1}}{H_{0}}=\frac{\sqrt{1+8F^{2}}-1}{2}\end{eqnarray}$$

is the consequence of (4.10) and (4.11). Taking into account (4.12), one can express $U_{0}$ from (4.11) in terms of the Froude number:

(4.13) $$\begin{eqnarray}U_{0}=U_{1}+\sqrt{gH_{0}}\left(F-\frac{1+\sqrt{1+8F^{2}}}{4F}\right).\end{eqnarray}$$

Figure 5 shows the results of computations at dimensionless time instant $t_{\ast }=90$ for $F=1.28$ (undular bore) and $F=1.40$ (breaking bore) which correspond to the following upstream velocities: $U_{0}=0.3511$ and $U_{0}=0.4921$ . Thus, the proposed model (4.1)–(4.3) allows one to describe both undular and breaking bores without any ‘switching’ criteria. In both cases, the surface turbulent layer is formed and shear velocity $q$ is generated. The thickness of the layer grows with increasing jump amplitude $a_{m}$ (or Froude number $F$ ). Also, it is interesting to note that the region of high turbulence (lower curve in the figure corresponding to the scaled variable $q^{\ast }=q/U_{0}$ ) is situated at the crests of the waves (figure 5 a for undular bores), or immediately at the wavefront (figure 5 b for breaking bores). In the latter case the dispersion effects are clearly negligible and the hydrostatic approximation can be used. Such an intensive region of turbulence can be associated with the roller appearance.

Further we compare the amplitudes of undular bores obtained by model (4.1)–(4.3) with experimental data of Treske (Reference Treske1994, p. 365, figure 21). In this case we perform calculations in the domain $x\in [0,200]$ with $N=4000$ for different Froude numbers from the interval $F\in [1.10,1.30]$ . For larger Froude numbers, the amplitude of the leading wave rapidly decreases and transition to a breaking bore occurs. As in the previous case, we choose $g=1$ , $H_{0}=1$ , $\unicode[STIX]{x1D70E}=0.15$ and $\unicode[STIX]{x1D705}=3$ . The maximum and minimum amplitudes of the leading wave are taken at $t_{\ast }=190$ . Figure 6 shows that the results obtained by the model (4.1)–(4.3) are in good agreement with the experimental data.

Figure 5.

Favre waves: the free surface $h+\unicode[STIX]{x1D702}$ , the thickness of the lower potential layer $h$ , and the scaled shear velocity $q^{\ast }=q/U_{0}$ at $t_{\ast }=90$ , $t_{\ast }=t\sqrt{g/H_{0}}$ for $F=1.28$ (a) and $F=1.40$ (b). The parameters are as follows: $g=1$ , $\unicode[STIX]{x1D70E}=0.15$ and $\unicode[STIX]{x1D705}=3$ .

Figure 6.

Amplitude of undular bores for different Froude numbers.

4.3 The shoaling and breaking of a solitary wave

Here we consider the evolution of breaking solitary waves on a mildly sloping beach and compare numerical results with the experiments of Hsiao et al. (Reference Hsiao, Hsu, Lin and Chang2008). We perform calculations of the model (4.1)–(4.3) in the interval $x\in [0,200]$ . The bottom topography is specified as follows: $b(x)=0$ for $x\in [0,50]$ ; $b(x)=(x-50)/60$ for $x\in (50,175)$ ; and $b(x)=25/12$ for $x\in [175,200]$ . To avoid modelling of the wave propagation over dry bottom (run-up), we introduce a shelf zone near the right boundary. On both (left and right) walls we assume reflecting boundary conditions. Initial data represent a solitary wave having amplitude $a_{0}=0.3344~\text{m}$ that propagates with velocity $C_{0}=\sqrt{g(a_{0}+H_{0})}$ , where $H_{0}=2.2~\text{m}$ is the undisturbed water depth and $g=9.8~\text{m}~\text{s}^{-2}$ . These data correspond to trial 43 in the experiments by Hsiao et al. (Reference Hsiao, Hsu, Lin and Chang2008). The initial thickness of the upper turbulent layer $\unicode[STIX]{x1D702}_{0}$ is equal to $0.01~\text{m}$ . The parameter values are the same as those in the case of Favre waves: $\unicode[STIX]{x1D70E}=0.15$ and $\unicode[STIX]{x1D705}=3$ . We take $N=2000$ nodes for space resolution. Here we use dimensional variables.

Figure 7 shows the results of numerical calculations of the free surface at different time moments. As we can see from figure 7, when the wave breaks, the turbulent layer near the free surface is formed ( $\unicode[STIX]{x1D702}$ is increasing).

Figure 7.

The evolution of a solitary wave on a mildly sloping beach: 1, bottom topography; 2, 3 and 4, free surface at $t_{\ast }=0$ , $50$ and $75$ , respectively; and 5, the boundary of the lower potential layer at $t_{\ast }=75$ , where the dimensionless time $t_{\ast }$ is given by $t_{\ast }=t\sqrt{g/H_{0}}$ .

The temporal evolution of the free-surface displacements at different positions along the channel is shown in figure 8. The numerical results were compared with the experimental data corresponding to trial 43 of the experiments by Hsiao et al. (Reference Hsiao, Hsu, Lin and Chang2008) (see figure 4(a) on p. 979 therein). Excellent agreement with the experimental data is observed everywhere except in the region near the right boundary corresponding to $x=172~\text{m}$ (the last graphic in figure 8). This is due to the fact that we have introduced the shelf zone for $x>175~\text{m}$ , while experimentally the wave continues running up the slope.

Figure 8.

Time history of free-surface displacement: solid curves, experimental data (Hsiao et al. Reference Hsiao, Hsu, Lin and Chang2008); dashed curves, numerical results.